A Prehistory of N-Categorical Physics: John C. Baez Aaron Lauda August 18, 2009

A Prehistory of n-Categorical Physics
arXiv:0908.2469v1 [hep-th] 18 Aug 2009
John C. Baez∗ Aaron Lauda†

August 18, 2009
Abstract
This paper traces the growing role of categories and n-categories in
physics, starting with groups and their role in relativity, and leading up
to more sophisticated concepts which manifest themselves in Feynman
diagrams, spin networks, string theory, loop quantum gravity, and topo-
logical quantum field theory. Our chronology ends around 2000, with just
a taste of later developments such as open-closed topological string theory,
the categorification of quantum groups, Khovanov homology, and Lurie’s
work on the classification of topological quantum field theories.
1 Introduction
This paper is a highly subjective chronology describing how physicists have be-
gun to use ideas from n-category theory in their work, often without making
this explicit. Somewhat arbitrarily, we start around the discovery of relativity
and quantum mechanics, and lead up to conformal field theory and topological
field theory. In parallel, we trace a bit of the history of n-categories, from Eilen-
berg and Mac Lane’s introduction of categories, to later work on monoidal and
braided monoidal categories, to Grothendieck’s dreams involving ∞-categories,
and subsequent attempts to realize this dream. Our chronology ends at the
dawn of the 21st century; after then, developments have been coming so thick
and fast that we have not had time to put them in proper perspective.
We call this paper a ‘prehistory’ because n-categories and their applications
to physics are still in their infancy. We call it ‘a’ prehistory because it represents
just one view of a multi-faceted subject: many other such stories can and should
be told. Ross Street’s Conspectus of Australian Category Theory [1] is a good
example: it overlaps with ours, but only slightly. There are many aspects of
n-categorical physics that our chronology fails to mention, or touches on very
briefly; other stories could redress these deficiencies. It would also be good
to have a story of n-categories that focused on algebraic topology, one that
∗ Department of Mathematics, University of California, Riverside, CA 92521, USA. Email:
baez@math.ucr.edu
† Department of Mathematics, Columbia University, New York, NY 10027, USA. Email:
lauda@math.columbia.edu
1
2
focused on algebraic geometry, and one that focused on logic. For n-categories
in computer science, we have John Power’s Why Tricategories? [2], which while
not focused on history at least explains some of the issues at stake.
What is the goal of this prehistory? We are scientists rather than historians
of science, so we are trying to make a specific scientific point, rather than accu-
rately describe every twist and turn in a complex sequence of events. We want
to show how categories and even n-categories have slowly come to be seen as
a good way to formalize physical theories in which ‘processes’ can be drawn as
diagrams—for example Feynman diagrams—but interpreted algebraically—for
example as linear operators. To minimize the prerequisites, we include a gentle
introduction to n-categories (in fact, mainly just categories and bicategories).
We also include a review of some relevant aspects of 20th-century physics.
The most obvious roads to n-category theory start from issues internal to
pure mathematics. Applications to physics only became visible much later,
starting around the 1980s. So far, these applications mainly arise around theo-
ries of quantum gravity, especially string theory and ‘spin foam models’ of loop
quantum gravity. These theories are speculative and still under development,
not ready for experimental tests. They may or may not succeed. So, it is too
early to write a real history of n-categorical physics, or even to know if this
subject will become important. We believe it will—but so far, all we have is a
‘prehistory’.
2 Road Map
Before we begin our chronology, to help the reader keep from getting lost in a
cloud of details, it will be helpful to sketch the road ahead. Why did categories
turn out to be useful in physics? The reason is ultimately very simple. A cate-
gory consists of ‘objects’ x, y, z, . . . and ‘morphisms’ which go between objects,
for example
f : x → y.
A good example is the category of Hilbert spaces, where the objects are Hilbert
spaces and the morphisms are bounded operators. In physics we can think
of an object as a ‘state space’ for some physical system, and a morphism as
a ‘process’ taking states of one system to states of another (perhaps the same
one). In short, we use objects to describe kinematics, and morphisms to describe
dynamics.
Why n-categories? For this we need to understand a bit about categories
and their limitations. In a category, the only thing we can do with morphisms
is ‘compose’ them: given a morphism f : x → y and a morphism g : y → z, we
can compose them and obtain a morphism gf : x → z. This corresponds to our
basic intuition about processes, namely that one can occur after another. While
this intuition is temporal in nature, it lends itself to a nice spatial metaphor.
We can draw a morphism f : x → y as a ‘black box’ with an input of type x
3
and an output of type y:

x
f
y
Composing morphisms then corresponds to feeding the output of one black box
into another:
x
f
y
g
z
This sort of diagram might be sufficient to represent physical processes if the
universe were 1-dimensional: no dimensions of space, just one dimension of time.
But in reality, processes can occur not just in series but also in parallel —‘side
by side’, as it were:
x x′
f f′
y y′
To formalize this algebraically, we need something more than a category: at the

very least a ‘monoidal category’, which is a special sort of ‘bicategory’. The
term ‘bicategory’ hints at the two ways of combining processes: in series and in
parallel.
Similarly, the mathematics of bicategories might be sufficient for physics if
the universe were only 2-dimensional: one dimension of space, one dimension
of time. But in our universe, is also possible for physical systems to undergo a
special sort of process where they ‘switch places’:
x y
To depict this geometrically requires a third dimension, hinted at here by the

crossing lines. To formalize it algebraically, we need something more than a
monoidal category: at the very least a ‘braided monoidal category’, which is a
special sort of ‘tricategory’.
This escalation of dimensions can continue. In the diagrams Feynman used
to describe interacting particles, we can continuously interpolate between this
4
way of switching two particles:
x y
and this:
x y
This requires four dimensions: one of time and three of space. To formalize this
algebraically we need a ‘symmetric monoidal category’, which is a special sort
of ‘tetracategory’.
More general n-categories, including those for higher values of n, may also be
useful in physics. This is especially true in string theory and spin foam models
of quantum gravity. These theories describe strings, graphs, and their higher-
dimensional generalizations propagating in spacetimes which may themselves
have more than 4 dimensions.
So, in abstract the idea is simple: we can use n-categories to algebraically
formalize physical theories in which processes can be depicted geometrically
using n-dimensional diagrams. But the development of this idea has been long
and convoluted. It is also far from finished. In our chronology we describe its
development up to the year 2000. To keep the tale from becoming unwieldy, we
have been ruthlessly selective in our choice of topics.
In particular, we can roughly distinguish two lines of thought leading towards
n-categorical physics: one beginning with quantum mechanics, the other with
general relativity. Since a major challenge in physics is reconciling quantum
mechanics and general relativity, it is natural to hope that these lines of thought
will eventually merge. We are not sure yet how this will happen, but the two
lines have already been interacting throughout the 20th century. Our chronology
will focus on the first. But before we start, let us give a quick sketch of both.
The first line of thought starts with quantum mechanics and the realization
that in this subject, symmetries are all-important. Taken abstractly, the sym-
metries of any system form a group G. But to describe how these symmetries
act on states of a quantum system, we need a ‘unitary representation’ ρ of this
group on some Hilbert space H. This sends any group element g ∈ G to a
unitary operator ρ(g) : H → H.
The theory of n-categories allows for drastic generalizations of this idea. We
can see any group G as a category with one object where all the morphisms are
invertible: the morphisms of this category are just the elements of the group,
while composition is multiplication. There is also a category Hilb where objects
are Hilbert spaces and morphisms are linear operators. A representation of G
can be seen as a map from the first category to the second:
ρ : G → Hilb.
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Such a map between categories is called a ‘functor’. The functor ρ sends the
one object of G to the Hilbert space H, and it sends each morphism g of G to
a unitary operator ρ(g) : H → H. In short, it realizes elements of the abstract
group G as actual transformations of a specific physical system.
The advantage of this viewpoint is that now the group G can be replaced
by a more general category. Topological quantum field theory provides the
most famous example of such a generalization, but in retrospect the theory
of Feynman diagrams provides another, and so does Penrose’s theory of ‘spin
networks’.
More dramatically, both G and Hilb may be replaced by a more general
sort of n-category. This allows for a rigorous treatment of physical theories
where physical processes are described by n-dimensional diagrams. The basic
idea, however, is always the same: a physical theory is a map sending ‘abstract’
processes to actual transformations of a specific physical system.
The second line of thought starts with Einstein’s theory of general relativity,
which explains gravity as the curvature of spacetime. Abstractly, the presence
of ‘curvature’ means that as a particle moves through spacetime from one point
to another, its internal state transforms in a manner that depends nontrivially
on the path it takes. Einstein’s great insight was that this notion of curvature
completely subsumes the older idea of gravity as a ‘force’. This insight was later
generalized to electromagnetism and the other forces of nature: we now treat
them all as various kinds of curvature.
In the language of physics, theories where forces are explained in terms of
curvature are called ‘gauge theories’. Mathematically, the key concept in a gauge
theory is that of a ‘connection’ on a ‘bundle’. The idea here is to start with a
manifold M describing spacetime. For each point x of spacetime, a bundle gives
a set Ex of allowed internal states for a particle at this point. A connection
then assigns to each path γ from x ∈ M to y ∈ M a map ρ(γ) : Ex → Ey . This
map, called ‘parallel transport’, says how a particle starting at x changes state
if it moves to y along the path γ.
Category theory lets us see that a connection is also a kind of functor.
There is a category called the ‘path groupoid’ of M , denoted P1 (M ), whose
objects are points of M : the morphisms are paths, and composition amounts
to concatenating paths. Similarly, any bundle E gives a ’transport category’,
denoted Trans(E), where the objects are the sets Ex and the morphisms are
maps between these. A connection gives a functor
ρ : P1 (M ) → Trans(P ).
This functor sends each object x of P1 (M ) to the set Ex , and sends each path
γ to the map ρ(γ).
So, the ‘second line of thought’, starting from general relativity, leads to
a picture strikingly similar to the first! Just as a unitary group representa-
tion is a functor sending abstract symmetries to transformations of a specific
physical system, a connection is a functor sending paths in spacetime to trans-
formations of a specific physical system: a particle. And just as unitary group
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representations are a special case of physical theories described as maps between

n-categories, when we go from point particles to higher-dimensional objects we
meet ‘higher gauge theories’, which use maps between n-categories to describe
how such objects change state as they move through spacetime [3]. In short:
the first and second lines of thought are evolving in parallel—and intimately
linked, in ways that still need to be understood.
Sadly, we will not have much room for general relativity, gauge theories, or
higher gauge theories in our chronology. We will be fully occupied with group
representations as applied to quantum mechanics, Feynman diagrams as applied
to quantum field theory, how these diagrams became better understood with
the rise of n-categories, and how higher-dimensional generalizations of Feynman
diagrams arise in string theory, loop quantum gravity, topological quantum field
theory, and the like.
3 Chronology
Maxwell (1876)
In his book Matter and Motion, Maxwell [4] wrote:
Our whole progress up to this point may be described as a gradual

development of the doctrine of relativity of all physical phenomena.
Position we must evidently acknowledge to be relative, for we cannot
describe the position of a body in any terms which do not express
relation. The ordinary language about motion and rest does not so
completely exclude the notion of their being measured absolutely,
but the reason of this is, that in our ordinary language we tacitly
assume that the earth is at rest.... There are no landmarks in space;
one portion of space is exactly like every other portion, so that we
cannot tell where we are. We are, as it were, on an unruffled sea,
without stars, compass, sounding, wind or tide, and we cannot tell in
what direction we are going. We have no log which we can case out
to take a dead reckoning by; we may compute our rate of motion
with respect to the neighboring bodies, but we do not know how
these bodies may be moving in space.
Readers less familiar with the history of physics may be surprised to see
these words, written 3 years before Einstein was born. In fact, the relative
nature of velocity was already known to Galileo, who also used a boat analogy
to illustrate this. However, Maxwell’s equations describing light made relativity
into a hot topic. First, it was thought that light waves needed a medium to
propagate in, the ‘luminiferous aether’, which would then define a rest frame.
Second, Maxwell’s equations predicted that waves of light move at a fixed speed
in vacuum regardless of the velocity of the source! This seemed to contradict
the relativity principle. It took the genius of Lorentz, Poincaré, Einstein and
Minkowski to realize that this behavior of light is compatible with relativity of
7
motion if we assume space and time are united in a geometrical structure we now
call Minkowski spacetime. But when this realization came, the importance of
the relativity principle was highlighted, and with it the importance of symmetry
groups in physics.
Poincaré (1894)
In 1894, Poincaré invented the fundamental group: for any space X with a
basepoint ∗, homotopy classes of loops based at ∗ form a group π1 (X). This
hints at the unification of space and symmetry, which was later to become one
of the main themes of n-category theory. In 1945, Eilenberg and Mac Lane
described a kind of ‘inverse’ to the process taking a space to its fundamental
group. Since the work of Grothendieck in the 1960s, many have come to believe
that homotopy theory is secretly just the study of certain vast generalizations
of groups, called ‘n-groupoids’. From this point of view, the fundamental group
is just the tip of an iceberg.
Lorentz (1904)
Already in 1895 Lorentz had invented the notion of ‘local time’ to explain the
results of the Michelson–Morley experiment, but in 1904 he extended this work
and gave formulas for what are now called ‘Lorentz transformations’ [5].
Poincaré (1905)
In his opening address to the Paris Congress in 1900, Poincaré asked ‘Does the
aether really exist?’ In 1904 he gave a talk at the International Congress of
Arts and Science in St. Louis, in which he noted that “. . . as demanded by the
relativity principle the observer cannot know whether he is at rest or in absolute
motion”.
On the 5th of June, 1905, he wrote a paper ‘Sur la dynamique de l’electron’
[6] in which he stated: “It seems that this impossibility of demonstrating abso-
lute motion is a general law of nature”. He named the Lorentz transformations
after Lorentz, and showed that these transformations, together with the rota-
tions, form a group. This is now called the ‘Lorentz group’.
Einstein (1905)
Einstein’s first paper on relativity, ‘On the electrodynamics of moving bodies’ [7]
was received on June 30th, 1905. In the first paragraph he points out problems
that arise from applying the concept of absolute rest to electrodynamics. In the
second, he continues:
Examples of this sort, together with the unsuccessful attempts to dis-
cover any motion of the earth relative to the ‘light medium,’ suggest
that the phenomena of electrodynamics as well as of mechanics pos-
sess no properties corresponding to the idea of absolute rest. They
8
suggest rather that, as already been shown to the first order of small
quantities, the same laws of electrodynamics and optics hold for all
frames of reference for which the equations of mechanics hold good.
We will raise this conjecture (the purport of which will hereafter be
called the ‘Principle of Relativity’) to the status of a postulate, and
also introduce another postulate, which is only apparently irrecon-
cilable with the former, namely, that light is always propagated in
empty space with a definite velocity c which is independent of the
state of motion of the emitting body.
From these postulates he derives formulas for the transformation of coordi-

nates from one frame of reference to another in uniform motion relative to the
first, and shows these transformations form a group.
Minkowski (1908)
In a famous address delivered at the 80th Assembly of German Natural Scientists
and Physicians on September 21, 1908, Hermann Minkowski declared:
The views of space and time which I wish to lay before you have
sprung from the soil of experimental physics, and therein lies their
strength. They are radical. Henceforth space by itself, and time by
itself, are doomed to fade away into mere shadows, and only a kind
of union of the two will preserve an independent reality.
He formalized special relativity by treating space and time as two aspects of

a single entity: spacetime. In simple terms we may think of this as R4 , where a
point x = (t, x, y, z) describes the time and position of an event. Crucially, this
R4 is equipped with a bilinear form, the Minkowski metric:
x · x′ = tt′ − xx′ − yy ′ − zz ′
which we use as a replacement for the usual dot product when calculating times
and distances. With this extra structure, R4 is now called Minkowski space-
time. The group of all linear transformations
T : R4 → R4
preserving the Minkowski metric is called the Lorentz group, and denoted
O(3, 1).
Heisenberg (1925)
In 1925, Werner Heisenberg came up with a radical new approach to physics in
which processes were described using matrices [8]. What makes this especially
remarkable is that Heisenberg, like most physicists of his day, had not heard
of matrices! His idea was that given a system with some set of states, say
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{1, . . . , n}, a process U would be described by a bunch of complex numbers Uji

specifying the ‘amplitude’ for any state i to turn into any state j. He composed
processes by summing over all possible intermediate states:
X j
(V U )ik = Vk Uji .
j
Later he discussed his theory with his thesis advisor, Max Born, who informed
him that he had reinvented matrix multiplication.
Heisenberg never liked the term ‘matrix mechanics’ for his work, because he
thought it sounded too abstract. However, it is an apt indication of the algebraic
flavor of quantum physics.
Born (1928)
In 1928, Max Born figured out what Heisenberg’s mysterious ‘amplitudes’ actu-
ally meant: the absolute value squared |Uji |2 gives the probability for the initial
state i to become the final state j via the process U . This spelled the end of the
deterministic worldview built into Newtonian mechanics [9]. More shockingly
still, since amplitudes are complex, a sum of amplitudes can have a smaller
absolute value than those of its terms. Thus, quantum mechanics exhibits de-
structive interference: allowing more ways for something to happen may reduce
the chance that it does!
Von Neumann (1932)

In 1932, John von Neumann published a book on the foundations of quantum
mechanics [10], which helped crystallize the now-standard approach to this the-
ory. We hope that the experts will forgive us for omitting many important
subtleties and caveats in the following sketch.
Every quantum system has a Hilbert space of states, H. A state of the
system is described by a unit vector ψ ∈ H. Quantum theory is inherently
probabilistic: if we put the system in some state ψ and immediately check to
see if it is in the state φ, we get the answer ‘yes’ with probability equal to
|hφ, ψi|2 .
A reversible process that our system can undergo is called a symmetry.
Mathematically, any symmetry is described by a unitary operator U : H → H.
If we put the system in some state ψ and apply the symmetry U it will then be in
the state U ψ. If we then check to see if it is in some state φ, we get the answer
‘yes’ with probability |hφ, U ψi|2 . The underlying complex number hφ, U ψi is
called a transition amplitude. In particular, if we have an orthonormal basis
ei of H, the numbers
Uji = hej , U ei i
are Heisenberg’s matrices!
Thus, Heisenberg’s matrix mechanics is revealed to be part of a framework
in which unitary operators describe physical processes. But, operators also play
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another role in quantum theory. A real-valued quantity that we can measure by

doing experiments on our system is called an observable. Examples include
energy, momentum, angular momentum and the like. Mathematically, any ob-
servable is described by a self-adjoint operator A on the Hilbert space H for the
system in question. Thanks to the probabilistic nature of quantum mechanics,
we can obtain various different values when we measure the observable A in the
state ψ, but the average or ‘expected’ value will be hψ, Aψi.
If a group G acts as symmetries of some quantum system, we obtain a uni-
tary representation of G, meaning a Hilbert space H equipped with unitary
operators
ρ(g) : H → H,
one for each g ∈ G, such that
ρ(1) = 1H
and
ρ(gh) = ρ(g)ρ(h).
Often the group G will be equipped with a topology. Then we want symmetry
transformation close to the identity to affect the system only slightly, so we
demand that if gi → 1 in G, then ρ(gi )ψ → ψ for all ψ ∈ H. Professionals
use the term strongly continuous for representations with this property, but
we shall simply call them continuous, since we never discuss any other sort of
continuity.
Continuity turns out to have powerful consequences, such as the Stone–von
Neumann theorem: if ρ is a continuous representation of R on H, then
ρ(s) = exp(−isA)
for a unique self-adjoint operator A on H. Conversely, any self-adjoint operator

gives a continuous representation of R this way. In short, there is a corre-
spondence between observables and one-parameter groups of symmetries. This
links the two roles of operators in quantum mechanics: self-adjoint operators
for observables, and unitary operators for symmetries.
Wigner (1939)
We have already discussed how the Lorentz group O(3, 1) acts as symmetries of
spacetime in special relativity: it is the group of all linear transformations
T : R4 → R4
preserving the Minkowski metric. However, the full symmetry group of Minkowski
spacetime is larger: it includes translations as well. So, the really important
group in special relativity is the so-called ‘Poincaré group’:
P = O(3, 1) ⋉ R4
generated by Lorentz transformations and translations.

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Some subtleties appear when we take some findings from particle physics
into account. Though time reversal
(t, x, y, z) 7→ (−t, x, y, z)
and parity
(t, x, y, z) 7→ (t, −x, −y, −z)
are elements of P, not every physical system has them as symmetries. So it
is better to exclude such elements of the Poincaré group by working with the
connected component of the identity, P0 . Furthermore, when we rotate an
electron a full turn, its state vector does not come back to where it stated: it
gets multiplied by -1. If we rotate it two full turns, it gets back to where it
started. To deal with this, we should replace P0 by its universal cover, P̃0 . For
lack of a snappy name, in what follows we call this group the Poincaré group.
We have seen that in quantum mechanics, physical systems are described by
continuous unitary representations of the relevant symmetry group. In relativis-
tic quantum mechanics, this symmetry group is P̃0 . The Stone-von Neumann
theorem then associates observables to one-parameter subgroups of this group.
The most important observables in physics—energy, momentum, and angular
momentum—all arise this way!
For example, time translation
gs : (t, x, y, z) 7→ (t + s, x, y, z)
gives rise to an observable A with
ρ(gs ) = exp(−isA).
and this observable is the energy of the system, also known as the Hamiltonian.
If the system is in a state described by the unit vector ψ ∈ H, the expected
value of its energy is hψ, Aψi. In the context of special relativity, the energy
of a system is always greater than or equal to that of the vacuum (the empty
system, as it were). The energy of the vacuum is zero, so it makes sense to focus
attention on continuous unitary representations of the Poincaré group with
hψ, Aψi ≥ 0.
These are usually called positive-energy representations.

In a famous 1939 paper, Eugene Wigner [11] classified the positive-energy
representations of the Poincaré group. All these representations can be built
as direct sums of irreducible ones, which serve as candidates for describing
‘elementary particles’: the building blocks of matter. To specify one of these
representations, we need to give a number m ≥ 0 called the ‘mass’ of the particle,
a number j = 0, 21 , 1, . . . called its ‘spin’, and sometimes a little extra data.
For example, the photon has spin 1 and mass 0, while the electron has spin 12
and mass equal to about 9 · 10−31 kilograms. Nobody knows why particles have
the masses they do—this is one of the main unsolved problems in physics—but
they all fit nicely into Wigner’s classification scheme.
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Eilenberg–Mac Lane (1945)

Eilenberg and Mac Lane [12] invented the notion of a ‘category’ while work-
ing on algebraic topology. The idea is that whenever we study mathematical
gadgets of any sort—sets, or groups, or topological spaces, or positive-energy
representations of the Poincaré group, or whatever—we should also study the
structure-preserving maps between these gadgets. We call the gadgets ‘objects’
and the maps ‘morphisms’. The identity map is always a morphism, and we
can compose morphisms in an associative way.
Eilenberg and Mac Lane thus defined a category C to consist of:
• a collection of objects,
• for any pair of objects x, y, a set of hom(x, y) of morphisms from x to
y, written f : x → y,
equipped with:
• for any object x, an identity morphism 1x : x → x,
• for any pair of morphisms f : x → y and g : y → z, a morphism gf : x → z

called the composite of f and g,
such that:
• for any morphism f : x → y, the left and right unit laws hold: 1y f =
f = f 1x .
• for any triple of morphisms f : w → x, g : x → y, h : y → z, the associa-
tive law holds: (hg)f = h(gf ).
Given a morphism f : x → y, we call x the source of f and y the target of y.
Eilenberg and Mac Lane did much more than just define the concept of
category. They also defined maps between categories, which they called ‘func-
tors’. These send objects to objects, morphisms to morphisms, and preserve
all the structure in sight. More precisely, given categories C and D, a functor
F : C → D consists of:
• a function F sending objects in C to objects in D, and
• for any pair of objects x, y ∈ Ob(C), a function also called F sending
morphisms in hom(x, y) to morphisms in hom(F (x), F (y))
such that:
• F preserves identities: for any object x ∈ C, F (1x ) = 1F (x) ;

• F preserves composition: for any pair of morphisms f : x → y, g : y → z
in C, F (gf ) = F (g)F (f ).
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Many of the famous invariants in algebraic topology are actually functors,

and this is part of how we convert topology problems into algebra problems and
solve them. For example, the fundamental group is a functor
π1 : Top∗ → Grp.
from the category of topological spaces equipped with a basepoint to the cate-
gory of groups. In other words, not only does any topological space with base-
point X have a fundamental group π1 (X), but also any continuous map f : X →
Y preserving the basepoint gives a homomorphism π1 (f ) : π1 (X) → π1 (Y ), in
a way that gets along with composition. So, to show that the inclusion of the
circle in the disc S1 D2
i /
does not admit a retraction—that is, a map

D2 S1
r /
such that this diagram commutes:

D2
? ??
i ??r
??
S1
? S1
/
1S 1
we simply hit this question with the functor π1 and note that the homomorphism
π1 (i) : π1 (S 1 ) → π1 (D2 )
cannot have a homomorphism
π1 (r) : π1 (D2 ) → π1 (S 1 )
for which π1 (r)π1 (i) is the identity, because π1 (S 1 ) = Z and π1 (D2 ) = 0.

However, Mac Lane later wrote that the real point of this paper was not
to define categories, nor to define functors between categories, but to define
‘natural transformations’ between functors! These can be drawn as follows:
F
α

C• B•D

G
Given functors F, G : C → D, a natural transformation α : F ⇒ G consists

of:
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• a function α mapping each object x ∈ C to a morphism αx : F (x) → G(x)

such that:
• for any morphism f : x → y in C, this diagram commutes:
F (f )
F (x) / F (y)
αx αy

G(x) G(f )
/ G(y)
The commuting square here conveys the ideas that α not only gives a morphism
αx : F (x) → G(x) for each object x ∈ C, but does so ‘naturally’—that is, in a
way that is compatible with all the morphisms in C.
The most immediately interesting natural transformations are the natural
isomorphisms. When Eilenberg and Mac Lane were writing their paper, there
were many different recipes for computing the homology groups of a space, and
they wanted to formalize the notion that these different recipes give groups that
are not only isomorphic, but ‘naturally’ so. In general, we say a morphism
g : y → x is an isomorphism if it has an inverse: that is, a morphism f : x →
y for which f g and gf are identity morphisms. A natural isomorphism
between functors F, G : C → D is then a natural transformation α : F ⇒ G
such that αx is an isomorphism for all x ∈ C. Alternatively, we can define how
to compose natural transformations, and say a natural isomorphism is a natural
transformation with an inverse.
Invertible functors are also important—but here an important theme known
as ‘weakening’ intervenes for the first time. Suppose we have functors F : C → D
and G : D → C. It is unreasonable to demand that if we apply first F and then
G, we get back exactly the object we started with. In practice all we really
need, and all we typically get, is a naturally isomorphic object. So, we say
a functor F : C → D is an equivalence if it has a weak inverse, that is, a
functor G : D → C such that there exist natural isomorphisms α : GF ⇒ 1C ,
β : F G ⇒ 1D .
In the first applications to topology, the categories involved were mainly
quite large: for example, the category of all topological spaces, or all groups. In
fact, these categories are even ‘large’ in the technical sense, meaning that their
collection of objects is not a set but a proper class. But later applications of
category theory to physics often involved small categories.
For example, any group G can be thought of as a category with one object
and only invertible morphisms: the morphisms are the elements of G, and com-
position is multiplication in the group. A representation of G on a Hilbert space
is then the same as a functor
ρ : G → Hilb,
where Hilb is the category with Hilbert spaces as objects and bounded linear
operators as morphisms. While this viewpoint may seem like overkill, it is a
15
prototype for the idea of describing theories of physics as functors, in which

‘abstract’ physical processes (e.g. symmetries) get represented in a ‘concrete’
way (e.g. as operators). However, this idea came long after the work of Eilenberg
and Mac Lane: it was born sometime around Lawvere’s 1963 thesis, and came
to maturity in Atiyah’s 1988 definition of ‘topological quantum field theory’.
Feynman (1947)
After World War II, many physicists who had been working in the Manhattan
project to develop the atomic bomb returned to work on particle physics. In
1947, a small conference on this subject was held at Shelter Island, attended by
luminaries such as Bohr, Oppenheimer, von Neumann, Weisskopf, and Wheeler.
Feynman presented his work on quantum field theory, but it seems nobody
understood it except Schwinger, who was later to share the Nobel prize with
him and Tomonaga. Apparently it was a bit too far-out for most of the audience.
Feynman described a formalism in which time evolution for quantum sys-
tems was described using an integral over the space of all classical histories: a
‘Feynman path integral’. These are notoriously hard to make rigorous. But, he
also described a way to compute these perturbatively as a sum over diagrams:
‘Feynman diagrams’. For example, in QED, the amplitude for an electron to
absorb a photon is given by:
'' -- I '' -- I
'' K K - - I ''#c _ J J G -- I
' K -I ' ' J G - I
'' K J V
+ Z + '' J QM + · · · + Y Y + · · ·
U O J ^ ^ Y T O
I G ^ J
D C E
All these diagrams describe ways for an electron and photon to come in and an
electron to go out. Lines with arrows pointing downwards stand for electrons.
Lines with arrows pointing upwards stand for positrons: the positron is the
‘antiparticle’ of an electron, and Feynman realized that this could thought of as
an electron going backwards in time. The wiggly lines stand for photons. The
photon is its own antiparticle, so we do not need arrows on these wiggly lines.
Mathematically, each of the diagrams shown above is shorthand for a linear
operator
f : H e ⊗ Hγ → H e
where He is the Hilbert space for an electron, and Hγ is a Hilbert space for a
photon. We take the tensor product of group representations when combining
two systems, so He ⊗ Hγ is the Hilbert space for an photon together with an
electron.
As already mentioned, elementary particles are described by certain special
representations of the Poincaré group—the irreducible positive-energy ones. So,
He and Hγ are representations of this sort. We can tensor these to obtain
positive-energy representations describing collections of elementary particles.
Moreover, each Feynman diagram describes an intertwining operator: an
16
operator that commutes with the action of the Poincaré group. This expresses
the fact that if we, say, rotate our laboratory before doing an experiment, we
just get a rotated version of the result we would otherwise get.
So, Feynman diagrams are a notation for intertwining operators between
positive-energy representations of the Poincaré group. However, they are so
powerfully evocative that they are much more than a mere trick! As Feynman
recalled later [13]:
The diagrams were intended to represent physical processes and the

mathematical expressions used to describe them. Each diagram sig-
nified a mathematical expression. In these diagrams I was seeing
things that happened in space and time. Mathematical quantities
were being associated with points in space and time. I would see
electrons going along, being scattered at one point, then going over
to another point and getting scattered there, emitting a photon and
the photon goes there. I would make little pictures of all that was
going on; these were physical pictures involving the mathematical
terms.
Feynman first published papers containing such diagrams in 1949 [14, 15].
However, his work reached many physicists through expository articles pub-
lished even earlier by one of the few people who understood what he was up to:
Freeman Dyson [16, 17]. For more on the history of Feynman diagrams, see the
book by Kaiser [18].
The general context for such diagrammatic reasoning came much later, from
category theory. The idea is that we can draw a morphism f : x → y as an
arrow going down:
x
f

y
but then we can switch to a style of drawing in which the objects are depicted
not as dots but as ‘wires’, while the morphisms are drawn not as arrows but as
‘black boxes’ with one input wire and one output wire:
x x
f • or f
y y
This is starting to look a bit like a Feynman diagram! However, to get really
interesting Feynman diagrams we need black boxes with many wires going in
and many wires going out. The mathematics necessary for this was formalized
later, in Mac Lane’s 1963 paper on monoidal categories (see below) and Joyal
and Street’s 1980s work on ‘string diagrams’ [19].
17
Yang–Mills (1953)
In modern physics the electromagnetic force is described by a U(1) gauge field.
Most mathematicians prefer to call this a ‘connection on a principal U(1) bun-
dle’. Jargon aside, this means that if we carry a charged particle around a loop
in spacetime, its state will be multiplied by some element of U(1)—that is, a
phase—thanks to the presence of the electromagnetic field. Moreover, every-
thing about electromagnetism can be understood in these terms!
In 1953, Chen Ning Yang and Robert Mills [20] formulated a generalization
of Maxwell’s equations in which forces other than electromagnetism can be
described by connections on G-bundles for groups other than U(1). With a vast
amount of work by many great physicists, this ultimately led to the ‘Standard
Model’, a theory in which all forces other than gravity are described using a
connection on a principal G-bundle where
G = U(1) × SU(2) × SU(3).
Though everyone would like to more deeply understand this curious choice of
G, at present it is purely a matter of fitting the experimental data.
In the Standard Model, elementary particles are described as irreducible
positive-energy representations of P̃0 × G. Perturbative calculations in this
theory can be done using souped-up Feynman diagrams, which are a notation
for intertwining operators between positive-energy representations of P̃0 × G.
While efficient, the mathematical jargon in the previous paragraphs does
little justice to how physicists actually think about these things. For example,
Yang and Mills did not know about bundles and connections when formulating
their theory. Yang later wrote [21]:
What Mills and I were doing in 1954 was generalizing Maxwell’s

theory. We knew of no geometrical meaning of Maxwell’s theory, and
we were not looking in that direction. To a physicist, gauge potential
is a concept rooted in our description of the electromagnetic field.
Connection is a geometrical concept which I only learned around
1970.
Mac Lane (1963)

In 1963 Mac Lane published a paper describing the notion of a ‘monoidal cat-
egory’ [22]. The idea was that in many categories there is a way to take the
‘tensor product’ of two objects, or of two morphisms. A famous example is the
category Vect, where the objects are vector spaces and the morphisms are linear
operators. This becomes a monoidal category with the usual tensor product of
vector spaces and linear maps. Other examples include the category Set with
the cartesian product of sets, and the category Hilb with the usual tensor prod-
uct of Hilbert spaces. We will also be interested in FinVect and FinHilb, where
the objects are finite-dimensional vector spaces (resp. Hilbert spaces) and the
morphisms are linear maps. We will also get many examples from categories
18
of representations of groups. The theory of Feynman diagrams, for example,

turns out to be based on the symmetric monoidal category of positive-energy
representations of the Poincaré group!
In a monoidal category, given morphisms f : x → y and g : x′ → y ′ there is
a morphism
f ⊗ g : x ⊗ x′ → y ⊗ y ′ .
We can also draw this as follows:
x x′
f • •g
y y′
This sort of diagram is sometimes called a ‘string diagram’; the mathematics of

these was formalized later [19], but we can’t resist using them now, since they
are so intuitive. Notice that the diagrams we could draw in a mere category
were intrinsically 1-dimensional, because the only thing we could do is compose
morphisms, which we draw by sticking one on top of another. In a monoidal
category the string diagrams become 2-dimensional, because now we can also
tensor morphisms, which we draw by placing them side by side.
This idea continues to work in higher dimensions as well. The kind of cate-
gory suitable for 3-dimensional diagrams is called a ‘braided monoidal category’.
In such a category, every pair of objects x, y is equipped with an isomorphism
called the ‘braiding’, which switches the order of factors in their tensor product:
Bx,y : x ⊗ y → y ⊗ x.
We can draw this process of switching as a diagram in 3 dimensions:
x y
and the braiding Bx,y satisfies axioms that are related to the topology of 3-
dimensional space.
All the examples of monoidal categories given above are also braided monoidal
categories. Indeed, many mathematicians would shamelessly say that given vec-
tor spaces V and W , the tensor product V ⊗ W is ‘equal to’ the tensor product
W ⊗ V . But this is not really true; if you examine the fine print you will see
that they are just isomorphic, via this braiding:
BV,W : v ⊗ w 7→ w ⊗ v.
Actually, all the examples above are not just braided but also ‘symmetric’
monoidal categories. This means that if you switch two things and then switch
them again, you get back where you started:
Bx,y By,x = 1x⊗y .
19
Because all the braided monoidal categories Mac Lane knew satisfied this extra
axiom, he only considered symmetric monoidal categories. In diagrams, this
extra axiom says that:
x y x y
=
In 4 or more dimensions, any knot can be untied by just this sort of process.
Thus, the string diagrams for symmetric monoidal categories should really be
drawn in 4 or more dimensions! But, we can cheat and draw them in the plane,
as we have above.
It is worth taking a look at Mac Lane’s precise definitions, since they are a
bit subtler than our summary suggests, and these subtleties are actually very
interesting.
First, he demanded that a monoidal category have a unit for the tensor
product, which he call the ‘unit object’, or ‘1’. For example, the unit for tensor
product in Vect is the ground field, while the unit for the Cartesian product in
Set is the one-element set. (Which one-element set? Choose your favorite one!)
Second, Mac Lane did not demand that the tensor product be associative
‘on the nose’:
(x ⊗ y) ⊗ z = x ⊗ (y ⊗ z),
but only up a specified isomorphism called the ‘associator’:
ax,y,z : (x ⊗ y) ⊗ z → x ⊗ (y ⊗ z).
Similarly, he didn’t demand that 1 act as the unit for the tensor product ‘on the
nose’, but only up to specified isomorphisms called the ‘left and right unitors’:
ℓx : 1 ⊗ x → x, rx : x ⊗ 1 → x.
The reason is that in real life, it is usually too much to expect equations between
objects in a category: usually we just have isomorphisms, and this is good
enough! Indeed this is a basic moral of category theory: equations between
objects are bad; we should instead specify isomorphisms.
Third, and most subtly of all, Mac Lane demanded that the associator and
left and right unitors satisfy certain ‘coherence laws’, which let us work with
them as smoothly as if they were equations. These laws are called the pentagon
and triangle identities.
Here is the actual definition. A monoidal category consists of:
• a category M .
20
• a functor called the tensor product ⊗ : M × M → M , where we write

⊗(x, y) = x ⊗ y and ⊗(f, g) = f ⊗ g for objects x, y ∈ M and morphisms
f, g in M .
• an object called the identity object 1 ∈ M .
• natural isomorphisms called the associator:
ax,y,z : (x ⊗ y) ⊗ z → x ⊗ (y ⊗ z),
the left unit law:

ℓx : 1 ⊗ x → x,
and the right unit law:
rx : x ⊗ 1 → x.
such that the following diagrams commute for all objects w, x, y, z ∈ M :

• the pentagon identity:
((w ⊗ x) ⊗ y) ⊗ z
HH
vv HH
vvv HH
vv HH
aw,x,y ⊗1zvv v HHaw⊗x,y,z
v HH
v HH
v vv HH
v v HH
v v HH
{vv H#
(w ⊗ (x ⊗ y)) ⊗ z (w ⊗ x) ⊗ (y ⊗ z)
))
))
))
))
))
)) a
aw,x⊗y,z )
)) w,x,y⊗z
))
))
))
))

w ⊗ ((x ⊗ y) ⊗ z) 1w ⊗ax,y,z / w ⊗ (x ⊗ (y ⊗ z))
governing the associator.
• the triangle identity:

ax,1,y
(x ⊗ 1) ⊗ y / x ⊗ (1 ⊗ y)
LLL rr
LLL rrr
L
rx ⊗1y LL rr1 ⊗ℓ
L& xrrr x y
x⊗y
21
governing the left and right unitors.

The pentagon and triangle identities are the least obvious part of this definition—
but also the truly brilliant part. The point of the pentagon identity is that
when we have a tensor product of four objects, there are five ways to parenthe-
size it, and at first glance the associator gives two different isomorphisms from
w ⊗ (x⊗ (y ⊗ z)) to ((w ⊗ x)⊗ y)⊗ z. The pentagon identity says these are in fact
the same! Of course when we have tensor products of even more objects there
are even more ways to parenthesize them, and even more isomorphisms between
them built from the associator. However, Mac Lane showed that the pentagon
identity implies these isomorphisms are all the same. If we also assume the
triangle identity, all isomorphisms with the same source and target built from
the associator, left and right unit laws are equal.
In fact, the pentagon was also introduced in 1963 by James Stasheff [23], as
part of an infinite sequence of polytopes called ‘associahedra’. Stasheff defined
a concept of ‘A∞ -space’, which is roughly a topological space having a product
that is associative up to homotopy, where this homotopy satisfies the pentagon
identity up homotopy, that homotopy satisfies yet another identity up to homo-
topy, and so on, ad infinitum. The nth of these identities is described by the
n-dimensional associahedron. The first identity is just the associative law, which
plays a crucial role in the definition of monoid: a set with associative product
and identity element. Mac Lane realized that the second, the pentagon identity,
should play a similar role in the definition of monoidal category. The higher
ones show up in the theory of monoidal bicategories, monoidal tricategories and
so on.
With the concept of monoidal category in hand, one can define a braided
monoidal category to consist of:
• a monoidal category M , and
• a natural isomorphism called the braiding:
Bx,y : x ⊗ y → y ⊗ x.
such that these two diagrams commute, called the hexagon identities:
Bx,y ⊗z
(x ⊗ y) ⊗ z / (y ⊗ x) ⊗ z
a−1 rr9 LLL
x,y,z
rrr LaLy,x,z
LLL
rrr %
x ⊗ (y ⊗ z) y ⊗ (x ⊗ z)
LLL r
LLL r rrr
Bx,y⊗z LL% r
yrr y⊗Bx,z
(y ⊗ z) ⊗ x ay,z,x
/ y ⊗ (z ⊗ x)
22
x⊗By,z
x ⊗ (y ⊗ z) / x ⊗ (z ⊗ y)
9 LLLa−1
ax,y,z rrr LLx,z,y
rr LLL
rrr %
(x ⊗ y) ⊗ z (x ⊗ z) ⊗ y
LLL rr
LLL rrr
Bx⊗y,z LL% yrrrBx,z ⊗y
z ⊗ (x ⊗ y) / (z ⊗ x) ⊗ y
a−1
z,x,y
The first hexagon equation says that switching the object x past y ⊗ z all at
once is the same as switching it past y and then past z (with some associators
thrown in to move the parentheses). The second one is similar: it says switching
x ⊗ y past z all at once is the same as doing it in two steps.
We define a symmetric monoidal category to be a braided monoidal
−1
category M for which the braiding satisfies Bx,y = By,x for all objects x and y.
A monoidal, braided monoidal, or symmetric monoidal category is called strict
if ax,y,z , ℓx , and rx are always identity morphisms. In this case we have
(x ⊗ y) ⊗ z = x ⊗ (y ⊗ z),
1 ⊗ x = x, x ⊗ 1 = x.
Mac Lane showed in a certain precise sense, every monoidal or symmetric
monoidal category is equivalent to a strict one. The same is true for braided
monoidal categories. However, the examples that turn up in nature, like Vect,
are rarely strict.
Lawvere (1963)
The famous category theorist F. William Lawvere began his graduate work un-
der Clifford Truesdell, an expert on ‘continuum mechanics’, that very practical
branch of classical field theory which deals with fluids, elastic bodies and the
like. In the process, Lawvere got very interested in the foundations of physics,
particularly the notion of ‘physical theory’, and his research took a very abstract
turn. Since Truesdell had worked with Eilenberg and Mac Lane during World
War II, he sent Lawvere to visit Eilenberg at Columbia University, and that is
where Lawvere wrote his thesis.
In 1963, Lawvere finished a thesis on ‘functorial semantics’ [24]. This is a
general framework for theories of mathematical or physical objects in which a
‘theory’ is described by a category C, and a ‘model’ of this theory is described
by a functor Z : C → D. Typically C and D are equipped with extra structure,
and Z is required to preserve this structure. The category D plays the role of
an ‘environment’ in which the models live; often we take D = Set.
Variants of this idea soon became important in topology, especially ‘PROPs’
and ‘operads’. In the late 1960’s and early 70’s, Mac Lane [25], Boardmann and
Vogt [26], May [27] and others used these variants to study ‘homotopy-coherent’
23
algebraic structures: that is, structures with operations satisfying laws only up
to homotopy, with the homotopies themselves obeying certain laws, but only
up to homotopy, ad infinitum. The easiest examples are Stasheff’s A∞ -spaces,
which we mentioned in the previous section. The laws governing A∞ -spaces
are encoded in ‘associahedra’ such as the pentagon. In later work, it was seen
that the associahedra form an operad. By the 90’s, operads had become very
important both in mathematical physics [28, 29] and the theory of n-categories
[30]. Unfortunately, explaining this line of work would take us far afield.
Other outgrowths of Lawvere’s vision of functorial semantics include the
definitions of ‘conformal field theory’ and ‘topological quantum field theory’,
propounded by Segal and Atiyah in the late 1980s. We will have much more to
say about these. In keeping with physics terminology, these later authors use the
word ‘theory’ for what Lawvere called a ‘model’: namely, a structure-preserving
functor Z : C → D. There is, however, a much more important difference.
Lawvere focused on classical physics, and took C and D to be categories with
cartesian products. Segal and Atiyah focused on quantum physics, and took C
and D to be symmetric monoidal categories, not necessarily cartesian.
Bénabou (1967)
In 1967 Bénabou [31] introduced the notion of a ‘bicategory’, or as it is some-
times now called, a ‘weak 2-category’. The idea is that besides objects and
morphisms, a bicategory has 2-morphisms going between morphisms, like this:
objects morphisms 2-morphisms

f
x f /•y
• x• • α
D•

g
In a bicategory we can compose morphisms as in an ordinary category, but also

we can compose 2-morphisms in two ways: vertically and horizontally:
f f f′
α
α
β

• /•
E • D• D•
β

g g g′
There are also identity morphisms and identity 2-morphisms, and various axioms
governing their behavior. Most importantly, the usual laws for composition
of morphisms—the left and right unit laws and associativity—hold only up to
specified 2-isomorphisms. (A 2-isomorphism is a 2-morphism that is invertible
with respect to vertical composition.) For example, given morphisms h : w → x,
24
g : x → y and f : y → z, we have a 2-isomorphism called the ‘associator’:
af,g,h : (f g)h → f (gh).
As in a monoidal category, this should satisfy the pentagon identity.

Bicategories are everywhere once you know how to look. For example, there
is a bicategory Cat in which:
• the objects are categories,
• the morphisms are functors,
• the 2-morphisms are natural transformations.
This example is unusual, because composition of morphisms happens to satisfy
the left and right unit laws and associativity on the nose, as equations. A more
typical example is Bimod, in which:
• the objects are rings,
• the morphisms from R to S are R − S-bimodules,
• the 2-morphisms are bimodule homomorphisms.
Here composition of morphisms is defined by tensoring: given an R−S-bimodule
M and an S − T -bimodule, we can tensor them over S to get an R − T -
bimodule. In this example the laws for composition hold only up to specified
2-isomorphisms.
Another class of examples comes from the fact that a monoidal category is
secretly a bicategory with one object! The correspondence involves a kind of
‘reindexing’ as shown in the following table:
Monoidal Category Bicategory

— objects
objects morphisms
morphisms 2-morphisms
tensor product of objects composite of morphisms
composite of morphisms vertical composite of 2-morphisms
tensor product of morphisms horizontal composite of 2-morphisms
In other words, to see a monoidal category as a bicategory with only one object,
we should call the objects of the monoidal category ‘morphisms’, and call its
morphisms ‘2-morphisms’.
A good example of this trick involves the monoidal category Vect. Start with
Bimod and pick out your favorite object, say the ring of complex numbers. Then
take all those bimodules of this ring that are complex vector spaces, and all the
bimodule homomorphisms between these. You now have a sub-bicategory with
just one object—or in other words, a monoidal category! This is Vect.
The fact that a monoidal category is secretly just a degenerate bicategory
eventually stimulated a lot of interest in higher categories: people began to
25
wonder what kinds of degenerate higher categories give rise to braided and
symmetric monoidal categories. The impatient reader can jump ahead to 1995,
when the pattern underlying all these monoidal structures and their higher-
dimensional analogs became more clear.
Penrose (1971)
In general relativity people had been using index-ridden expressions for a long
time. For example, suppose we have a binary product on a vector space V :
m : V ⊗ V → V.
A normal person would abbreviate m(v ⊗ w) as v · w and write the associative
law as
(u · v) · w = u · (v · w).
A mathematician might show off by writing
m(m ⊗ 1) = m(1 ⊗ m)
instead. But physicists would pick a basis ei of V and set
X ij
m(ei ⊗ ej ) = mk e k
k
or
m(ei ⊗ ej ) = mij
k e
k
for short, using the ‘Einstein summation convention’ to sum over any repeated
index that appears once as a superscript and once as a subscript. Then, they
would write the associative law as follows:
pk iq jk
mij
p ml = ml mq .
Mathematicians would mock them for this, but until Penrose came along there
was really no better completely general way to manipulate tensors. Indeed,
before Einstein introduced his summation convention in 1916, things were even
worse. He later joked to a friend [32]:
I have made a great discovery in mathematics; I have suppressed the
summation sign every time that the summation must be made over
an index which occurs twice....
In 1971, Penrose [33] introduced a new notation where tensors are drawn as
‘black boxes’, with superscripts corresponding to wires coming in from above,
and subscripts corresponding to wires going out from below. For example, he
might draw m : V ⊗ V → V as:
22
i 22 j
2
m
k
26
and the associative law as:

22 44 22
i 22 j k i 44 j 22 k
2 44 2
44
m = 44 m
//
p /
44 q

m m
l l
In this notation we sum over the indices labelling ‘internal wires’—by which we
mean wires that are the output of one box and an input of another. This is just
the Einstein summation convention in disguise: so the above picture is merely
an artistic way of drawing this:
pk iq jk
mij
p ml = ml mq .
But it has an enormous advantage: no ambiguity is introduced if we leave out

the indices, since the wires tell us how the tensors are hooked together:
22 44 22
22 44 22
2 44 2
44
m = 44 m
//
/ 44

m m
This is a more vivid way of writing the mathematician’s equation
m(m ⊗ 1V ) = m(1V ⊗ m)
because tensor products are written horizontally and composition vertically,

instead of trying to compress them into a single line of text.
In modern language, what Penrose had noticed here was that FinVect, the
category of finite-dimensional vector spaces and linear maps, is a symmetric
monoidal category, so we can draw morphisms in it using string diagrams. But
he probably wasn’t thinking about categories: he was probably more influenced
by the analogy to Feynman diagrams.
Indeed, Penrose’s pictures are very much like Feynman diagrams, but sim-
pler. Feynman diagrams are pictures of morphisms in the symmetric monoidal
category of positive-energy representations of the Poincaré group! It is amusing
that this complicated example was considered long before Vect. But that is how
it often works: simple ideas rise to consciousness only when difficult problems
make them necessary.
27
Penrose also considered some examples more complicated than FinVect but
simpler than full-fledged Feynman diagrams. For any compact Lie group K,
there is a symmetric monoidal category Rep(K). Here the objects are finite-
dimensional continuous unitary representations of K—that’s a bit of a mouthful,
so we will just call them ‘representations’. The morphisms are intertwining
operators between representations: that is, operators f : H → H ′ with
f (ρ(g)ψ) = ρ′ (g)f (ψ)
for all g ∈ K and ψ ∈ H, where ρ(g) is the unitary operator by which g acts on
H, and ρ′ (g) is the one by which g acts on H ′ . The category Rep(K) becomes
symmetric monoidal with the usual tensor product of group representations:
(ρ ⊗ ρ′ )(g) = ρ(g) ⊗ ρ(g ′ )
and the obvious braiding.

As a category, Rep(K) is easy to describe. Every object is a direct sum of
finitely many irreducible representations: that is, representations that are not
themselves a direct sum in a nontrivial way. So, if we pick a collection Ei of
irreducible representations, one from each isomorphism class, we can write any
object H as M
H∼ = H i ⊗ Ei
i
i
where the H is the finite-dimensional Hilbert space describing the multiplicity
with which the irreducible Ei appears in H:
H i = hom(Ei , H)
Then, we use Schur’s Lemma, which describes the morphisms between irre-
ducible representations:
• When i = j, the space hom(Ei , Ej ) is 1-dimensional: all morphisms from
Ei to Ej are multiples of the identity.
• When i 6= j, the space hom(Ei , Ej ) is 0-dimensional: all morphisms from
E to E ′ are zero.
So, every representation is a direct sum of irreducibles, and every morphism
between irreducibles is a multiple of the identity (possibly zero). Since compo-
sition is linear in each argument, this means there’s only one way composition
of morphisms can possibly work. So, the category is completely pinned down
as soon as we know the set of irreducible representations.
One nice thing about Rep(K) is that every object has a dual. If H is some
representation, the dual vector space H ∗ also becomes a representation, with
(ρ∗ (g)f )(ψ) = f (ρ(g)ψ)
for all f ∈ H ∗ , ψ ∈ H. In our string diagrams, we can use little arrows to

distinguish between H and H ∗ : a downwards-pointing arrow labelled by H
28
stands for the object H, while an upwards-pointing one stands for H ∗ . For
example, this:
O H
is the string diagram for the identity morphism 1H ∗ . This notation is meant
to remind us of Feynman’s idea of antiparticles as particles going backwards in
time.
The dual pairing
eH : H ∗ ⊗ H → C
f ⊗ v 7→ f (v)
is an intertwining operator, as is the operator
iH : C → H ⊗ H∗
c 7 → c 1H
where we think of 1H ∈ hom(H, H) as an element of H ⊗ H ∗ . We can draw
these operators as a ‘cup’:
H∗ ⊗ H
S
H H stands for eH

C
and a ‘cap’:
C
H
T
H stands for iH

H ⊗ H∗
Note that if no edges reach the bottom (or top) of a diagram, it describes a
morphism to (or from) the trivial representation of G on C—since this is the
tensor product of no representations.
The cup and cap satisfy the zig-zag identities:
P =
O P = O
29
These identities are easy to check. For example, the first zig-zag gives a mor-
phism from H to H which we can compute by feeding in a vector ψ ∈ H:
ψ
_
eH

P
H ei ⊗ e_i ⊗ ψ
iH

ei ⊗ ψ i = ψ
So indeed, this is the identity morphism. But, the beauty of these identities is
that they let us straighten out a portion of a string diagram as if it were actually
a piece of string! Algebra is becoming topology.
Furthermore, we have:
H O
= dim(H)
O
This requires a little explanation. A ‘closed’ diagram—one with no edges coming

in and no edges coming out—denotes an intertwining operator from the trivial
representation to itself. Such a thing is just multiplication by some number.
The equation above says the operator on the left is multiplication by dim(H).
We can check this as follows:
1_

H O
ei ⊗
_ ei

i
O ei ⊗
_e

δii = dim(H)
So, a loop gives a dimension. This explains a big problem that plagues Feynman
diagrams in quantum field theory—namely, the ‘divergences’ or ‘infinities’ that
show up in diagrams containing loops, like this:
M

30
or more subtly, like this:

O
O
P
O
O
These infinities come from the fact that most positive-energy representations of
the Poincaré group are infinite-dimensional. The reason is that this group is
noncompact. For a compact Lie group, all the irreducible continuous represen-
tations are finite-dimensional.
So far we have been discussing representations of compact Lie groups quite
generally. In his theory of ‘spin networks’ [34, 35], Penrose worked out all the
details for SU(2): the group of 2 × 2 unitary complex matrices with determinant
1. This group is important because it is the universal cover of the 3d rotation
group. This lets us handle particles like the electron, which doesn’t come back
to its original state after one full turn—but does after two!
The group SU(2) is the subgroup of the Poincaré group whose corresponding
observables are the components of angular momentum. Unlike the Poincaré
group, it is compact. As already mentioned, we can specify an irreducible
positive-energy representation of the Poincaré group by choosing a mass m ≥ 0,
a spin j = 0, 12 , 1, 32 , . . . and sometimes a little extra data. Irreducible unitary
representations of SU(2) are simpler: for these, we just need to choose a spin.
The group SU(2) has one irreducible unitary representation of each dimension.
Physicists call the representation of dimension 2j + 1 the ‘spin-j’ representation,
or simply ‘j’ for short.
Every representation of SU(2) is isomorphic to its dual, so we can pick an
isomorphism
♯ : j → j∗
for each j. Using this, we can stop writing little arrows on our string diagrams.
For example, we get a new ‘cup’
j⊗j
j j ♯⊗1

j∗ ⊗ j
ej

C
and similarly a new cap. These satisfy an interesting relation:
j j
= (−1)2j
31
Physically, this means that when we give a spin-j particle a full turn, its state
transforms trivially when j is an integer:
ψ 7→ ψ
but it picks up a sign when j is an integer plus 21 :
ψ 7→ −ψ.
Particles of the former sort are called bosons; those of the latter sort are called
fermions.
The funny minus sign for fermions also shows up when we build a loop with
our new cup and cap:
= (−1)2j (2j + 1)
We get, not the usual dimension of the spin-j representation, but the dimension
times a sign depending on whether this representation is bosonic or fermionic!
This is sometimes called the superdimension, since its full explanation involves
what physicists call ‘supersymmetry’. Alas, we have no time to discuss this here:
we must hasten on to Penrose’s theory of spin networks!
Spin networks are a nice notation for morphisms between tensor products of
irreducible representations of SU(2). The key underlying fact is that:
j⊗k ∼
= |j − k| ⊕ |j − k| + 1 ⊕ · · · ⊕ j + k
Thus, the space of intertwining operators hom(j ⊗ k, l) has dimension 1 or 0

depending on whether or not l appears in this direct sum. We say the triple
(j, k, l) is admissible when this space has dimension 1. This happens when the
triangle inequalities are satisfied:
|j − k| ≤ l ≤ j + k
and also j + k + l ∈ Z.
For any admissible triple (j, k, l) we can choose a nonzero intertwining op-
erator from j ⊗ k to l, which we draw as follows:
??
j ???? k
?
•
l
32
Using the fact that a closed diagram gives a number, we can normalize these
intertwining operators so that the ‘theta network’ takes a convenient value, say:
j
•
k • = 1
l
When the triple (j, k, l) is not admissible, we define
??
j ???? k
?
•
l
to be the zero operator, so that
j
•
k • = 0
ℓ
We can then build more complicated intertwining operators by composing
and tensoring the ones we have described so far. For example, this diagram
shows an intertwining operator from the representation 2 ⊗ 23 ⊗ 1 to the repre-
sentation 25 ⊗ 2: 11
2 111 32 1
•'
'
5 ''
2 '
•'
'' 7
'' 2
•:
5
:::
2
:2
A diagram of this sort is called a ‘spin network’. The resemblance to a Feynman
diagram is evident. There is a category where the morphisms are spin networks,
and a functor from this category to Rep(SU(2)). A spin network with no edges
coming in from the top and no edges coming out at the bottom is called closed.
A closed spin network determines an intertwining operator from the trivial rep-
resentation of SU(2) to itself, and thus a complex number. For more details, see
the paper by Major [36].
Penrose noted that spin networks satisfy a bunch of interesting rules. For
example, we can deform a spin network in various ways without changing the
33
operator it describes. We have already seen the zig-zag identity, which is an

example of this. Other rules involve changing the topology of the spin network.
The most important of these is the binor identity for the spin- 21 representation:
1 1 1 1 1 1
2 2 2 2 2 2
= +
1 1
2 2
We can use this to prove something we have already seen:

1 1 1 1
2 2 2 2
= + = −
Physically, this says that turning a spin- 21 particle around 360 degrees multiplies
its state by −1.
There are also interesting rules involving the spin-1 representation, which
imply some highly nonobvious results. For example, every trivalent planar graph
with no edge-loops and all edges labelled by the spin-1 representation:
1 1
##
•
1 •
# •
1 ## 1 1 •
• 1 •
1 1
1
• YYYYYY
1 1 1
YY•
•
• 1
1
evaluates to a nonzero number [37]. But, Penrose showed this fact is equivalent
to the four-color theorem!
By now, Penrose’s diagrammatic approach to the finite-dimensional repre-
sentations of SU(2) has been generalized to many compact simple Lie groups.
A good treatment of this material is the free online book by Cvitanović [38].
His book includes a brief history of diagrammatic methods that makes a nice
companion to the present paper. Much of the work in his book was done in the
1970’s. However, the huge burst of work on diagrammatic methods for algebra
came later, in the 1980’s, with the advent of ‘quantum groups’.
Ponzano–Regge (1968)
Sometimes history turns around and goes back in time, like an antiparticle. This
seems like the only sensible explanation of the revolutionary work of Ponzano
34
and Regge [39], who applied Penrose’s theory of spin networks before it was
invented to relate tetrahedron-shaped spin networks to gravity in 3 dimensional
spacetime. Their work eventually led to a theory called the Ponzano–Regge
model, which allows for an exact solution of many problems in 3d quantum
gravity [40].
In fact, Ponzano and Regge’s paper on this topic appeared in the proceedings
of a conference on spectroscopy, because the 6j symbol is important in chemistry.
But for our purposes, the 6j symbol is just the number obtained by evaluating
this spin network:
q
b b
i
j
b
l k
p
b
depending on six spins i, j, k, l, p, q.

In the Ponzano–Regge model of 3d quantum gravity, spacetime is made of
tetrahedra, and we label the edges of tetrahedra with spins to specify their
lengths. To compute the amplitude for spacetime to have a particular shape,
we multiply a bunch of amplitudes (that is, complex numbers): one for each
tetrahedron, one for each triangle, and one for each edge. The most interesting
ingredient in this recipe is the amplitude for a tetrahedron. This is given by the
6j symbol.
But, we have to be a bit careful! Starting from a tetrahedron whose edge
lengths are given by spins:
l
j
k
we compute its amplitude using the ‘Poincaré dual’ spin network, which has:
• one vertex at the center of each face of the original tetrahedron;
• one edge crossing each edge of the original tetrahedron.
35
It looks like this:
l
j
k
Its edges inherit spin labels from the edges of the original tetrahedron:
q
b b
i
j
b
l k
p
b
Voilà! The 6j symbol!

It is easy to get confused, since the Poincaré dual of a tetrahedron just hap-
pens to be another tetrahedron. But, there are good reasons for this dualization
process. For example, the 6j symbol vanishes if the spins labelling three edges
meeting at a vertex violate the triangle inequalities, because then these spins
will be ‘inadmissible’. For example, we need
|i − j| ≤ p ≤ i + j
or the intertwining operator

??
??
i ?? j
?
•
will vanish, forcing the 6j symbols to vanish as well. But in the original tetra-
hedron, these spins label the three sides of a triangle:
33
33
i 33 j
33
33
3
p
So, the amplitude for a tetrahedron vanishes if it contains a triangle that violates
the triangle inequalities!
36
This is exciting because it suggests that the representations of SU(2) some-

how know about the geometry of tetrahedra. Indeed, there are other ways for
a tetrahedron to be ‘impossible’ besides having edge lengths that violate the
triangle inequalities. The 6j symbol does not vanish for all these tetrahedra,
but it is exponentially damped—very much as a particle in quantum mechanics
can tunnel through barriers that would be impenetrable classically, but with an
amplitude that decays exponentially with the width of the barrier.
In fact the relation between Rep(SU(2)) and 3-dimensional geometry goes
much deeper. Regge and Ponzano found an excellent asymptotic formula for
the 6j symbol that depends entirely on geometrically interesting aspects of
the corresponding tetrahedron: its volume, the dihedral angles of its edges,
and so on. But, what is truly amazing is that this asymptotic formula also
matches what one would want from a theory of quantum gravity in 3 dimensional
spacetime!
More precisely, the Ponzano–Regge model is a theory of ‘Riemannian’ quan-
tum gravity in 3 dimensions. Gravity in our universe is described with a
Lorentzian metric on 4-dimensional spacetime, where each tangent space has
the Lorentz group acting on it. But, we can imagine gravity in a universe where
spacetime is 3-dimensional and the metric is Riemannian, so each tangent space
has the rotation group SO(3) acting on it. The quantum description of gravity
in this universe should involve the double cover of this group, SU(2) — essen-
tially because it should describe not just how particles of integer spin transform
as they move along paths, but also particles of half-integer spin. And it seems
the Ponzano–Regge model is the right theory to do this.
A rigorous proof of Ponzano and Regge’s asymptotic formula was given only
in 1999, by Justin Roberts [41]. Physicists are still finding wonderful surprises
in the Ponzano–Regge model. For example, if we study it on a 3-manifold with
a Feynman diagram removed, with edges labelled by suitable representations, it
describes not only ‘pure’ quantum gravity but also matter! The series of papers
by Freidel and Louapre explain this in detail [42].
Besides its meaning for geometry and physics, the 6j symbol also has a purely
category-theoretic significance: it is a concrete description of the associator in
Rep(SU(2)). The associator gives a linear operator
ai,j,k : (i ⊗ j) ⊗ k → i ⊗ (j ⊗ k).
The 6j symbol is a way of expressing this operator as a bunch of numbers. The

idea is to use our basic intertwining operators to construct operators
S : (i ⊗ j) ⊗ k → l, T : l → i ⊗ (j ⊗ k),
37
namely:
22
i 222 j k
22 l
•//
p // //
•
//
S = T = //
•
//q
•2
22
l 22
i j 22k
Using the associator to bridge the gap between (i ⊗ j) ⊗ k and i ⊗ (j ⊗ k), we

can compose S and T and take the trace of the resulting operator, obtaining a
number. These numbers encode everything there is to know about the associator
in the monoidal category Rep(SU(2)). Moreover, these numbers are just the 6j
symbols: q
b b
i
j
tr(T ai,j,k S) =
b
l k
p
b
This can be proved by gluing the pictures for S and T together and warping
the resulting spin network until it looks like a tetrahedron! We leave this as an
exercise for the reader.
The upshot is a remarkable and mysterious fact: the associator in the
monoidal category of representations of SU(2) encodes information about 3-
dimensional quantum gravity! This fact will become less mysterious when we
see that 3-dimensional quantum gravity is almost a topological quantum field
theory, or TQFT. In our discussion of Barrett and Westbury’s 1992 paper on
TQFTs, we will see that a large class of 3d TQFTs can be built from monoidal
categories.
Grothendieck (1983)
In his 600–page letter entitled Pursuing Stacks, Grothendieck fantasized about
n-categories for higher n—even n = ∞—and their relation to homotopy theory
[43]. The rough idea of an ∞-category is that it should be a generalization of
a category which has objects, morphisms, 2-morphisms and so on forever. In
the fully general, ‘weak’ ∞-categories, all the laws governing composition of j-
morphisms should hold only up to a specified (j + 1)-morphisms, which in turn
satisfy laws of their own, but only up to specified (j + 2)-morphisms, and so on.
Furthermore, all these higher morphisms which play the role of ‘laws’ should
be equivalences—where a k-morphism is an ‘equivalence’ if it is invertible up to
equivalence. The circularity here is not necessarily vicious, but it hints at how
tricky ∞-categories can be.
38
Grothendieck believed that among the weak ∞-categories there should be

a special class, the ‘weak ∞-groupoids’, in which all j-morphisms (j ≥ 1) are
equivalences. He also believed that every space X should have a weak ∞-
groupoid Π∞ (X) called its ‘fundamental ∞-groupoid’, in which:
• the objects are points of X,
• the morphisms are paths in X,
• the 2-morphisms are paths of paths in X,
• the 3-morphisms are paths of paths of paths in X,
• etc.
Moreover, Π∞ (X) should be a complete invariant of the homotopy type of X, at
least for nice spaces like CW complexes. In other words, two nice spaces should
have equivalent fundamental ∞-groupoids if and only if they are homotopy
equivalent.
The above sketch of Grothendieck’s dream is phrased in terms of a ‘glob-
ular’ approach to n-categories, where the n-morphisms are modeled after n-
dimensional discs:
objects morphisms 2-morphisms 3-morphisms ···

• • /• • D• • _*4
E• Globes

% y
However, he also imagined other approaches based on j-morphisms with differ-

ent shapes, such as simplices:

E •)) DD
F • 22 )) DDD
22 )) !2 •
• • /• 22 D Simplices
• VVVVV ))
• / • V+ •
In fact, simplicial weak ∞-groupoids had already been developed in a 1957

paper by Kan [44]; these are now called ‘Kan complexes’. In this framework
Π∞ (X) is indeed a complete invariant of the homotopy type of any nice space
X. So, the real challenge is to define weak ∞-categories in the simplicial and
other approaches, and then define weak ∞-groupoids as special cases of these,
and prove their relation to homotopy theory.
39
Great progress towards fulfilling Grothendieck’s dream has been made in

recent years. We cannot possibly do justice to the enormous body of work
involved, so we simply offer a quick thumbnail sketch. Starting around 1977,
Street began developing a simplicial approach to ∞-categories [45, 46] based
on ideas from the physicist Roberts [47]. Thanks in large part to the recently
published work of Verity, this approach has begun to really take off [48, 49, 50].
In 1995, Baez and Dolan initiated another approach to weak n-categories,
the ‘opetopic’ approach [51]:

•E / ... •/ 22
2 • /: • • / •+
• • /• •O • I uuuu ++ _*4
I + Opetopes
u / / •
• • •
• /
•
The idea here is that an (n + 1)-dimensional opetope describes a way of gluing

together n-dimensional opetopes. The opetopic approach was corrected and
clarified by various authors [52, 53, 54, 55, 56, 57], and by now it has been
developed by Makkai [58] into a full-fledged foundation for mathematics. We
have already mentioned how in category theory it is considered a mistake to
assert equations between objects: instead, one should specify an isomorphism
between them. Similarly, in n-category theory it is a mistake to assert an
equation between j-morphisms for any j < n: one should instead specify an
equivalence. In Makkai’s approach to the foundations of mathematics based on
weak ∞-categories, equality plays no role, so this mistake is impossible to make.
Instead of stating equations one must always specify equivalences.
Also starting around 1995, Tamsamani [59] and Simpson [61, 62, 63, 64]
developed a ‘multisimplicial’ approach to weak n-categories. In a 1998 paper,
Batanin [65, 66] initiated a globular approach to weak ∞-categories. Penon
[67] gave a related, very compact definition of ∞-category, which was later
improved by Batanin, Cheng and Makkai [68, 69]. There is also a topologically
motivated approach using operads due to Trimble [70], which was studied and
generalized by Cheng and Gurski [71, 72]. Yet another theory is due to Joyal,
with contributions by Berger [73, 74].
This great diversity of approaches raises the question of when two definitions
of n-category count as ‘equivalent’. In Pursuing Stacks, Grothendieck proposed
the following answer. Suppose that for all n we have two different definitions
of weak n-category, say ‘n-category1 ’ and ‘n-category2 ’. Then we should try to
construct the (n + 1)-category1 of all n-categories1 and the (n + 1)-category1 of
all n-categories2 and see if these are equivalent as objects of the (n+2)-category1
of all (n + 1)-categories1 . If so, we may say the two definitions are equivalent
as seen from the viewpoint of the first definition.
Of course, this strategy for comparing definitions of weak n-category requires
a lot of work. Nobody has carried it out for any pair of significantly different
definitions. There is also some freedom of choice involved in constructing the
40
two (n + 1)-categories1 in question. One should do it in a ‘reasonable’ way, but

what does that mean? And what if we get a different answer when we reverse
the roles of the two definitions?
A somewhat less strenuous strategy for comparing definitions is suggested
by homotopy theory. Many different approaches to homotopy theory are in use,
and though superficially very different, there is by now a well-understood sense
in which they are fundamentally the same. Different approaches use objects
from different categories to represent topological spaces, or more precisely, the
homotopy-invariant information in topological spaces, called their ‘homotopy
types’. These categories are not equivalent, but each one is equipped with a
class of morphisms called ‘weak equivalences’, which play the role of homotopy
equivalences. Given a category C equipped with a specified class of weak equiv-
alences, under mild assumptions one can throw in inverses for these morphisms
and obtain a category called the ‘homotopy category’ Ho(C). Two categories
with specified equivalences may be considered the same for the purposes of ho-
motopy theory if their homotopy categories are equivalent in the usual sense of
category theory. The same strategy —or more sophisticated variants—can be
applied to comparing definitions of n-category, so long as one can construct a
category of n-categories.
Starting around 2000, work began on comparing different approaches to n-
category theory [56, 71, 74, 76, 75]. There has also been significant progress
towards achieving Grothendieck’s dream of relating weak n-groupoids to homo-
topy theory [60, 77, 78, 79, 80, 81, 82]. But n-category theory is still far from
mature. This is one reason the present paper is just a ‘prehistory’.
Luckily, Leinster has written a survey of definitions of n-category [83], and
also a textbook on the role of operads and their generalizations in higher cate-
gory theory [84]. Cheng and Lauda have prepared an ‘illustrated guidebook’ of
higher categories, for those who like to visualize things [85]. The forthcoming
book by Baez and May [86] provides more background for readers who want to
learn the subject. And for applications to algebra, geometry and physics, try the
conference proceedings edited by Getzler and Kapranov [87] and by Davydov et
al [88].
String theory (1980’s)

In the 1980’s there was a huge outburst of work on string theory. There is no
way to summarize it all here, so we shall content ourselves with a few remarks
about its relation to n-categorical physics. For a general overview the reader
can start with the introductory text by Zweibach [89], and then turn to the
book by Green, Schwarz and Witten [90], which was written in the 1980s, or
the book by Polchinski [91], which covers more recent developments.
String theory goes beyond ordinary quantum field theory by replacing 0-
dimensional point particles by 1-dimensional objects: either circles, called ‘closed
strings’, or intervals, called ‘open strings’. So, in string theory, the essentially
1-dimensional Feynman diagrams depicting worldlines of particles are replaced
by 2-dimensional diagrams depicting ‘string worldsheets’:
41
??
??
? ?

??
7→
O
O
O
O
This is a hint that as we pass from ordinary quantum field theory to string
theory, the mathematics of categories is replaced by the mathematics of bicate-
gories. However, this hint took a while to be recognized.
To compute an operator from a Feynman diagram, only the topology of
the diagram matters, including the specification of which edges are inputs and
which are outputs. In string theory we need to equip the string worldsheet with
a conformal structure, which is a recipe for measuring angles. More precisely:
a conformal structure on a surface is an orientation together with an equiv-
alence class of Riemannian metrics, where two metrics counts as equivalent if
they give the same answers whenever we use them to compute angles between
tangent vectors.
A conformal structure is also precisely what we need to do complex analysis
on a surface. The power of complex analysis is what makes string theory so
much more tractable than theories of higher-dimensional membranes.
Joyal–Street (1985)
Around 1985, Joyal and Street introduced braided monoidal categories [92].
The story is nicely told in Street’s Conspectus [1], so here we focus on the
mathematics.
As we have seen, braided monoidal categories are just like Mac Lane’s sym-
metric monoidal categories, but without the law
−1
Bx,y = By,x .
The point of dropping this law becomes clear if we draw the isomorphism
Bx,y : x ⊗ y → y ⊗ x as a little braid:
x y
Then its inverse is naturally drawn as
y x

42
−1
since then the equation Bx,y Bx,y = 1 makes topological sense:
y x y x

=

y x y x
−1
and similarly for Bx,y Bx,y = 1:
x y x y

=

x y x y
In fact, these equations are familiar in knot theory, where they describe ways
of changing a 2-dimensional picture of a knot (or braid, or tangle) without
changing it as a 3-dimensional topological entity. Both these equations are
called the second Reidemeister move.
−1
On the other hand, the law Bx,y = By,x would be drawn as
x y x y
=

and this is not a valid move in knot theory: in fact, using this move all knots
become trivial. So, it make some sense to drop it, and this is just what the
definition of braided monoidal category does.
Joyal and Street constructed a very important braided monoidal category
called Braid. Every object in this category is a tensor product of copies of a
special object x, which we draw as a point. So, we draw the object x⊗n as a row
of n points. The unit for the tensor product, I = x⊗0 , is drawn as a blank space.
All the morphisms in Braid are endomorphisms: they go from an object to
itself. In particular, a morphism f : x⊗n → x⊗n is an n-strand braid:

and composition is defined by stacking one braid on top of another. We tensor

morphisms in Braid by setting braids side by side. The braiding is defined in
43
an obvious way: for example, the braiding
B2,3 : x⊗2 ⊗ x⊗3 → x⊗3 ⊗ x⊗2
looks like this:

Joyal and Street showed that Braid is the ‘free braided monoidal category on
one object’. This and other results of theirs justify the use of string diagrams as a
technique for doing calculations in braided monoidal categories. They published
a paper on this in 1991, aptly titled The Geometry of Tensor Calculus [19].
Let us explain more precisely what it means that Braid is the free braided
monoidal category on one object. For starters, Braid is a braided monoidal
category containing a special object x: the point. But when we say Braid is
the free braided monoidal category on this object, we are saying much more.
Intuitively, this means two things. First, every object and morphism in Braid
can be built from 1 using operations that are part of the definition of ‘braided
monoidal category’. Second, every equation that holds in Braid follows from
the definition of ‘braided monoidal category’.
To make this precise, consider a simpler but related example. The group of
integers Z is the free group on one element, namely the number 1. Intuitively
speaking this means that every integer can be built from the integer 1 using
operations built into the definition of ‘group’, and every equation that holds in
Z follows from the definition of ‘group’. For example, (1 + 1) + 1 = 1 + (1 + 1)
follows from the associative law.
To make these intuitions precise it is good to use the idea of a ‘universal
property’. Namely: for any group G containing an element g there exists a
unique homomorphism
ρ: Z → G
such that
ρ(1) = g.
The uniqueness clause here says that every integer is built from 1 using the
group operations: that is why knowing what ρ does to 1 determines ρ uniquely.
The existence clause says that every equation between integers follows from
the definition of a group: if there were extra equations, these would block the
existence of homomorphisms to groups where these equations failed to hold.
So, when we say that Braid is the ‘free’ braided monoidal category on the
object 1, we mean something roughly like this: for any braided monoidal cat-
egory C, and any object c ∈ C, there is a unique map of braided monoidal
categories
Z : Braid → C
44
such that
Z(x) = c.
This will not be not precise until we define a map of braided monoidal
categories. The correct concept is that of a ‘braided monoidal functor’. But we
also need to weaken the universal property. To say that Z is ‘unique’ means
that any two candidates sharing the desired property are equal. But this is
too strong: it is bad to demand equality between functors. Instead, we should
say that any two candidates are isomorphic. For this we need the concept of
‘braided monoidal natural isomorphism’.
Once we have these concepts in hand, the correct theorem is as follows. For
any braided monoidal category C, and any object x ∈ C, there exists a braided
monoidal functor
Z : Braid → C
such that
Z(x) = c.
Moreover, given two such braided monoidal functors, there is a braided monoidal
natural isomorphism between them.
Now we just need to define the necessary concepts. The definitions are a bit
scary at first sight, but they illustrate the idea of weakening: that is, replacing
equations by isomorphisms which satisfy equations of their own. They will also
be needed for the definition of ‘topological quantum field theories’, which we
will present in our discussion of Atiyah’s 1988 paper.
To begin with, a functor F : C → D between monoidal categories is monoidal
if it is equipped with:
• a natural isomorphism Φx,y : F (x) ⊗ F (y) → F (x ⊗ y), and
• an isomorphism φ : 1D → F (1C )
such that:
• the following diagram commutes for any objects x, y, z ∈ C:
Φx,y ⊗1F (z) Φx⊗y,z
(F (x) ⊗ F (y)) ⊗ F (z) - F (x ⊗ y) ⊗ F (z) - F ((x ⊗ y) ⊗ z)
aF (x),F (y),F (z) F (ax,y,z )
? ?
F (x) ⊗ (F (y) ⊗ F (z)) - F (x) ⊗ F (y ⊗ z) - F (x ⊗ (y ⊗ z))
1F (x) ⊗Φy,z Φx,y⊗z
45
• the following diagrams commute for any object x ∈ C:

ℓF (x)
1 ⊗ F (x) - F (x)
6
φ⊗1F (x) F (ℓx )
?
F (1) ⊗ F (x) - F (1 ⊗ x)
Φ1,x
rF (x)
F (x) ⊗ 1 - F (x)
6
1F (x) ⊗φ F (rx )
?
F (x) ⊗ F (1) - F (x ⊗ 1)
Φx,1
Note that we do not require F to preserve the tensor product or unit ‘on the
nose’. Instead, it is enough that it preserve them up to specified isomorphisms,
which must in turn satisfy some plausible equations called ‘coherence laws’.
This is typical of weakening.
A functor F : C → D between braided monoidal categories is braided
monoidal if it is monoidal and it makes the following diagram commute for
all x, y ∈ C:
BF (x),F (y)
F (x) ⊗ F (y) - F (y) ⊗ F (x)
Φx,y Φy,x
? ?
F (x ⊗ y) - F (y ⊗ x)
F (Bx,y )
This condition says that F preserves the braiding as best it can, given the
fact that it only preserves tensor products up to a specified isomorphism. A
symmetric monoidal functor is just a braided monoidal functor that happens
to go between symmetric monoidal categories. No extra condition is involved
here.
Having defined monoidal, braided monoidal and symmetric monoidal func-
tors, let us next do the same for natural transformations. Recall that a monoidal
functor F : C → D is really a triple (F, Φ, φ) consisting of a functor F , a
natural isomorphism Φx,y : F (x) ⊗ F (y) → F (x ⊗ y), and an isomorphism
φ : 1D → F (1C ). Suppose that (F, Φ, φ) and (G, Γ, γ) are monoidal functors
from the monoidal category C to the monoidal category D. Then a natural
46
transformation α : F ⇒ G is monoidal if the diagrams
F (x) ⊗ F (y)
αx ⊗αy
- G(x) ⊗ G(y)
Φx,y Γx,y
? ?
F (x ⊗ y) - G(x ⊗ y)
αx⊗y
and
1D
Q
Q γ
Q
φ Q
Q
? Qs
Q
F (1C ) - G(1C )
α1C
commute. There are no extra condition required of braided monoidal or

symmetric monoidal natural transformations.
The reader, having suffered through these definitions, is entitled to see an
application besides Joyal and Street’s algebraic description of the category of
braids. At the end of our discussion of Mac Lane’s 1963 paper on monoidal
categories, we said that in a certain sense every monoidal category is equivalent
to a strict one. Now we can make this precise! Suppose C is a monoidal category.
Then there is a strict monoidal category D that is monoidally equivalent to
C. That is: there are monoidal functors F : C → D, G : D → C and monoidal
natural isomorphisms α : F G ⇒ 1D , β : GF ⇒ 1C .
This result allows us to work with strict monoidal categories, even though
most monoidal categories found in nature are not strict: we can take the
monoidal category we are studying and replace it by a monoidally equivalent
strict one. The same sort of result is true for braided monoidal and symmetric
monoidal categories.
A very similar result holds for bicategories: they are all equivalent to strict
2-categories: that is, bicategories where all the associators and unitors are
identity morphisms. However, the pattern breaks down when we get to tricate-
gories: not every tricategory is equivalent to a strict 3-category! At this point
the necessity for weakening becomes clear.
Jones (1985)
A knot is a circle smoothly embedded in R3 :
...
.... ....
... ` ....
........ .................................
......... ..

....@. ... ... ...
...... .. ......
......
... .................... ....$
.... .......p... ........ ..
.. ......
47
More generally, a link is a collection of disjoint knots. In topology, we consider

two links to be the same, or ‘isotopic’, if we can deform one smoothly without
its strands crossing until it looks like the other. Classifying links up to isotopy
is a challenging task that has spawned many interesting theorems and conjec-
tures. To prove these, topologists are always looking for link invariants: that is,
quantities they can compute from a link, which are equal on isotopic links.
In 1985, Jones [93] discovered a new link invariant, now called the ‘Jones
polynomial’. To everyone’s surprise he defined this using some mathematics
with no previously known connection to knot theory: the operator algebras
developed in the 1930s by Murray and von Neumann [10] as part of a general
formalism for quantum theory. Shortly thereafter, the Jones polynomial was
generalized by many authors obtaining a large family of so-called ‘quantum
invariants’ of links.
Of all these new link invariants, the easiest to explain is the ‘Kauffman
bracket’ [94]. The Kauffman bracket can be thought of as a simplified version of
the Jones polynomial. It is also a natural development of Penrose’s 1971 work
on spin networks [95].
As we have seen, Penrose gave a recipe for computing a number from any
spin network. The case relevant here is a spin network with vertices at all, with
every edge labelled by the spin 21 . For spin networks like this we can compute
the number by repeatedly using the binor identity:
= +
and this formula for the ‘unknot’:
= −2
The Kauffman bracket obeys modified versions of the above identities. These
involve a parameter that we will call q:
= q + q −1
and
= −(q 2 + q −2 )
48
Among knot theorists, identities of this sort are called ‘skein relations’.
Penrose’s original recipe is unable to detecting linking or knotting, since it
also satisfies this identity:
coming from the fact that Rep(SU(2)) is a symmetric monoidal category. The
Kauffman bracket arises from a more interesting braided monoidal category: the
category of representations of the ‘quantum group’ associated to SU(2). This
quantum group depends on a parameter q, which conventionally is related to
quantity we are calling q above by a mildly annoying formula. To keep our story
simple, we identify these two parameters.
When q = 1, the category of representations of the quantum group associated
to SU(2) reduces to Rep(SU(2)), and the Kauffman bracket reduces to Penrose’s
original recipe. At other values of q, this category is not symmetric, and the
Kauffman bracket detects linking and knotting.
In fact, all the quantum invariants of links discovered around this time turned
out to come from braided monoidal categories: namely, categories of representa-
tions of quantum groups. When q = 1, these quantum groups reduce to ordinary
groups, their categories of representations become symmetric, and the quantum
invariants of links become boring.
A basic result in knot theory says that given diagrams of two isotopic links,
we can get from one to the other by warping the page on which they are drawn,
together with a finite sequence of steps where we change a small portion of the
diagram. There are three such steps, called the first Reidemeister move:
the second Reidemeister move:
and the third Reidemeister move:

49
=
%%
%%
% %
Kauffman gave a beautiful purely diagrammatic argument that his bracket was
invariant under the second and third Reidemeister moves. We leave it as a
challenge to the reader to find this argument, which looks very simple after
one has seen it. On the other hand, the bracket is not invariant under the
first Reidemeister move. But, it transforms in a simple way, as this calculation
shows:
= q + q −1 = −q −3
where we used the skein relations and did a little algebra. So, while the Kauff-
man bracket is not an isotopy invariant of links, it comes close: we shall later
see that it is an invariant of ‘framed’ links, made from ribbons. And with a bit
of tweaking, it gives the Jones polynomial, which is an isotopy invariant.
This and other work by Kauffman helped elevate string diagram techniques
from a curiosity to a mainstay of modern mathematics. His book Knots and
Physics was especially influential in this respect [96]. Meanwhile, the work
of Jones led researchers towards a wealth of fascinating connections between
von Neumann algebras, higher categories, and quantum field theory in 2- and
3-dimensional spacetime.
Freyd–Yetter (1986)
Among the many quantum invariants of links that appeared after Jones poly-
nomial, one of the most interesting is the ‘HOMFLY-PT’ polynomial, which, it
later became clear, arises from the category of representations of the quantum
group associated to SU(n). This polynomial got its curious name because it
was independently discovered by many mathematicians, some of whom teamed
up to write a paper about it for the Bulletin of the American Mathematical
Society in 1985: Hoste, Ocneanu, Millet, Freyd, Lickorish and Yetter [97]. The
‘PT’ refers to Przytycki and Traczyk, who published separately [98].
Different authors of this paper took different approaches. Freyd and Yetter’s
approach is particularly germane to our story because they used a category
50
where morphisms are tangles. A ‘tangle’ is a generalization of a braid that

allows strands to double back, and also allows closed loops:
UU **
**
**
*
So, a link is just a tangle with no strands coming in on top, and none leaving
at the bottom. The advantage of tangles is that we can take a complicated link
and chop it into simple pieces, which are tangles.
Shortly after Freyd heard Street give a talk on braided monoidal categories
and the category of braids, Freyd and Yetter found a similar purely algebraic
description of the category of oriented tangles [99]. A tangle is ‘oriented’ if each
strand is equipped with a smooth nowhere vanishing field of tangent vectors,
which we can draw as little arrows. We have already seen what an orientation
is good for: it lets us distinguish between representations and their duals—or
in physics, particles and antiparticles.
There is a precisely defined but also intuitive notion of when two oriented
tangles count as the same: roughly speaking, whenever we can go from the
first to the second by smoothly moving the strands without moving their ends
or letting the strands cross. In this case we say these oriented tangles are
‘isotopic’.
The category of oriented tangles has isotopy classes of oriented tangles as
morphisms. We compose tangles by sticking one on top of the other. Just
like Joyal and Street’s category of braids, Tang is a braided monoidal category,
where we tensor tangles by placing them side by side, and the braiding is defined
using the fact that a braid is a special sort of tangle.
In fact, Freyd and Yetter gave a purely algebraic description of the category
of oriented tangles as a ‘compact’ braided monoidal category. Here a monoidal
category C is compact if every object x ∈ C has a dual: that is, an object x∗
together with morphisms called the unit:
1
T = ix

x ⊗ x∗
and the counit:

x∗ ⊗ x
S
= ex

1
51
satisfying the zig-zag identities:
P
=
O P = O
We have already seen these identities in our discussion of Penrose’s work. In-
deed, some classic examples of compact symmetric monoidal categories include
FinVect, where x∗ is the usual dual of the vector space x, and Rep(K) for any
compact Lie group K, where x∗ is the dual of the representation x. But the
zig-zag identities clearly hold in the category of oriented tangles, too, and this
example is braided but not symmetric.
There are some important subtleties that our sketch has overlooked so far.
For example, for any object x in a compact braided monoidal category, this
string diagram describes an isomorphism dx : x → x∗∗ :
x

x∗∗
But if we think of this diagram as an oriented tangle, it is isotopic to a straight
line. This suggests that dx should be an identity morphism. To implement
this idea, Freyd and Yetter used braided monoidal categories where each object
has a chosen dual, and this equation holds: x∗∗ = x. Then they imposed the
equation dx = 1x , which says that
=
This seems sensible, but it in category theory it is always dangerous to im-

pose equations between objects, like x∗∗ = x. And indeed, the danger becomes
52
clear when we remember that Penrose’s spin networks violate the above rule:
instead, they satisfy
j j
= (−1)2j+1
The Kauffman bracket violates the rule in an even more complicated way. As
mentioned in our discussion of Jones’ 1985 paper, the Kauffman bracket satisfies
= −q −3
So, while Freyd and Yetter’s theorem is correct, it needs some fine-tuning to
cover all the interesting examples.
For this reason, Street’s student Shum [100] considered tangles where each
strand is equipped with both an orientation and a framing — a nowhere van-
ishing smooth field of unit normal vectors. We can draw a framed tangle as
made of ribbons, where one edge of each ribbon is black, while the other is red.
The black edge is the actual tangle, while the normal vector field points from
the black edge to the red edge. But in string diagrams, we usually avoid drawing
the framing by using a standard choice: the blackboard framing, where the
unit normal vector points at right angles to the page, towards the reader.
There is an evident notion of when two framed oriented tangles count as
the same, or ‘isotopic’. Any such tangle is isotopic to one where we use the
blackboard framing, so we lose nothing by making this choice. And with this
choice, the following framed tangles are not isotopic:
6=
The problem is that if we think of these tangles as ribbons, and pull the left
one tight, it has a 360 degree twist in it.
53
What is the framing good for in physics? The picture above is the answer.
We can think of each tangle as a physical process involving particles. The pres-
ence of the framing means that the left-hand process is topologically different
than the right-hand process, in which a particle just sits there unchanged.
This is worth pondering in more detail. Consider the left-hand picture:
Reading this from top to bottom, it starts with a single particle. Then a virtual
particle-antiparticle pair is created on the left. Then the new virtual particle
and the original particle switch places by moving around each other clockwise.
Finally, the original particle and its antiparticle annihilate each other. So, this
is all about a particle that switches places with a copy of itself.
But we can also think of this picture as a ribbon. If we pull it tight, we
get a ribbon that is topologically equivalent—that is, isotopic. It has a 360◦
clockwise twist in it. This describes a particle that rotates a full turn:
So, as far as topology is concerned, we can express the concept of rotating a

single particle a full turn in terms of switching two identical particles—at least
in situations where creation and annihilation of particle-antiparticle pairs is
possible. This fact is quite remarkable. As emphasized by Feynman [101], it lies
at the heart of the famous ‘spin-statistics theorem’ in quantum field theory. We
have already seen that in theories of physics where spacetime is 4-dimensional,
the phase of a particle is multiplied by either 1 or −1 when we rotate it a full
turn: 1 for bosons, and −1 for fermions. The spin-statistics theorem says that
switching two identical copies of this particle has the same effect on their phase:
1 for bosons, −1 for fermions.
The story becomes even more interesting in theories of physics where space-
time is 3-dimensional. In this situation space is 2-dimensional, so we can distin-
guish between clockwise and counterclockwise rotations. Now the spin-statistics
theorem says that rotating a single particle a full turn clockwise gives the same
phase as switching two identical particles of this type by moving them around
each other clockwise. Rotating a particle a full turn clockwise need not have
the same effect as rotating it counterclockwise, so this phase need not be its
54
own inverse. In fact, it can be any unit complex number. This allows for ‘ex-
otic’ particles that are neither bosons nor fermions. In 1982, such particles were
dubbed anyons by Frank Wilczek [102].
Anyons are not just mathematical curiosities. Superconducting thin films
appear to be well described by theories in which the dimension of spacetime
is 3: two dimensions for the film, and one for time. In such films, particle-
like excitations arise, which act like anyons to a good approximation. The
presence of these ‘quasiparticles’ causes the film to respond in a surprising way to
magnetic fields when current is running through it. This is called the ‘fractional
quantum Hall effect’ [103].
In 1983, Robert Laughlin [104] published an explanation of the fractional
quantum Hall effect in terms of anyonic quasiparticles. He won the Nobel prize
for this work in 1998, along with Horst Störmer and Daniel Tsui, who observed
this effect in the lab [105]. By now we have an increasingly good understanding
of anyons in terms of a quantum field theory called Chern–Simons theory, which
also explains knot invariants such as the Kauffman bracket. For a bit more on
this, see our discussion of Witten’s 1989 paper on Chern–Simons theory.
But we are getting ahead of ourselves! Let us return to the work of Shum.
She constructed a category where the objects are finite collections of oriented
points in the unit square. By ‘oriented’ we mean that each point is labelled either
x or x∗ . We call a point labelled by x positively oriented, and one labelled
by x∗ negatively oriented. The morphisms in Shum’s category are isotopy
classes of framed oriented tangles. As usual, composition is defined by gluing
the top of one tangle to the bottom of the other. We shall call this category
1Tang2 . The reason for this curious notation is that the tangles themselves
have dimension 1, but they live in a space — or spacetime, if you prefer — of
dimension 1 + 2 = 3. The number 2 is called the ‘codimension’. It turns out
that varying these numbers leads to some very interesting patterns.
Shum’s theorem gives a purely algebraic description of 1Tang2 in terms of
‘ribbon categories’. We have already seen that in a compact braided monoidal
category C, every object x ∈ C comes equipped with an isomorphism to its
double dual, which we denoted dx : x → x∗∗ . A ribbon category is a compact
braided monoidal category where each object x is also equipped with another
isomorphism, cx : x∗∗ → x, which must satisfy a short list of axioms. We call
this a ‘ribbon structure’. Composing this ribbon structure with dx , we get an
isomorphism
bx = cx dx : x → x.
Now the point is that we can draw a string diagram for bx which is very much
55
like the diagram for dx , but with x as the output instead of x∗∗ :
x

x
This is the composite of dx , which we know how to draw, and cx , which we leave
invisible, since we do not know how to draw it.
In modern language, Shum’s theorem says that 1Tang2 is the ‘free ribbon
category on one object’, namely the positively oriented point, x. The definition
of ribbon category is designed to make it obvious that 1Tang2 is a ribbon cate-
gory. But in what sense is it ‘free on one object’ ? For this we define a ‘ribbon
functor’ to be a braided monoidal functor between ribbon categories that pre-
serves the ribbon structure. Then the statement is this. First, given any ribbon
category C and any object c ∈ C, there is a ribbon functor
Z : 1Tang2 → C
such that
Z(x) = c.
Second, Z is unique up to a braided monoidal natural isomorphism.
For a thorough account of Shum’s theorem and related results, see Yetter’s
book [106]. We emphasized some technical aspects of this theorem because
they are rather strange. As we shall see, the theme of n-categories ‘with duals’
becomes increasingly important as our history winds to its conclusion, but duals
remain a bit mysterious. Shum’s theorem is the first hint of this: to avoid
the equation between objects x∗∗ = x, it seems we are forced to introduce an
isomorphism cx : x∗∗ → x with no clear interpretation as a string diagram. We
will see similar mysteries later.
Shum’s theorem should remind the reader of Joyal and Street’s theorem
saying that Braid is the free braided monoidal category on one object. They are
the first in a long line of results that describe interesting topological structures
as free structures on one object, which often corresponds to a point. This idea
has been dubbed “the primacy of the point”.
Drinfel’d (1986)
In 1986, Vladimir Drinfel’d won the Fields medal for his work on quantum
groups [107]. This was the culmination of a long line of work on exactly solvable
problems in low-dimensional physics, which we can only briefly sketch.
Back in 1926, Heisenberg [109] considered a simplified model of a ferromagnet
like iron, consisting of spin- 12 particles—electrons in the outermost shell of the
56
iron atoms—sitting in a cubical lattice and interacting only with their nearest
neighbors. In 1931, Bethe [110] proposed an ansatz which let him exactly solve
for the eigenvalues of the Hamiltonian in Heisenberg’s model, at least in the
even simpler case of a 1-dimensional crystal. This was subsequently generalized
by Onsager [111], C. N. and C. P. Yang [112], Baxter [113] and many others.
The key turns out to be something called the ‘Yang–Baxter equation’. It’s
easiest to understand this in the context of 2-dimensional quantum field theory.
Consider a Feynman diagram where two particles come in and two go out:
//
/
/
B/
///

This corresponds to some operator
B: H ⊗ H → H ⊗ H
where H is the Hilbert space of states of the particle. It turns out that the
physics simplifies immensely, leading to exactly solvable problems, if:
??
??
?? ??
??
B ?? ??
? B
? ? ??

B ?? = B ??
?? ? ?
??
B ?? B ??
?
? ??
?

This says we can slide the lines around in a certain way without changing the
operator described by the Feynman diagram. In terms of algebra:
(B ⊗ 1)(1 ⊗ B)(B ⊗ 1) = (1 ⊗ B)(B ⊗ 1)(1 ⊗ B).
This is the Yang–Baxter equation; it makes sense in any monoidal category.
In their 1985 paper, Joyal and Street noted that given any object x in a
braided monoidal category, the braiding
Bx,x : x ⊗ x → x ⊗ x
is a solution of the Yang–Baxter equation. If we draw this equation using string
diagrams, it looks like the third Reidemeister move in knot theory:
=
%%
%%
% %
57
Joyal and Street also showed that given any solution of the Yang–Baxter equa-
tion in any monoidal category, we can build a braided monoidal category.
Mathematical physicists enjoy exactly solvable problems, so after the work
of Yang and Baxter a kind of industry developed, devoted to finding solutions
of the Yang–Baxter equation. The Russian school, led by Faddeev, Sklyanin,
Takhtajan and others, were especially successful [114]. Eventually Drinfel’d
discovered how to get solutions of the Yang–Baxter equation from any simple
Lie algebra. The Japanese mathematician Jimbo did this as well, at about the
same time [108].
What they discovered was that the universal enveloping algebra Ug of any
simple Lie algebra g can be ‘deformed’ in a manner depending on a parameter q,
giving a one-parameter family of ‘Hopf algebras’ Uq g. Since Hopf algebras are
mathematically analogous to groups and in some physics problems the parame-
ter q is related to Planck’s constant ~ by q = e~ , the Hopf algebras Uq g are called
‘quantum groups’. There is by now an extensive theory of these [115, 116, 117].
Moreover, these Hopf algebras have a special property which implies that
any representation of Uq g on a vector space V comes equipped with an operator
B: V ⊗ V → V ⊗ V
satisfying the Yang–Baxter equation. We shall say a bit more about this in our
discussion of a 1989 paper by Reshetikhin and Turaev.
This work led to a far more thorough understanding of exactly solvable prob-
lems in 2d quantum field theory [118]. It was also the first big explicit intrusion
of category theory into physics. As we shall see, Drinfel’d’s constructions can be
nicely explained in the language of braided monoidal categories. This led to the
widespread adoption of this language, which was then applied to other problems
in physics. Everything beforehand only looks category-theoretic in retrospect.
Segal (1988)
In an attempt to formalize some of the key mathematical structures underlying
string theory, Graeme Segal [119] proposed axioms describing a ‘conformal field
theory’. Roughly, these say that it is a symmetric monoidal functor
Z : 2CobC → Hilb
with some nice extra properties. Here 2CobC is the category whose morphisms
are ‘string worldsheets’, like this:

S′
58
We compose these morphisms by gluing them end to end, like this:

S′
M′

S ′′
A bit more precisely, an object 2CobC as a union of parametrized circles,
while a morphism M : S → S ′ is a 2-dimensional ‘cobordism’ equipped with
some extra structure. Here an n-dimensional ‘cobordism’ is roughly an n-
dimensional compact oriented manifold with boundary, M , whose boundary
has been written as the disjoint union of two (n − 1)-dimensional manifolds S
and S ′ , called the ‘source’ and ‘target’.
In the case of 2CobC , we need these cobordisms to be equipped with a confor-
mal structure and a parametrization of each boundary circle. The parametriza-
tion lets us give the composite of two cobordisms a conformal structure built
from the conformal structures on the two parts.
In fact we are glossing over many subtleties here; we hope the above sketch
gets the idea across. In any event, 2CobC is a symmetric monoidal category,
where we tensor objects or morphisms by setting them side by side:
S1 ⊗ S2
M⊗M ′

S1′ ⊗ S2′
Similarly, Hilb is a symmetric monoidal category with the usual tensor product
of Hilbert spaces. A basic rule of quantum physics is that the Hilbert space for
a disjoint union of two physical systems should be the tensor product of their
Hilbert spaces. This suggests that a conformal field theory, viewed as a functor
Z : 2CobC → Hilb, should preserve tensor products—at least up to a specified
isomorphism. So, we should demand that Z be a monoidal functor. A bit more
reflection along these lines leads us to demand that Z be a symmetric monoidal
functor.
There is more to the full definition of a conformal field theory than merely
a symmetric monoidal functor Z : 2CobC → Hilb. For example, we also need
a ‘positive energy’ condition reminiscent of the condition we already met for
59
representations of the Poincaré group. Indeed there is a profusion of different

ways to make the idea of conformal field theory precise, starting with Segal’s
original definition. But the different approaches are nicely related, and the
subject of conformal field theory is full of deep results, interesting classification
theorems, and applications to physics and mathematics. A good introduction
is the book by Di Francesco, Mathieu and Senechal [120].
Atiyah (1988)
Shortly after Segal proposed his definition of ‘conformal field theory’, Atiyah
[121] modified it by dropping the conformal structure and allowing cobordisms
of an arbitrary fixed dimension. He called the resulting structure a ‘topological
quantum field theory’, or ‘TQFT’ for short. One of his goals was to formalize
some work by Witten [122] on invariants of 4-dimensional manifolds coming
from a quantum field theory sometimes called ‘Donaldson theory’, which is
related to Yang–Mills theory. These invariants have led to a revolution in our
understanding of 4-dimensional topology—but ironically, Donaldson theory has
never been successfully dealt with using Atiyah’s axiomatic approach. We will
say more about this in our discussion of Crane and Frenkel’s 1994 paper. For
now, let us simply explain Atiyah’s definition of a TQFT.
In modern language, an n-dimensional TQFT is a symmetric monoidal
functor
Z : nCob → FinVect.
Here FinVect stands for the category of finite-dimensional complex vector spaces
and linear operators between them, while nCob is the category with:
• compact oriented (n − 1)-dimensional manifolds as objects;
• oriented n-dimensional cobordisms as morphisms.
Taking the disjoint union of manifolds makes nCob into a monoidal category.
The braiding in nCob can be drawn like this:
S ⊗ S′
BS,S ′

S′ ⊗ S
but because we are interested in ‘abstract’ cobordisms, not embedded in any

ambient space, this braiding will be symmetric.
Physically, idea of a TQFT is that it describes a featureless universe that
looks locally the same in every state. In such an imaginary universe, the only way
to distinguish different states is by doing ‘global’ observations, for example by
carrying a particle around a noncontractible loop in space. Thus, TQFTs appear
60
to be very simple toy models of physics, which ignore most of the interesting
features of what we see around us. It is precisely for this reason that TQFTs are
more tractable than full-fledged quantum field theories. In what follows we shall
spend quite a bit of time explaining how TQFTs are related to n-categories. If
n-categorical physics is ever to blossom, we must someday go further. There
are some signs that this may be starting [123]. But attempting to discuss this
would lead us out of our ‘prehistory’.
Mathematically, the study of topological quantum field theories quickly leads
to questions involving duals. In our explanation of the work of Freyd and Yetter
we mentioned ‘compact’ monoidal categories, where every object has a dual.
One can show that nCob is compact, with the dual x∗ of an object x being
the same manifold equipped with the opposite orientation. Similarly, FinVect
is compact with the usual notion of dual for vector spaces. The categories Vect
and Hilb are not compact, since we can always define a ‘dimension’ of an object
in a compact braided monoidal category by
1_

x O
ei ⊗
_ ei

i
O ei ⊗
_e

δii = dim(H)
but this diverges for an infinite-dimensional vector space, or Hilbert space. As

we have seen, the infinities that plague ordinary quantum field theory arise from
his fact.
As a category, FinVect is equivalent to FinHilb, the category of finite-
dimensional complex Hilbert spaces and linear operators. However, FinHilb
and also Hilb have something in common with nCob that Vect lacks: they have
‘duals for morphisms’. In nCob, given a morphism

S′
we can reverse its orientation and switch its source and target to obtain a
61
morphism going ‘backwards in time’:
S′
M†

S
Similarly, given a linear operator T : H → H ′ between Hilbert spaces, we can

define an operator T † : H ′ → H by demanding that
hT † φ, ψi = hφ, T ψi
for all vectors ψ ∈ H, φ ∈ H ′ .

Isolating the common properties of these constructions, we say a category
has duals for morphisms if for any morphism f : x → y there is a morphism
f † : y → x such that
(f † )† = f, (f g)† = g † f † , 1†x = 1x .
We then say morphism f is unitary if f † is the inverse of f . In the case of Hilb

this is just a unitary operator in the usual sense.
As we have seen, symmetries in quantum physics are described not just
by group representations on Hilbert spaces, but by unitary representations.
This is a hint of the importance of ‘duals for morphisms’ in physics. We can
always think of a group G as a category with one object and with all morphisms
invertible. This becomes a category with duals for morphisms by setting g † =
g −1 for all g ∈ G. A representation of G on a Hilbert space is the same as a
functor ρ : G → Hilb, and this representation is unitary precisely when
ρ(g † ) = ρ(g)† .
The same sort of condition shows up in many other contexts in physics. So,
quite generally, given any functor F : C → D between categories with duals for
morphisms, we say F is unitary if F (f † ) = F (f )† for every morphism in C.
It turns out that the physically most interesting TQFTs are the unitary ones,
which are unitary symmetric monoidal functors
Z : nCob → FinHilb.
While categories with duals for morphisms play a crucial role in this defini-
tion, and also 1989 paper by Doplicher and Roberts, and also the 1995 paper by
Baez and Dolan, they seem to have been a bit neglected by category theorists
until 2005, when Selinger [124] introduced them under the name of ‘dagger cate-
gories’ as part of his work on the foundations of quantum computation. Perhaps
62
one reason for this neglect is that their definition implicitly involves an equation
between objects—something normally shunned in category theory.
To see this equation between objects explicitly, note that a category with
duals for morphisms, or dagger category, may be defined as a category C
equipped with a contravariant functor † : C → C such that
†2 = 1 C
and x† = x for every object x ∈ C. Here by a contravariant functor we mean

one that reverses the order of composition: this is a just way of saying that
(f g)† = g † f † .
Contravariant functors are well-accepted in category theory, but it raises
eyebrows to impose equations between objects, like x† = x. This is not just a
matter of fashion. Such equations cause real trouble: if C is a dagger category,
and F : C → D is an equivalence of categories, we cannot use F to give D the
structure of a dagger category, precisely because of this equation. Nonetheless,
the concept of dagger category seems crucial in quantum physics. So, there is a
tension that remains to be resolved here.
The reader may note that this is not the first time an equation between
objects has obtruded in the study of duals. We have already seen one in our
discussion of Freyd and Yetter’s 1986 paper. In that case the problem involved
duals for objects, rather than morphisms. And in that case, Shum found a way
around the problem. When it comes to duals for morphisms, no comparable fix
is known. However, in our discussion of Doplicher and Roberts 1989 paper, we
will see that the two problems are closely connected.
Dijkgraaf (1989)
Shortly after Atiyah defined TQFTs, Dijkgraaf gave a purely algebraic charac-
terization of 2d TQFTs in terms of commutative Frobenius algebras [125].
Recall that a 2d TQFT is a symmetric monoidal functor Z : 2Cob → Vect.
An object of 2Cob is a compact oriented 1-dimensional manifold—a disjoint
union of copies of the circle S 1 . A morphism of 2Cob is a 2d cobordism between
such manifolds. Using ‘Morse theory’, we can chop any 2d cobordism M into
elementary building blocks that contain only a single critical point. These are
called the birth of a circle, the upside-down pair of pants, the death of
a circle and the pair of pants:
Every 2d cobordism is built from these by composition, tensoring, and the

other operations present in any symmetric monoidal category. So, we say that
2Cob is ‘generated’ as a symmetric monoidal category by the object S 1 and
63
these morphisms. Moreover, we can list a complete set of relations that these
generators satisfy:
= = = (1)
= = = (2)
= = (3)
= (4)
2Cob is completely described as a symmetric monoidal category by means of

these generators and relations.
Applying the functor Z to the circle gives a vector space F = Z(S 1 ), and
applying it to the cobordisms shown below gives certain linear maps:
i: C → F m: F ⊗ F → F ε: F → C ∆: F → F ⊗ F
This means that our 2-dimensional TQFT is completely determined by choosing
a vector space F equipped with linear maps i, m, ε, ∆ satisfying the relations
drawn as pictures above.
Surprisingly, all this stuff amounts to a well-known algebraic structure: ‘com-
mutative Frobenius algebra’. For starters, Equation 1:
F ⊗F ⊗F F ⊗F ⊗F
1F ⊗m m⊗1F

F ⊗F = F ⊗F
µ m

F F
says that the map m defines an associative multiplication on F . The second re-
lation says that the map i gives a unit for the multiplication on F . This makes F
64
into an algebra. The upside-down versions of these relations appearing in 2 say

that F is also a coalgebra. An algebra that is also a coalgebra where the multi-
plication and comultiplication are related by Equation 3 is called a Frobenius
algebra. Finally, Equation 4 is the commutative law for multiplication.
In 1996, Abrams [126] was able to construct a category of 2d TQFTs and
prove it is equivalent to the category of commutative Frobenius algebras. This
makes precise the sense in which a 2-dimensional topological quantum field
theory ‘is’ a commutative Frobenius algebra. It implies that when one has a
commutative Frobenius algebra in the category FinVect, one immediately gets a
symmetric monoidal functor Z : 2Cob → Vect, hence a 2-dimensional topological
quantum field theory. This perspective is explained in great detail in the book
by Kock [127].
In modern language, the essence of Abrams’ result is contained in the follow-
ing theorem: 2Cob is the free symmetric monoidal category on a commutative
Frobenius algebra. To make this precise, we first define a commutative Frobenius
algebra in any symmetric monoidal category, using the same diagrams as above.
Next, suppose C is any symmetric monoidal category and c ∈ C is a commu-
tative Frobenius algebra in C. Then first, there exists a symmetric monoidal
functor
Z : 2Cob → C
with
Z(S 1 ) = c
and such that Z sends the multiplication, unit, cocomultiplication and counit
for S 1 to those for c. Second, Z is unique up to a symmetric monoidal natural
isomorphism.
This result should remind the reader of Joyal and Street’s algebraic charac-
terization of the category of braids, and Shum’s characterization of the category
of framed oriented tangles. It is a bit more complicated, because the circle is a
bit more complicated than the point. The idea of an ‘extended’ TQFT, which
we shall describe later, strengthens the concept of a TQFT so as to restore “the
primacy of the point”.
Doplicher–Roberts (1989)
In 1989, Sergio Doplicher and John Roberts published a paper [128] showing
how to reconstruct a compact topological group K—for example, a compact
Lie group—from its category of finite-dimensional continuous unitary represen-
tations, Rep(K). They then used this to show one could start with a fairly
general quantum field theory and compute its gauge group, instead of putting
the group in by hand [129].
To do this, they actually needed some extra structure on Rep(K). For
our purposes, the most interesting thing they needed was its structure as a
‘symmetric monoidal category with duals’. Let us define this concept.
In our discussion of Atiyah’s 1988 paper on TQFTs, we explained what it
means for a category to be a ‘dagger category’, or have ‘duals for morphisms’.
65
When such a category is equipped with extra structure, it makes sense to de-
mand that this extra structure be compatible with this duality. For example,
we can demand that an isomorphism f : x → y be unitary, meaning
f † f = 1x , f f † = 1y .
So, we say a monoidal category C has duals for morphisms if its underlying
category has duals for morphisms, the duality preserves the tensor product:
(f ⊗ g)† = f † ⊗ g †
and moreover all the relevant isomorphisms are unitary: the associators ax,y,z ,
and the left and right unitors ℓx and rx . We say a braided or symmetric monoidal
category has duals for morphisms if all these conditions hold and in addition
the braiding Bx,y is unitary. There is an easy way to make 1Tang2 into a
braided monoidal category with duals for morphisms. Both nCob and FinHilb
are symmetric monoidal categories with duals for morphisms.
Besides duals for morphisms, we may consider duals for objects. In our
discussion of Freyd and Yetter’s 1986 paper, we said a monoidal category has
‘duals for objects’, or is ‘compact’, if for each object x there is an object x∗
together with a unit ix : 1 → x ⊗ x∗ and counit ex : x∗ ⊗ x → 1 satisfying the
zig-zag identities.
Now suppose a braided monoidal category has both duals for morphisms
and duals for objects. Then there is yet another compatibility condition we
can—and should—demand. Any object has a counit, shaped like a cup:
x∗ ⊗ x
S
ex

1
and taking the dual of this morphism we get a kind of cap:
1
T ix

x∗ ⊗ x
Combining these with the braiding we get a morphism like this:
x

x
This looks just like the morphism bx : x → x that we introduced in our discussion
of Freyd and Yetter’s 1986 paper—only now it is the result of combining duals
66
for objects and duals for morphisms! Some string diagram calculations suggest
that bx should be unitary. So, we say a braided monoidal category has duals
if it has duals for objects, duals for morphisms, and the twist isomorphism
bx : x → x, constructed as above, is unitary for every object x.
In a symmetric monoidal category with duals on can show that b2x = 1x . In
physics this leads to the boson/fermion distinction mentioned earlier, since a
boson is any particle that remains unchanged when rotated a full turn, while a
fermion is any particle whose phase gets multiplied by −1 when rotated a full
turn. Both nCob and Hilb are symmetric monoidal categories with duals, and
both are ‘bosonic’ in the sense that bx = 1x for every object. The same is true
for Rep(K) for any compact group K. This features prominently in the paper
by Doplicher and Roberts.
In recent years, interest has grown in understanding the foundations of quan-
tum physics with the help of category theory. One reason is that in theoretical
work on quantum computation, there is a fruitful overlap between the cate-
gory theory used in quantum physics and that used in computer science. In a
2004 paper on this subject, Abramsky and Coecke [130] introduced symmetric
monoidal categories with duals under the name of ‘strongly compact closed cate-
gories’. These entities were later dubbed ‘dagger compact categories’ by Selinger
[124], and this name seems to have caught on. What we are calling symmet-
ric monoidal categories with duals for morphisms, he calls ‘dagger symmetric
monoidal categories’.
Reshetikhin–Turaev (1989)
We have mentioned how Jones discovery in 1985 of a new invariant of knots
led to a burst of work on related invariants. Eventually it was found that all
these so-called ‘quantum invariants’ of knots can be derived in a systematic way
from quantum groups. A particularly clean treatment using braided monoidal
categories can be found in a paper by Nikolai Reshetikhin and Vladimir Turaev
[131]. This is a good point to summarize a bit of the theory of quantum groups
in its modern form.
The first thing to realize is that a quantum group is not a group: it is a
special sort of algebra. What quantum groups and groups have in common is
that their categories of representations have similar properties. The category of
finite-dimensional representations of a group is a symmetric monoidal category
with duals for objects. The category of finite-dimensional representations of a
quantum group is a braided monoidal category with duals for objects.
As we saw in our discussion of Freyd and Yetter’s 1986 paper, the category
1Tang2 of tangles in 3 dimensions is the free braided monoidal category with
duals on one object x. So, if Rep(A) is the category of finite-dimensional repre-
sentations of a quantum group A, any object V ∈ Rep(A) determines a braided
monoidal functor
Z : 1Tang2 → Rep(A).
67
with
Z(x) = V.
This functor gives an invariant of tangles: a linear operator for every tangle,
and in particular a number for every knot or link.
So, what sort of algebra has representations that form a braided monoidal
category with duals for objects? This turns out to be one of a family of related
questions with related answers. The more extra structure we put on an algebra,
the nicer its category of representations becomes:
algebra category
bialgebra monoidal category
quasitriangular bialgebra braided monoidal category
triangular bialgebra symmetric monoidal category
Hopf algebra monoidal category
with duals for objects
quasitriangular braided monoidal category
Hopf algebra with duals for objects
triangular symmetric monoidal category
Hopf algebra with duals for objects
Algebras and their categories of representations
For each sort of algebra A in the left-hand column, its category of representations
Rep(A) becomes a category of the sort listed in the right-hand column. In
particular, a quantum group is a kind of ‘quasitriangular Hopf algebra’.
In fact, the correspondence between algebras and their categories of repre-
sentations works both ways. Under some mild technical assumptions, we can
recover A from Rep(A) together with the ‘forgetful functor’ F : Rep(A) → Vect
sending each representation to its underlying vector space. The theorems guar-
anteeing this are called ‘Tannaka–Krein reconstruction theorems’ [132]. They
are reminiscent of the Doplicher–Roberts reconstruction theorem, which allows
us to recover a compact topological group G from its category of representations.
However, they are easier to prove, and they came earlier.
So, someone who strongly wishes to avoid learning about quasitriangular
Hopf algebras can get away with it, at least for a while, if they know enough
about braided monoidal categories with duals for objects. The latter subject is
ultimately more fundamental. Nonetheless, it is very interesting to see how the
correspondence between algebras and their categories of representations works.
So, let us sketch how any bialgebra has a monoidal category of representations,
and then give some examples coming from groups and quantum groups.
First, recall that an algebra is a vector space A equipped with an associative
multiplication
m: A ⊗ A → A
a ⊗ b 7→ ab
68
together with an element 1 ∈ A satisfying the left and right unit laws: 1a = a =
a1 for all a ∈ A. We can draw the multiplication using a string diagram:
Y E
Y E
Y E
m
O
O
We can also describe the element 1 ∈ A using the unique operator i : C → A

that sends the complex number 1 to 1 ∈ A. Then we can draw this operator
using a string diagram:
i
O
O
In this notation, the associative law looks like this:

Y E E Y Y E
Y E E Y Y E
Y E E Y Y E
E Y
m E
E Y
Y m
W E Y G
W E Y G
m = m
O O
O O
O O
O O
while the left and right unit laws look like this:
E O Y
E O Y
E O Y
E O Y
i E O Y i
E Y
W E O Y G
W E O Y G
O
m = O = m
O O O
O O O
O O O
O O O
O
A representation of an algebra is a lot like a representation of a group, except

that instead of writing ρ(g)v for the action of a group element g on a vector v,
we write ρ(a ⊗ v) for the action of an algebra element a on a vector v. More
precisely, a representation of an algebra A is a vector space V equipped with
an operator
ρ: A ⊗ V → V
satisfying these two laws:
ρ(1 ⊗ v) = v, ρ(ab ⊗ v) = ρ(a ⊗ ρ(b ⊗ v)).

69
Using string diagrams can draw ρ as follows:

Y
Y
Y
ρ
Note that wiggly lines refer to the object A, while straight ones refer to V . Then
the two laws obeyed by ρ look very much like associativity and the left unit law:
Y E Y Y
Y E Y Y
Y E Y Y
Y
m Y
Y ρ
W Y
W Y
ρ = ρ
i
W
W
ρ =
To make the representations of an algebra into the objects of a category, we

must define morphisms between them. Given two algebra representations, say
ρ : A ⊗ V → V and ρ′ : A ⊗ V ′ → V ′ , we define an intertwining operator
f : V → V ′ to be a linear operator such that
f (ρ(a ⊗ v)) = ρ′ (a ⊗ f (v)).
This closely resembles the definition of an intertwining operator between group

representations. It says that acting by a ∈ A and then applying the intertwining
operator is the same as applying the intertwining operator and then acting by
a.
With these definitions, we obtain a category Rep(A) with finite-dimensional
representations of A as objects and intertwining operators as morphisms. How-
ever, unlike group representations, there is no way in general to define the tensor
product of algebra representations! For this, we need A to be a ‘bialgebra’. To
understand what this means, first recall from our discussion of Dijkgraaf’s 1989
thesis that a coalgebra is just like an algebra, only upside-down. More pre-
70
cisely, it is a vector space equipped with a comultiplication:
O
O
∆
E Y
E Y
E Y
and counit:
O
O
ε
satisfying the coassociative law:

O O
O O
O O
O O
∆ = ∆
G Y E W
G Y E W
Y E
∆ Y
Y E
E ∆
E Y Y E E Y
E Y Y E E Y
E Y Y E E Y
and left/right counit laws:
O O O
O O O
O O O
O O O
O
O
∆ = O = ∆
G Y O E W
G Y O E W
Y O E
ε Y
Y O E
E ε
Y O E
Y O E
Y O E
A bialgebra is a vector space equipped with an algebra and coalgebra struc-

ture that are compatible in a certain way. We have already seen that a Frobenius
algebra is both an algebra and a coalgebra, with the multiplication and comulti-
plication obeying the compatibility conditions in Equation 3. A bialgebra obeys
different compatibility conditions. These can be drawn using string diagrams,
but it is more enlightening to note that they are precisely the conditions we
need to make the category of representations of an algebra A into a monoidal
category. The idea is that the comultiplication ∆ : A → A⊗ A lets us ‘duplicate’
an element A so it can act on both factors in a tensor product of representations,
71
say ρ and ρ′ :
O
O

O

∆
]
]

M ]
P
S #c #c
X
\ #c c#
ρ ρ′
This gives Rep(A) a tensor product. Similarly, we use the counit to let A act
on C as follows:
O
O C
O
O
ε
We can then write down equations saying that Rep(A) is a monoidal category
with the same associator and unitors as in Vect, and with C as its unit object.
These equations are then the definition of ‘bialgebra’.
As we have seen, the category of representations of a compact Lie group K
is also a monoidal category. In this sense, bialgebras are a generalization of such
groups. Indeed, there is a way to turn any group of this sort into a bialgebra
A, and when the group is simply connected, this bialgebra has an equivalent
category of representations:
Rep(K) ≃ Rep(A).
So, as far as its representations are concerned, there is really no difference. But
a big advantage of bialgebras is that we can often ‘deform’ them to obtain new
bialgebras that don’t come from groups.
The most important case is when K is not only simply-connected and com-
pact, but also simple, which for Lie groups means that all its normal subgroups
are finite. We have already been discussing an example: SU(2). Groups of this
sort were classified by Élie Cartan in 1894, and by the mid-1900s their theory
had grown to one of the most enormous and beautiful edifices in mathematics.
The fact that one can deform them to get interesting bialgebras called ‘quantum
groups’ opened a brand new wing in this edifice, and the experts rushed in.
A basic fact about groups of this sort is that they have ‘complex forms’. For
example, SU(2) has the complex form SL(2), consisting of 2×2 complex matrices
with determinant 1. This group contains SU(2) as a subgroup. The advantage of
SU(2) is that it is compact, which implies that its finite-dimensional continuous
72
representations can always be made unitary. The advantage of SL(2) is that it is

a complex manifold, with all the group operations being analytic functions; this
allows us to define ‘analytic’ representations of this group. For our purposes,
another advantage of SL(2) is that its Lie algebra is a complex vector space.
Luckily we do not have to choose one group over the other, since the finite-
dimensional continuous unitary representations of SU(2) correspond precisely
to the finite-dimensional analytic representations of SL(2). And as emphasized
by Hermann Weyl, every simply-connected compact simple Lie group K has a
complex Lie group G for which this relation holds!
These facts let us say a bit more about how to get a bialgebra with the same
representations as our group K. First, we take the complex form G of the group
K, and consider its Lie algebra, g. Then we let g freely generate an algebra in
which these relations hold:
xy − yx = [x, y]
for all x, y ∈ g. This algebra is called the universal enveloping algebra of
g, and denoted Ug. It is in fact a bialgebra, and we have an equivalence of
monoidal categories:
Rep(K) ≃ Rep(Ug).
What Drinfel’d discovered is that we can ‘deform’ Ug and get a quantum
group Uq g. This is a family of bialgebras depending on a complex parameter
q, with the property that Uq g ∼ = U g when q = 1. Moreover, these bialgebras
are unique, up to changes of the parameter q and other inessential variations.
In fact, quantum groups are much better than mere bialgebras: they are
‘quasitriangular Hopf algebras’. This is just an intimidating way of saying that
Rep(Uq g) is not merely a monoidal category, but in fact a braided monoidal
category with duals for objects. And this, in turn, is just an intimidating way
of saying that any representation of Uq g gives an invariant of framed oriented
tangles! Reshetikhin and Turaev’s paper explained exactly how this works.
If all this seems too abstract, take K = SU(2). From what we have already
said, these categories are equivalent:
Rep(SU(2)) ≃ Rep(Usl(2))
where sl(2) is the Lie algebra of SL(2). So, we get a braided monoidal category
with duals for objects, Rep(Uq sl(2)), which reduces to Rep(SU(2)) when we set
q = 1. This is why Uq sl(2) is often called ‘quantum SU(2)’, especially in the
physics literature.
Even better, the quantum group Uq sl(2) has a 2-dimensional representation
which reduces to the usual spin- 12 representation of SU(2) at q = 1. Using
this representation to get a tangle invariant, we obtain the Kauffman bracket—
at least up to some minor normalization issues that we shall ignore here. So,
Reshetikhin and Turaev’s paper massively generalized the Kauffman bracket
and set it into its proper context: the representation theory of quantum groups!
In our discussion of Kontsevich’s 1993 paper we will sketch how to actually
get our hands on quantum groups.
73
Witten (1989)
In the 1980s there was a lot of work on the Jones polynomial [133], leading up to
the result we just sketched: a beautiful description of this invariant in terms of
representations of quantum SU(2). Most of this early work on the Jones poly-
nomial used 2-dimensional pictures of knots and tangles—the string diagrams
we have been discussing here. This was unsatisfying in one respect: researchers
wanted an intrinsically 3-dimensional description of the Jones polynomial.
In his paper ‘Quantum field theory and the Jones polynomial’ [134], Witten
gave such a description using a gauge field theory in 3d spacetime, called Chern–
Simons theory. He also described how the category of representations of SU(2)
could be deformed into the category of representations of quantum SU(2) using
a conformal field theory called the Wess–Zumino–Witten model, which is closely
related to Chern–Simons theory. We shall say a little about this in our discussion
of Kontsevich’s 1993 paper.
Rovelli–Smolin (1990)
Around 1986, Abhay Ashtekar discovered a new formulation of general relativity,
which made it more closely resemble gauge theories such as Yang–Mills theory
[135]. In 1990, Rovelli and Smolin [136] published a paper that used this to
develop a new approach to the old and difficult problem of quantizing gravity
— that is, treating it as a quantum rather than a classical field theory. This
approach is usually called ‘loop quantum gravity’, but in its later development
it came to rely heavily on Penrose’s spin networks [137, 140]. It reduces to the
Ponzano–Regge model in the case of 3-dimensional quantum gravity; the difficult
and so far unsolved challenge is finding a correct treatment of 4-dimensional
quantum gravity in this approach, if one exists.
As we have seen, spin networks are mathematically like Feynman diagrams
with the Poincaré group replaced by SU(2). However, Feynman diagrams de-
scribe processes in ordinary quantum field theory, while spin networks describe
states in loop quantum gravity. For this reason it seemed natural to explore
the possibility that some sort of 2-dimensional diagrams going between spin
networks are needed to describe processes in loop quantum gravity. These were
introduced by Reisenberger and Rovelli in 1996 [139], and further formalized
and dubbed ‘spin foams’ in 1997 [140, 141]. As we shall see, just as Feynman
diagrams can be used to do computations in categories like the category of
Hilbert spaces, spin foams can be used to do computations in bicategories like
the bicategory of ‘2-Hilbert spaces’.
For a review of loop quantum gravity and spin foams with plenty of references
for further study, start with the article by Rovelli [142]. Then try his book [143]
and the book by Ashtekar [144].
74
Kashiwara and Lusztig (1990)

Every matrix can be written as a sum of a lower triangular matrix, a diagonal
matrix and an upper triangular matrix. Similarly, for every simple Lie algebra
g, the quantum group Uq g has a ‘triangular decomposition’
Uq g ∼
= U − g ⊗ U 0 g ⊗ U + g.
q q q
If one is interested in the braided monoidal category of finite dimensional rep-

resentations of Uq g, then it turns out that one only needs to understand the
lower triangular part Uq− g of the quantum group. Using a sophisticated geo-
metric approach Lusztig [145, 146] defined a basis for Uq− g called the ‘canon-
ical basis’, which has remarkable properties. Using algebraic methods, Kashi-
wara [147, 148, 149] defined a ‘global crystal basis’ for Uq− (g), which was later
shown by Grojnowski and Lusztig [150] to coincide with the canonical basis.
What makes the canonical basis so interesting is that given two basis ele-
ments ei and ej , their product ei ej can be expanded in terms of basis elements
X ij
ei ej = mk e k
k
where the constants mij

k are polynomials in q and q −1 , and these polynomials
have natural numbers as coefficients. If we had chosen a basis at random, we
would only expect these constants to be rational functions of q, with rational
numbers as coefficients.
The appearance of natural numbers here hints that quantum groups are just
shadows of more interesting structures where the canonical basis elements be-
come objects of a category, multiplication becomes the tensor product in this
category, and addition becomes direct sum in this category. Such a structure
could be called a categorified quantum group. Its existence was explicitly con-
jectured in a paper by Crane and Frenkel, which we will discuss below. Indeed,
this was already visible in Lusztig’s geometric approach to studying quantum
groups using so-called ‘perverse sheaves’ [151].
For a simpler example of this phenomenon, recall our discussion of Penrose’s
1971 paper. We saw that if K is a compact Lie group, the category Rep(K)
has a tensor product and direct sums. If we pick one irreducible representation
E i from each isomorphism class, then every object in Rep(K) is a direct sum
of these objects E i , which thus act as a kind of ‘basis’ for Rep(K). As a result,
we have M ij
Ei ⊗ Ej ∼= Mk ⊗ E k
k
for certain finite-dimensional vector spaces Mkij . The dimensions of these vector
spaces, say
mij ij
k = dim(Mk ),
are natural numbers. We can define an algebra with one basis vector ei for each
E i , and with a multiplication defined by
X ij
ei ej = mk e k
k
75
This algebra is called the representation ring of K, and denoted R(K). It is

associative because the tensor product in Rep(K) is associative up to isomor-
phism.
In fact, representation rings were discovered before categories of represen-
tations. Suppose someone had handed us such ring and asked us to explain it.
Then the fact that it had a basis where the constants mij k are natural numbers
would be a clue that it came from a monoidal category with direct sums!
The special properties of the canonical basis are a similar clue, but here there
is an extra complication: instead of natural numbers, we are getting polynomials
in q and q −1 with natural number coefficients. We shall give an explanation of
this later, in our discussion of Khovanov’s 1999 paper.
Kapranov–Voevodsky (1991)
Around 1991, Kapranov and Voevodsky made available a preprint in which they
initiated work on ‘2-vector spaces’ and what we now call ‘braided monoidal bi-
categories’ [153]. They also studied a higher-dimensional analogue of the Yang–
Baxter equation called the ‘Zamolodchikov tetrahedron equation’. Recall from
our discussion of Joyal and Street’s 1985 paper that any solution of the Yang–
Baxter equation gives a braided monoidal category. Kapranov and Voevodsky
argued that similarly, any solution of the Zamolodchikov tetrahedron equation
gives a braided monoidal bicategory.
The basic idea of a braided monoidal bicategory is straightforward: it is like
a braided monoidal category, but with a bicategory replacing the underlying
category. This lets us ‘weaken’ equational laws involving 1-morphisms, replac-
ing them by specified 2-isomorphisms. To obtain a useful structure we also
need to impose equational laws on these 2-isomorphisms—so-called ‘coherence
laws’. This is the tricky part, which is why Kapranov and Voevodsky’s origi-
nal definition of ‘semistrict braided monoidal 2-category’ required a number of
fixes [154, 155, 156], leading ultimately to the fully general concept of braided
monoidal bicategory introduced by McCrudden [157].
However, their key insight was striking and robust. As we have seen, any
object in a braided monoidal category gives an isomorphism
B = Bx,x : x ⊗ x → x ⊗ x
satisfying the Yang–Baxter equation
(B ⊗ 1)(1 ⊗ B)(B ⊗ 1) = (1 ⊗ B)(B ⊗ 1)(1 ⊗ B)
which in pictures corresponds to the third Reidemeister move. In a braided

monoidal bicategory, the Yang–Baxter equation holds only up to a 2-isomorphism
Y : (B ⊗ 1)(1 ⊗ B)(B ⊗ 1) ⇒ (1 ⊗ B)(B ⊗ 1)(1 ⊗ B)
which in turn satisfies the ‘Zamolodchikov tetrahedron equation’.

76
This equation is best understood using diagrams. If we think of Y as the

surface in 4-space traced out by the process of performing the third Reidemeister
move:
Y: ⇒
%%
%%
%%
then the Zamolodchikov tetrahedron equation says the surface traced out
by first performing the third Reidemeister move on a threefold crossing and then
sliding the result under a fourth strand is isotopic to that traced out by first
sliding the threefold crossing under the fourth strand and then performing the
third Reidemeister move. So, this octagon commutes:
tt JJJJJJ
tt
ttttttt JJJJJJ
v~ t BB (
HH
LL ?????

????
????

? #
{
77
;; 7
;
777 ;
?????
????
????

? #
LL {
CC
BB
JJJJJJ tt
JJJJJJ ;;
tt
ttttttt
( 77 v~ t
Just as the Yang–Baxter equation relates two different planar projections of

3 lines in R3 , the Zamolodchikov tetrahedron relates two different projections
onto R3 of 4 lines in R4 . This suggests that solutions of the Zamolodchikov
equation can give invariants of ‘2-dimensional tangles’ in 4-dimensional space
(roughly, surfaces embedded in 4-space) just as solutions of the Yang–Baxter
equation can give invariants of tangles (roughly, curves embedded in 3-space).
Indeed, this was later confirmed [158, 159, 160].
Drinfel’d’s work on quantum groups naturally gives solutions of the Yang–
Baxter equation in the category of vector spaces. This suggested to Kapranov
and Voevodsky the idea of looking for solutions of the Zamolodchikov tetrahe-
77
dron equation in some bicategory of ‘2-vector spaces’. They defined 2-vector

spaces using the following analogy:
C Vect
+ ⊕
× ⊗
0 {0}
1 C
Analogy between ordinary linear algebra and higher linear algebra
So, just as a finite-dimensional vector space may be defined as a set of the

form Cn , they defined a 2-vector space to be a category of the form Vectn .
And just as a linear operator T : Cn → Cm may be described using an m × n
matrix of complex numbers, they defined a linear functor between 2-vector
spaces to be an m × n matrix of vector spaces! Such matrices indeed act to give
functors from Vectn to Vectm . We can also add and multiply such matrices in
the usual way, but with ⊕ and ⊗ taking the place of + and ×.
Finally, there is a new layer of structure: given two linear functors S, T : Vectn
→ Vectm , Kapranov and Voevodsky defined a linear natural transformation
α : S ⇒ T to be an m × n matrix of linear operators
αij : Sij → Tij
going between the vector spaces that are the matrix entries for S and T . This
new layer of structure winds up making 2-vector spaces into the objects of a
bicategory.
Kapranov and Voevodsky called this bicategory 2Vect. They also defined
a tensor product for 2-vector spaces, which turns out to make 2Vect into a
‘monoidal bicategory’. The Zamolodchikov tetrahedron equation makes sense
in any monoidal bicategory, and any solution gives a braided monoidal bicate-
gory. Conversely, any object in a braided monoidal bicategory gives a solution
of the Zamolodchikov tetrahedron equation. These results hint that the rela-
tion between quantum groups, solutions of the Yang–Baxter equation, braided
monoidal categories and 3d topology is not a freak accident: all these concepts
may have higher-dimensional analogues! To reach these higher-dimensional ana-
logues, it seems we need to take concepts and systematically ‘boost their dimen-
sion’ by making the following replacements:
78
elements objects
equations isomorphisms
between elements between objects
sets categories
functions functors
equations natural isomorphisms
between functions between functors
Analogy between set theory and category theory
In their 1994 paper, Crane and Frenkel called this process of dimension boosting
categorification. We have already seen, for example, that the representation
category Rep(K) of a compact Lie group is a categorification of its representa-
tion ring R(K). The representation ring is a vector space; the representation
category is a 2-vector space. In fact the representation ring is an algebra, and
as we shall in our discussion of Barrett and Westbury’s 1992 paper, the repre-
sentation category is a ‘2-algebra’.
Reshetikhin–Turaev (1991)
In 1991, Reshetikhin and Turaev [152] published a paper in which they con-
structed invariants of 3-manifolds from quantum groups. These invariants were
later seen to be part of a full-fledged 3d TQFT. Their construction made rig-
orous ideas from Witten’s 1989 paper on Chern–Simons theory and the Jones
polynomial, so this TQFT is now usually called the Witten–Reshetikhin–Turaev
theory.
Their construction uses representations of a quantum group Uq g, but not the
whole category Rep(Uq g). Instead they use a special subcategory, which can be
constructed when q is a suitable root of unity. This subcategory has many nice
properties: for example, it is a braided monoidal category with duals, and also
a 2-vector space with a finite basis of simple object. These and some extra
properties are summarized by saying that this subcategory is a ‘modular tensor
category’. Such categories were later intensively studied by Turaev [162] and
many others [163]. In this work, the Witten–Reshetikhin–Turaev construction
was generalized to obtain a 3d TQFT from any modular tensor category. More-
over, it was shown that any quantum group Uq g gives rise to a modular tensor
category when q is a suitable root of unity.
However, it was later seen that in most cases there is a 4d TQFT of which
the Witten–Reshetikhin–Turaev TQFT in 3 dimensions is merely a kind of side-
effect. So, for the purposes of understanding the relation between n-categories
and TQFTs in various dimensions, it is better to postpone further treatment
of the Witten–Reshetikhin–Turaev theory until our discussion of Turaev’s 1992
paper on the 4-dimensional aspect of this theory.
79
Turaev–Viro (1992)
In 1992, the topologists Turaev and Viro [161] constructed another invariant of
3-manifolds—which we now know is part of a full-fledged 3d TQFT—from the
modular category arising from quantum SU(2). Their construction was later
generalized to all modular tensor categories, and indeed beyond. By now, any
3d TQFT arising via this construction is called a Turaev–Viro model.
The relation between the Turaev–Viro model and the Witten–Reshetikhin–
Turaev theory is subtle and interesting, but for our limited purposes a few
words will suffice. Briefly: it later became clear that a sufficiently nice braided
monoidal category lets us construct a 4-dimensional TQFT, which has a Witten–
Reshetikhin–Turaev TQFT in 3 dimensions as a kind of shadow. On the other
hand, Barrett and Westbury discovered that we only need a sufficiently nice
monoidal category to construct a 3-dimensional TQFT—and the Turaev–Viro
models are among these. This outlook makes certain patterns clearer; we shall
explain these patterns further in sections to come.
When writing their original paper, Turaev and Viro did not know about
the Ponzano–Regge model of quantum gravity. However, their construction
amounts to taking the Ponzano–Regge model and curing it of its divergent sums
by replacing SU(2) by the corresponding quantum group. Despite the many
technicalities involved, the basic idea is simple. The Ponzano–Regge model
is not a 3d TQFT, because it assigns divergent values to the operator Z(M )
for many cobordisms M . The reason is that computing this operator involves
triangulating M , labelling the edges by spins j = 0, 21 , 1, . . . , and summing
over spins. Since there are infinitely many choices of the spins, the sum may
diverge. And since the spin labelling an edge describes its length, this divergence
arises physically from the fact that we are summing over geometries that can
be arbitrarily large.
Mathematically, spins correspond to irreducible representations of SU(2).
There are, of course, infinitely many of these. The same is true for the quantum
group Uq sl(2). But in the modular tensor category, we keep only finitely many
of the irreducible representations of Uq sl(2) as objects, corresponding to the
spins j = 0, 12 , 1, . . . , k2 , where k depends on the root of unity q. This cures the
Ponzano–Regge model of its infinities. Physically, introducing the parameter
q corresponds to introducing a nonzero ‘cosmological constant’ in 3d quantum
gravity. The cosmological constant endows the vacuum with a constant energy
density and forces spacetime to curl up instead of remaining flat. This puts an
upper limit on the size of spacetime, avoiding the divergent sum over arbitrarily
large geometries.
We postpone a detailed description of the Turaev–Viro model until our dis-
cussion of Barrett and Westbury’s 1992 paper. As mentioned, this paper strips
Turaev and Viro’s construction down to its bare essentials, building a 3d TQFT
from any sufficiently nice monoidal category: the braiding is inessential. But
the work of Barrett and Westbury is a categorified version of Fukuma, Hosono
and Kawai’s work on 2d TQFTs, so we should first discuss that.
80
Fukuma–Hosono–Kawai (1992)
Fukuma, Hosono and Kawai found a way to construct two-dimensional topo-
logical quantum field theories from semisimple algebras [164]. Though they did
not put it this way, they essentially gave a recipe to turn any 2-dimensional
cobordism
S

S′
into a string diagram, and use that diagram to define an operator between vector
spaces:
Z̃(M ) : Z̃(S) → Z̃(S ′ )
This gadget Z̃ is not quite a TQFT, but with a little extra work it gives a TQFT
which we will call Z.
The recipe begins as follows. Triangulate the cobordism M :
This picture already looks a bit like a string diagram, but never mind that.
Instead, take the Poincaré dual of the triangulation, drawing a string diagram
with:
• one vertex in the center of each triangle of the original triangulation;
• one edge crossing each edge of the original triangulation.
81
We then need a way to evaluate this string diagram and get an operator.
For this, fix an associative algebra A. Then using Poincaré duality, each
triangle in the triangulation can be reinterpreted as a string diagram for multi-
plication in A:
33 3
33 ?? 3 ??
??? 3 ??
33
?
33 3
33 m 3 m
3 _ _ _ _ _ _ 3

Actually there is a slight subtlety here. The above string diagram comes with
some extra information: little arrows on the edges, which tell us which edges are
coming in and which are going out. To avoid the need for this extra information,
let us equip A with an isomorphism to its dual vector space A∗ . Then we can
take any triangulation of M and read it as a string diagram for an operator
Z̃(M ). If our triangulation gives the manifold S some number of edges, say n,
and gives S ′ some other number of edges, say n′ , then we have
Z̃(M ) : Z̃(S) → Z̃(S ′ )
where ′
Z̃(S) = A⊗n , Z̃(S ′ ) = A⊗n .
We would like this operator Z̃(M ) to be well-defined and independent of
our choice of triangulation for M . And now a miracle occurs. In terms of
triangulations, the associative law:
22 44 22
22 44 22
2 44 2
44
m 44 m
//
/ 44

m = m
82
can be redrawn as follows:
??
??
??
??
= ??
??
??
??
?
This equation is already famous in topology! It is the 2-2 move: one of two
so-called Pachner moves for changing the triangulation of a surface without
changing the surface’s topology. The other is the 1-3 move:
33 33
33 33
33 33
33 = 33
33 rLLL 33
33 rr LLL33
rr LL3
3 rrr
By repeatedly using these moves, we can go between any two triangulations of
M that restrict to the same triangulation of its boundary.
The associativity of the algebra A guarantees that the operator Z̃(M ) does
not change when we apply the 2-2 move. To ensure that Z̃(M ) is also unchanged
by 1-3 move, we require A to be ‘semisimple’. There are many equivalent ways of
defining this concept. For example, given that we are working over the complex
numbers, we can define an algebra A to be semisimple if it is isomorphic to a
finite direct sum of matrix algebras. A more conceptual definition uses the fact
that any algebra A comes equipped with a bilinear form
g(a, b) = tr(La Lb )
where La stands for left multiplication by a:
La : A → A
x 7→ ax
and tr stands for the trace. We can reinterpret g as a linear operator g : A⊗A →
C, which we can draw as a ‘cup’:
A⊗A

C
We say g is nondegenerate if we can find a is a corresponding ‘cap’ that

satisfies the zig-zag equations. Then we say the algebra A is semisimple if g is
nondegenerate. In this case, the map a 7→ g(a, ·) gives an isomorphism A ∼
= A∗ ,
83
which lets us avoid writing little arrows on our string diagram. Even better,
with the chosen cap and cup, we get the equation:
where each circle denotes the multiplication m : A ⊗ A → A. This equation

then turns out to imply the 1-3 move! Proving this is a good workout in string
diagrams and Poincaré duality.
So: starting from a semisimple algebra A, we obtain an operator Z̃(M ) from
any triangulated 2d cobordism M . Moreover, this operator is invariant under
both Pachner moves. But how does this construction give us a 2d TQFT? It is
easy to check that
Z̃(M M ′ ) = Z̃(M ) Z̃(M ′ ),
which is a step in the right direction. We have seen that Z̃(M ) is the same re-
gardless of which triangulation we pick for M , as long as we fix the triangulation
of its boundary. Unfortunately, it depends on the triangulation of the boundary:
after all, if S is the circle triangulated with n edges then Z̃(S) = A⊗n . So, we
need to deal with this problem.
Given two different triangulations of the same 1-manifold, say S and S ′ ,
we can always find a triangulated cobordism M : S → S ′ which is a cylinder,
meaning it is homeomorphic to S × [0, 1], with S and S ′ as its two ends. For
example:

S′
This cobordism gives an operator Z̃(M ) : Z̃(S) → Z̃(S ′ ), and because this op-
erator is independent of the triangulation of the interior of M , we obtain a
canonical operator from Z̃(S) to Z̃(S ′ ). In particular, when S and S ′ are equal
as triangulated manifolds, we get an operator
pS : Z̃(S) → Z̃(S).
84
This operator is not the identity, but a simple calculation shows that it is a
projection, meaning
p2S = pS .
In physics jargon, this operator acts as a projection onto the space of ‘physical
states’. And if we define Z(S) to be the range of pS , and Z(M ) to be the
restriction of Z̃(M ) to Z(S), we can check that Z is a TQFT!
How does this construction relate to the construction of 2d TQFTs from
commutative Frobenius algebras explained our discussion of Dijkgraaf’s 1989
thesis? To answer this, we need to see how the commutative Frobenius algebra
Z(S 1 ) is related to the semisimple algebra A. In fact Z(S 1 ) turns out to be the
center of A: the set of elements that commute with all other elements of A.
The proof is a nice illustration of the power of string diagrams. Consider the
simplest triangulated cylinder from S 1 to itself. We get this by taking a square,
dividing it into two triangles by drawing a diagonal line, and then curling it up
to form a cylinder:
This gives a projection
p = pS 1 : Z̃(S 1 ) → Z̃(S 1 )
whose range is Z(S 1 ). Since we have triangulated S 1 with a single edge in this
picture, we have Z̃(S 1 ) = A. So, the commutative Frobenius algebra Z(S 1 ) sits
inside A as the range of the projection p : A → A.
Let us show that the range of p is precisely the center of A. First, take the
triangulated cylinder above and draw the Poincaré dual string diagram:
11
00
00
11
11
00
00
11
Erasing everything except this string diagram, we obtain a kind of ‘formula’ for
p:
p=
85
where the little circles stand for multiplication in A. To see that p maps A onto
into its center, it suffices to check that if a lies in the center of A then pa = a.
This is a nice string diagram calculation:
a a a a a
= = = =
In the second step we use the fact that a is in the center of A; in the last step
we use semisimplicity. Similarly, to see that p maps A into its center, it suffices
to check that for any a ∈ A, the element pa commutes with every other element
of A. In string diagram notation, this says that:
a a
The proof is as follows:

a a a a a a a
= = = = = =
Barrett–Westbury (1992)
In 1992, Barrett and Westbury completed a paper that treated ideas very similar
to those of Turaev and Viro’s paper from the same year [165]. Unfortunately,
it only reached publication much later, so everyone speaks of the Turaev–Viro
model. Barrett and Westbury showed that to construct 3d TQFTs we only need
a nice monoidal category, not a braided monoidal category. More technically:
we do not need a modular tensor category; a ‘spherical category’ will suffice
[166]. Their construction can be seen as a categorified version of the Fukuma–
Hosono–Kawai construction, and we shall present it from that viewpoint.
The key to the Fukuma–Hosono–Kawai construction was getting an operator
from a triangulated 2d cobordism and checking its invariance under the 2-2 and
1-3 Pachner moves. In both these moves, the ‘before’ and ‘after’ pictures can
be seen as the front and back of a tetrahedron:
??
??
←→ ??
??
?
33 33
33 33
33 ←→ r LLL33
3 r LL3
rrr
86
All this has an analogue one dimension up. For starters, there are also
Pachner moves in 3 dimensions. The 2-3 move takes us from two tetrahedra
attached along a triangle to three sharing an edge, or vice versa:
vH) H vH) H
vvv )) HHHH vvv )) HHHH
v )) HH vv ) H
vv H vv_ _ _ )) _ _H_H
H)v) vH )) v
)
H
_) H )
)) HHH )) ←→ )) HHH ))
)) HHH )) )) HHH ))
)) HH ) )) HH )
HH ) HH )
) HH) ) HH)

On the left side we see two tetrahedra sharing a triangle, the tall isosceles
triangle in the middle. On the right we see three tetrahedra sharing an edge,
the dashed horizontal line. The 1-4 move lets us split one tetrahedron into
four, or merge four back into one:
D)) D D)))DDD
)) DDD )) DD
)) DD D D
)) DD yY3 Y)) Y DDD
D y 3 )) Y YD )
)) ←→
) y 3)
VVVV VyVVV
VVVV ))) VVVV 3))3)
VVV VVV
Given a 3d cobordism M : S → S ′ , repeatedly applying these moves lets us go

between any two triangulations of M that restrict to the same triangulation of
its boundary. Moreover, for both these moves, the ‘before’ and ‘after’ pictures
can be seen as the front and back of a 4-simplex : the 4-dimensional analogue
of a tetrahedron.
Fukuma, Hosono and Kawai constructed 2d TQFTs from certain monoids:
namely, semisimple algebras. As we have seen, the key ideas were these:
• A triangulated 2d cobordism gives an operator by letting each triangle
correspond to multiplication in a semisimple algebra.
• Since the multiplication is associative, the resulting operator is invariant
under the 2-2 Pachner move.
• Since the algebra is semisimple, the operator is also invariant under the
1-3 move.
In a very similar way, Barrett and Westbury constructed 3d TQFTs from certain
monoidal categories called ‘spherical categories’. We can think of a spherical
category as a categorified version of a semisimple algebra. The key ideas are
these:
• A triangulated compact 2d manifold gives a vector space by letting each
triangle correspond to tensor product in a spherical category.
• A triangulated 3d cobordism gives an operator by letting each tetrahedron
correspond to the associator in the spherical category.
87
• Since the associator satisfies the pentagon identity, the resulting operator
is invariant under the 2-3 Pachner move.
• Since the spherical category is ‘semisimple’, the operator is also invariant
under the 1-4 move.
The details are a bit elaborate, so let us just sketch some of the simplest,
most beautiful aspects. Recall from our discussion of Kapranov and Voevodsky’s
1991 paper that categorifying the concept of ‘vector space’ gives the concept of
‘2-vector space’. Just as there is a category Vect of vector spaces, there is a
bicategory 2Vect of 2-vector spaces, with:
• 2-vector spaces as objects,
• linear functors as morphisms,
• linear natural transformations as 2-morphisms.
In fact 2Vect is a monoidal bicategory, with a tensor product satisfying
Vectm ⊗ Vectn ≃ Vectmn .
This lets us define a 2-algebra to be a 2-vector space A that is also a monoidal

category for which the tensor product extends to a linear functor
m : A ⊗ A → A,
and for which the associator and unitors extend to linear natural transforma-
tions. We have already seen a nice example of a 2-algebra, namely Rep(K) for
a compact Lie group K. Here the tensor product is the usual tensor product of
group representations.
Now let us fix a 2-algebra A. Given a triangulated compact 2-dimensional
manifold S, we can use Poincaré duality to reinterpret each triangle as a picture
of the multiplication m : A ⊗ A → A:
33 ??
33 ??
33
33 m
33
As in the Fukuma–Hosono–Kawai model, this lets us turn the triangulated mani-

fold into a string diagram. And as before, if A is ‘semisimple’—or more precisely,
if A is a spherical category—we do not need to write little arrows on the edges
of this string diagram for it to make sense. But since everything is categorified,
this string diagram now describes linear functor. Since S has no boundary, this
string diagram starts and ends with no edges, so it describes a linear functor
from A⊗0 to itself. Just as the tensor product of zero copies of a vector space
is defined to be C, the tensor product of no copies of 2-vector space is defined
to be Vect. But a linear functor from Vect to itself is given by a 1 × 1 matrix of
88
vector spaces—that is, a vector space! This recipe gives us a vector space Z̃(S)
for any compact 2d manifold S.
Next, from a triangulated 3d cobordism M : S → S ′ , we wish to obtain a
linear operator Z̃(M ) : Z̃(S) → Z̃(S ′ ). For this, we can use Poincaré duality
to reinterpret each tetrahedron as a picture of the associator. The ‘front’ and
‘back’ of the tetrahedron correspond to the two functors that the associator goes
between:
??
??
??
?? m? m
⇒ ?? ⇒
?? m m
??

A more 3-dimensional view is helpful here. Starting from a triangulated 3d

cobordism M : S → S ′ , we can use Poincaré duality to build a piecewise-linear
cell complex, or ‘2-complex’ for short. This is a 2-dimensional generalization of
a graph; just as a graph has vertices and edges, a 2-complex has vertices, edges
and polygonal faces. The 2-complex dual to the triangulation of a 3d cobordism
has:
• one vertex in the center of each tetrahedron of the original triangulation;
• one edge crossing each triangle of the original triangulation;
• one face crossing each edge of the original triangulation.
We can interpret this 2-complex as a higher-dimensional analogue of a string
diagram, and use this to compute an operator Z̃(M ) : Z̃(S) → Z̃(S ′ ). This
outlook is stressed in ‘spin foam models’ [140, 141], of which the Turaev–Viro–
Barrett–Westbury model is the simplest and most successful.
Each tetrahedron in M gives a little piece of the 2-complex, which looks like
this:
If we look at the string diagrams on the front and back of this picture, we see
they describe the two linear functors that the associator goes between:
89
This is just a deeper look at the something we already saw in our discussion
of Ponzano and Regge’s 1968 paper. There we saw a connection between the
tetrahedron, the 6j symbols, and the associator in Rep(SU(2)). Now we are
seeing that for any spherical category, a triangulated 3d cobordism gives a 2d
cell complex built out of pieces that we can interpret as associators. So, just
as triangulated 2-manifolds give us linear functors, triangulated 3d cobordisms
gives us linear natural transformations!
More precisely, recall that every compact triangulated 2-manifold S gave
a linear functor from Vect to Vect, or 1 × 1 matrix of vector spaces, which we
reinterpreted as a vector space Z̃(S). Similarly, every triangulated 3d cobordism
M : S → S ′ gives a linear natural transformation gives between such linear
functors. This amounts to a 1 × 1 matrix of linear operators, which we can
reinterpret as a linear operator Z̃(M ) : Z̃(S) → Z̃(S ′ ).
The next step is to show that Z̃(M ) is invariant under the 2-3 and 1-4
Pachner moves. If we can do this, the rest is easy: we can follow the strategy
we have already seen in the Fukuma–Hosono–Kawai construction and obtain a
3d TQFT.
At this point another miracle comes to our rescue: the pentagon identity
gives invariance under the 2-3 move! The 2-3 move goes from two tetrahedra to
three, but each tetrahedron corresponds to an associator, so we can interpret this
move as an equation between a natural transformation built from two associators
and one built from three. And this equation is just the pentagon identity.
To see why, ponder the ‘pentagon of pentagons’ in Figure 1. This depicts
five ways to parenthesize a tensor product of objects w, x, y, z in a monoidal
category. Each corresponds to a triangulation of a pentagon. (The repeated
appearance of the number five here is just a coincidence.) We can go between
these parenthesized tensor products using the associator. In terms of triangu-
lations, each use of the associator corresponds to a 2-2 move. We can go from
the top of the picture to the lower right in two ways: one using two steps and
one using three. The two-step method builds up this picture:
H)
vv ) HH
vvv )) HHHH
vv )) HH
)vH) vH ))
)) HHH )
)) HHH ))
)) HH )
HH )
) HH)

which shows two tetrahedra attached along a triangle. The three-step method
builds up this picture: H)
vv ) HH
vvv )) HHHH
vv ) H
_)vH) vH _ _ _ ))_ _ _H
H
)) HH ))
)) HHH ))
)) H HH ))
) HH )
H
which shows three tetrahedra sharing a common edge. The pentagon identity
90
H
xvvv
vv HHHHy
HH
vv HH
v)) v
))

v vv
) vv
w )) vvv z
)) vvvv
v v
H H)
xvvv
vv HHHHy xvvv
vv ))HHHHy
HH ((w⊗x)⊗y)⊗z ) H
vv HH HH vv )) HHH
v)) v vv HH v)) v
vv vvv HH ))
)) vv vv HH )) ))
) vv vv HH ))
w )) vv z v{ v H# w ) )) z
)) vvvv (w⊗(x⊗y))⊗z (w⊗x)⊗(y⊗z) )) ))
vv ))
))
) ))
))
))
)
w⊗((x⊗y)⊗z) / w⊗(x⊗(y⊗z))
H H)
xvvv
vv HHHHy vv ))HHHHy
xvvv
HH )) HHH
vv HH vv H
Hv)) vH v)H) vH ))
H
)) HH )) HHH ))
) HH
H ) HH
H ))
w) HH z w) HH ) z
)) HH )) HH )
) HH ) HH)

Figure 1: Deriving the 2-3 Pachner move from the pentagon identity.
91
thus yields the 2-3 move:

v)HH vH) H
vvv )) HHHH vvv )) HHHH
v )) H v )) H
vv )) HH vv H
H)v) vH _)vH) vH _ _ _ ))_ _ _H
)) HHH )) = )) HHH ))
)) HHH )) )) HHH ))
)) HH ) )) H HH ))
HH ) HH )
) HH) ) H

The other axioms in the definition of spherical category then yield the 1-4 move,
and so we get a TQFT.
At this point it is worth admitting that the link between the associative
law and 2-2 move and that between the pentagon identity and 2-3 move are not
really ‘miracles’ in the sense of unexplained surprises. This is just the beginning
of a pattern that relates the n-dimensional simplex and the (n − 1)-dimensional
Stasheff associahedron. An elegant explanation of this can be found in Street’s
1987 paper ‘The algebra of oriented simplexes’ [45]—the same one in which
he proposed a simplicial approach to weak ∞-categories. Since there are also
Pachner moves in every dimension [167], the Fukuma–Hosono–Kawai model and
the Turaev–Viro–Barrett–Westbury model should be just the first of an infinite
series of constructions building (n + 1)-dimensional TQFT from ‘semisimple n-
algebras’. But this is largely open territory, apart from some important work
in 4 dimensions, which we turn to next.
Turaev (1992)
As we already mentioned, the Witten–Reshetikhin–Turaev construction of 3-
dimensional TQFTs from modular tensor categories is really just a spinoff of a
way to get 4-dimensional TQFTs from modular tensor categories. This began
becoming visible in 1991, when Turaev released a preprint [168] on building 4d
TQFTs from modular tensor categories. In 1992 he published a paper with more
details [169], and his book explains the ideas even more thoroughly [162]. His
construction amounts to a 4-dimensional analogue of the Turaev–Viro–Barrett–
Westbury construction. Namely, from a 4d cobordism M : S → S ′ , one can com-
pute a linear operator Z̃(M ) : Z̃(S) → Z̃(S ′ ) with the help of a 2-dimensional
CW complex sitting inside M . As already mentioned, we think of this complex
as a higher-dimensional analogue of a string diagram.
In 1993, following work by the physicist Ooguri [170], Crane and Yetter [171]
gave a different construction of 4d TQFTs from the modular tensor category
associated to quantum SU(2). This construction used a triangulation of M .
It was later generalized to a large class of modular tensor categories [172], and
thanks to the work of Justin Roberts [173], it is clear that Turaev’s construction
is related to the Crane–Yetter construction by Poincaré duality, following a
pattern we have seen already.
At this point the reader, seeking simplicity amid these complex historical
developments, should feel a bit puzzled. We have seen that:
92
• The Fukuma–Hosono–Kawai construction gives 2d TQFTs from sufficiently

nice monoids (semisimple algebras).
• The Turaev–Viro–Bartlett–Westbury construction gives 3d TQFTs from
sufficiently nice monoidal categories (spherical categories).
Given this, it would be natural to expect:
• Some similar construction gives 4d TQFTs from sufficiently nice monoidal
bicategories.
Indeed, this is true! Mackaay [174] proved it in 1999. But how does this square
with the following fact?
• The Turaev–Crane–Yetter construction gives 4d TQFTs from sufficiently
nice braided monoidal categories (modular tensor categories).
The answer is very nice: it turns out that braided monoidal categories are a
special case of monoidal bicategories!
We should explain this, because it is part of a fundamental pattern called
the ‘periodic table of n-categories’. As a warmup, let us see why a commutative
monoid is the same as a monoidal category with only one object. This argu-
ment goes back to work of Eckmann and Hilton [175], published in 1962. A
categorified version of their argument shows that a braided monoidal category
is the same as monoidal bicategory with only one object. This seems to have
first been noticed by Joyal and Tierney [176] around 1984.
Suppose first that C is a category with one object x. Then composition of
morphisms makes the set of morphisms from x to itself, denoted hom(x, x), into
a monoid: a set with an associative multiplication and an identity element.
Conversely, any monoid gives a category with one object in this way.
But now suppose that C is a monoidal category with one object x. Then
this object must be the unit for the tensor product. As before, hom(x, x) be-
comes a monoid using composition of morphisms. But now we can also tensor
morphisms. By Mac Lane’s coherence theorem, we may assume without loss
of generality that C is a strict monoidal category. Then the tensor product is
associative, and we have 1x ⊗ f = f = f ⊗ 1x for every f ∈ hom(x, x). So,
hom(x, x) becomes a monoid in a second way, with the same identity element.
However, the fact that tensor product is a functor implies the interchange
law:
(f f ′ ) ⊗ (gg ′ ) = (f ⊗ g)(f ′ ⊗ g ′ ).
This lets us carry out the following remarkable argument, called the Eckmann–
Hilton argument:
f ⊗ g = (1f ) ⊗ (g1)
= (1 ⊗ g)(f ⊗ 1)
= gf
= (g ⊗ 1)(1 ⊗ f )
= (g1) ⊗ (1f )
= g ⊗ f.
93
In short: composition and tensor product are equal, and they are both commu-
tative! So, hom(x, x) is a commutative monoid. Conversely, one can show that
any commutative monoid can be thought of as the morphisms in a monoidal
category with just one object.
In fact, Eckmann and Hilton came up with their argument in work on topol-
ogy, and its essence is best revealed by a picture. Let us draw the composite of
morphisms by putting one on top of the other, and draw their tensor product by
putting them side by side. We have often done this using string diagrams, but
just for a change, let us draw morphisms as squares. Then the Eckmann–Hilton
argument goes as follows:
f 1 f 1 f
f g = = = = g f
1 g g g 1
f ⊗g (1⊗g)(f ⊗1) gf (g⊗1)(1⊗f ) g⊗f
We can categorify this whole discussion. For starters, we noted in our dis-
cussion of Bénabou’s 1967 paper that if C is a bicategory with one object x,
then hom(x, x) is a monoidal category—and conversely, any monoidal category
arises in this way. Then, the Eckmann–Hilton argument can be used to show
that a monoidal bicategory with one object is a braided monoidal category.
Since categorification amounts to replacing equations with isomorphisms, each
step in the argument now gives an isomorphism:
f ⊗g ∼
= (1f ) ⊗ (g1)
∼
= (1 ⊗ g)(f ⊗ 1)
∼
= gf
∼
= (g ⊗ 1)(1 ⊗ f )
∼
= (g1) ⊗ (1f )
∼
= g ⊗ f.
Composing these, we obtain an isomorphism from f ⊗ g to g ⊗ f , which we can
think of as a braiding:
Bf,g : f ⊗ g → g ⊗ f.
We can even go further and check that this makes hom(x, x) into a braided
monoidal category.
A picture makes this plausible. We can use the third dimension to record
the process of the Eckmann–Hilton argument. If we compress f and g to small
discs for clarity, it looks like this:
f g

g
f

This clearly looks like a braiding!

94
In the above pictures we are moving f around g clockwise. There is an

alternate version of the categorified Eckmann–Hilton argument that amounts
to moving f around g counterclockwise:
f ⊗g ∼
= (f 1) ⊗ (1g)
∼
= (f ⊗ 1)(1 ⊗ g)
∼
= fg
∼
= (1 ⊗ f )(g ⊗ 1)
∼
= (1g) ⊗ (f 1)
∼
= g ⊗ f.
This gives the following picture:
f g

g f

This picture corresponds to a different isomorphism from f ⊗ g to g ⊗ f , namely

the reverse braiding
−1
Bg,f : f ⊗ g → g ⊗ f.
This is a great example of how different proofs of the same equation may give
different isomorphisms when we categorify them.
The 4d TQFTs constructed from modular tensor categories were a bit dis-
appointing, in that they gave invariants of 4-dimensional manifolds that were
already known, and unable to shed light on the deep questions of 4-dimensional
topology. The reason could be that braided monoidal categories are rather
degenerate examples of monoidal bicategories. In their 1994 paper, Crane and
Frenkel began the search for more interesting monoidal bicategories coming from
the representation theory of categorified quantum groups. As of now, it is still
unknown if these give more interesting 4d TQFTs.
Kontsevich (1993)
In his famous paper of 1993, Kontsevich [177] arrived at a deeper understanding
of quantum groups, based on ideas of Witten, but making less explicit use of
the path integral approach to quantum field theory.
In a nutshell, the idea is this. Fix a compact simply-connected simple Lie
group K and finite-dimensional representations ρ1 , . . . , ρn . Then there is a way
to attach a vector space Z(z1 , . . . zn ) to any choice of distinct points z1 , . . . , zn
in the plane, and a way to attach a linear operator
Z(f ) : Z(z1 , . . . , zn ) → Z(z1′ , . . . , zn′ )
to any n-strand braid going from the points (z1 , . . . , zn ) to the points z1′ , . . . , zn′ .
The trick is to imagine each strand of the braid as the worldline of a particle in 3d
95
spacetime. As the particles move, they interact with each other via a gauge field
satisfying the equations of Chern–Simons theory. So, we use parallel transport
to describe how their internal states change. As usual in quantum theory, this
process is described by a linear operator, and this operator is Z(f ). Since
Chern–Simons theory describe a gauge field with zero curvature, this operator
depends only on the topology of the braid. So, with some work we get a braided
monoidal category from this data. With more work we can get operators not
just for braids but also tangles—and thus, a braided monoidal category with
duals for objects. Finally, using a Tannaka–Krein reconstruction theorem, we
can show this category is the category of finite-dimensional representations of a
quasitriangular Hopf algebra: the ‘quantum group’ associated to G.
Lawrence (1993)
In 1993, Lawrence wrote an influential paper on ‘extended topological quantum
field theories’ [178], which she developed further in later work [179]. As we
have seen, many TQFTs can be constructed by first triangulating a cobordism,
attaching a piece of algebraic data to each simplex, and then using these to
construct an operator. For the procedure to give a TQFT, the resulting operator
must remain the same when we change the triangulation by a Pachner move.
Lawrence tackled the question of precisely what is going on here. Her approach
was to axiomatize a structure with operations corresponding to ways of gluing
together n-dimensional simplexes, satisfying relations that guarantee invariance
under the Pachner moves.
The use of simplexes is not ultimately the essential point here: the essential
point is that we can build any n-dimensional spacetime out of a few standard
building blocks, which can be glued together locally in a few standard ways. This
lets us describe the topology of spacetime purely combinatorially, by saying how
the building blocks have been assembled. This reduces the problem of building
TQFTs to an essentially algebraic problem, though one of a novel sort.
(Here we are glossing over the distinction between topological, piecewise-
linear and smooth manifolds. Despite the term ‘TQFT’, our description is
really suited to the case of piecewise-linear manifolds, which can be chopped
into simplexes or other polyhedra. Luckily there is no serious difference between
piecewise-linear and smooth manifolds in dimensions below 7, and both these
agree with topological manifolds below dimension 4.)
Not every TQFT need arise from this sort of recipe: we loosely use the term
‘extended TQFT’ for those that do. The idea is that while an ordinary TQFT
only gives operators for n-dimensional manifolds with boundary (or more pre-
cisely, cobordisms), an ‘extended’ one assigns some sort of data to n-dimensional
manifolds with corners—for example, simplexes and other polyhedra. This is
a physically natural requirement, so it is believed that the most interesting
TQFT’s are extended ones.
In ordinary algebra we depict multiplication by setting symbols side-by-
side on a line: multiplying a and b gives ab. In category theory we visualize
morphisms as arrows, which we glue together end to end in a one-dimensional
96
way. In studying TQFTs we need ‘higher-dimensional algebra’ to describe how

to glue pieces of spacetime together.
The idea of higher-dimensional algebra had been around for several decades,
but by this time it began to really catch on. For example, in 1992 Brown wrote
a popular exposition of higher-dimensional algebra, aptly titled ‘Out of line’
[180]. It became clear that n-categories should provide a very general approach
to higher-dimensional algebra, since they have ways of composing n-morphisms
that mimic ways of gluing together n-dimensional simplexes, globes, or other
shapes. Unfortunately, the theory of n-categories was still in its early stages of
development, limiting its potential as a tool for studying extended TQFT’s.
For this reason, a partial implementation of the idea of extended TQFT
became of interest—see for example Crane’s 1995 paper [181]. Instead of work-
ing with the symmetric monoidal category nCob, he began to grapple with the
symmetric monoidal bicategory nCob2 , where, roughly speaking:
• objects are compact oriented (n − 2)-dimensional manifolds;
• morphisms are (n − 1)-dimensional cobordisms;
• 2-morphisms are n-dimensional ‘cobordisms between cobordisms’.
His idea was that a ‘once extended TQFT’ should be a symmetric monoidal
functor
Z : nCob2 → 2Vect.
In this approach, ‘cobordisms between cobordisms’ are described using man-
ifolds with corners. The details are still a bit tricky: it seems the first precise
general construction of nCob2 as a bicategory was given by Morton [182] in 2006,
and in 2009 Schommer-Pries proved that 2Cob2 was a symmetric monoidal bi-
category [183]. Lurie’s [184] more powerful approach goes in a somewhat differ-
ent direction, as we explain in our discussion of Baez and Dolan’s 1995 paper.
Since 2d TQFTs are completely classified by the result in Dijkgraaf’s 1989
thesis, the concept of ‘once extended TQFT’ may seem like overkill in dimension
2. But this would be a short-sighted attitude. Around 2001, motivated in part
by work on D-branes in string theory, Moore and Segal [185, 186] introduced
once extended 2d TQFTs under the name of ‘open-closed topological string
theories’. However, they did not describe these using the bicategory 2Cob2 .
Instead, they considered a symmetric monoidal category 2Cobext whose objects
include not just compact 1-dimensional manifolds like the circle (‘closed strings’)
but also 1-dimensional manifolds with boundary like the interval (‘open strings’).
Here are morphisms that generate 2Cobext as a symmetric monoidal category:
Using these, Moore and Segal showed that a once extended 2d TQFT gives
a Frobenius algebra for the interval and a commutative Frobenius algebra for
97
the circle. The operations in these Frobenius algebras account for all but the
last two morphisms shown above. The last two give a projection from the first
Frobenius algebra to the second, and an inclusion of the second into the center
of the first.
Later, Lauda and Pfeiffer [190] gave a detailed proof that 2Cobext is the free
symmetric monoidal category on a Frobenius algebra equipped with a projection
into its center satisfying certain relations. Using this, they showed [191] that
the Fukuma–Hosono–Kawai construction can be extended to obtain symmetric
monoidal functors Z : 2Cobext → Vect. Fjelstad, Fuchs, Runkel and Schweigert
have gone in a different direction, describing full-fledged open-closed conformal
field theories using Frobenius algebras [187, 188, 189].
Once extended TQFTs should be even more interesting in dimension 3. At
least in a rough way, we can see how the Turaev–Viro–Barrett–Westbury con-
struction should generalize to give examples of such theories. Recall that this
construction starts with a 2-algebra A ∈ 2Vect satisfying some extra conditions.
Then:
• A triangulated compact 1d manifold S gives a 2-vector space Z̃(S) built
by tensoring one copy of A for each edge in S.
• A triangulated 2d cobordism M : S → S ′ gives a linear functor Z̃(M ) :
Z̃(S) → Z̃(S ′ ) built out of one multiplication functor m : A ⊗ A → A for
each triangle in M .
• A triangulated 3d cobordism between cobordisms α : M ⇒ M ′ gives a
linear natural transformation Z̃(α) : Z̃(M ) ⇒ Z̃(M ′ ) built out of one as-
sociator for each tetrahedron in α.
From Z̃ we should then be able to construct a once extended 3d TQFT
Z : 3Cob2 → 2Vect.
However, to the best of our knowledge, this construction has not been carried
out. The work of Kerler and Lyubashenko constructs the Witten–Reshetikhin–
Turaev theory as a kind of extended 3d TQFT using a somewhat different
formalism: ‘double categories’ instead of bicategories [192].
Crane–Frenkel (1994)
In 1994, Louis Crane and Igor Frenkel wrote a paper entitled ‘Four dimensional
topological quantum field theory, Hopf categories, and the canonical bases’ [193].
In this paper they discussed algebraic structures that provide TQFTs in various
98
low dimensions:
n=4 trialgebras
BB Hopf categories
BB monoidal bicategories
BB ||| BB ||
BB || BB |||
BB || BB |
B || B |||
n=3 Hopf algebras
BB monoidal categories
BB ||
BB |||
BB |
B |||
n=2 algebras
This chart is a bit schematic, so let us expand on it a bit. In our discussion of

Fukuma, Hosono and Kawai’s 1992 paper, we have seen how they constructed
2d TQFTs from certain algebras, namely semisimple algebras. In our discussion
of Barrett and Westbury’s paper from the same year, we have seen how they
constructed 3d TQFTs from certain monoidal categories, namely spherical cat-
egories. But any Hopf algebra has a monoidal category of representations, and
we can use Tannaka–Krein reconstruction to recover a Hopf algebra from its
category of representations. This suggests that we might be able to construct
3d TQFTs directly from certain Hopf algebras. Indeed, this is the case, as was
shown by Kuperberg [194] and Chung–Fukuma–Shapere [195]. Indeed, there is
a beautiful direct relation between 3-dimensional topology and the Hopf algebra
axioms.
Crane and Frenkel speculated on how this pattern continues in higher di-
mensions. To anyone who understands the ‘dimension-boosting’ nature of cate-
gorification, it is natural to guess that one can construct 4d TQFTs from certain
monoidal bicategories. Indeed, as we have mentioned, this was later shown by
Mackaay [174], who was greatly influenced by the Crane–Frenkel paper. But
this in turn suggests that we could obtain monoidal bicategories by considering
‘2-representations’ of categorified Hopf algebras, or ‘Hopf categories’—and that
perhaps we could construct 4d TQFTs directly from certain Hopf categories.
This may be true. In 1997, Neuchl [196] gave a definition of Hopf categories
and showed that a Hopf category has a monoidal bicategory of 2-representations
on 2-vector spaces. In 1998, Carter, Kauffman and Saito [197] found beautiful
relations between 4-dimensional topology and the Hopf category axioms.
Crane and Frenkel also suggested that there should be some kind of algebra
whose category of representations was a Hopf category. They called this a ‘tri-
algebra’. They sketched the definition; in 2004 Pfeiffer [198] gave a more precise
treatment and showed that any trialgebra has a Hopf category of representa-
tions.
However, defining these structures is just the first step toward constructing
interesting 4d TQFTs. As Crane and Frenkel put it:
To proceed any further we need a miracle, namely, the existence of
an interesting family of Hopf categories.
99
Many of the combinatorial constructions of 3-dimensional TQFTs input a Hopf

algebra, or the representation category of a Hopf algebra, and produce a TQFT.
However, the most interesting class of 3-dimensional TQFTs come from Hopf
algebras that are deformed universal enveloping algebras Uq g. The question is
where can one find an interesting class of Hopf categories that will give invariants
that are useful in 4d topology.
Topology in 4 dimensions is very different from lower dimensions: it is the
first dimension where homeomorphic manifolds can fail to diffeomorphic. In
fact, there exist exotic R4 ’s: manifolds homeomorphic to R4 but not diffeo-
morphic to it. This is the only dimension in which exotic Rn ’s exist! The
discovery of exotic R4 ’s relied on invariants coming from quantum field theory
that can distinguish between homeomorphic 4-dimensional manifolds that are
not diffeomorphic. Indeed this subject, known as ‘Donaldson theory’ [199], is
what motivated Witten to invent the term ‘topological quantum field theory’
in the first place [122]. Later, Seiberg and Witten revolutionized this subject
with a streamlined approach [200, 201], and Donaldson theory was rebaptized
‘Seiberg–Witten theory’. There are by now some good introductory texts on
these matters [202, 203, 204]. The book by Scorpan [205] is especially inviting.
But this mystery remains: how—if at all! —can the 4-manifold invariants
coming from quantum field theory be computed using Hopf categories, trialge-
bras or related structures? While such structures would give TQFTs suitable
for piecewise-linear manifolds, there is no essential difference between piecewise-
linear and smooth manifolds in dimension 4. Unfortunately, interesting exam-
ples of Hopf categories seem hard to construct.
Luckily Crane and Frenkel did more than sketch the definition of a Hopf
category. They also conjectured where examples might arise:
The next important input is the existence of the canonical bases,
for a special family of Hopf algebras, namely, the quantum groups.
These bases are actually an indication of the existence of a family
of Hopf categories, with structures closely related to the quantum
groups.
Crane and Frenkel suggested that the existence of the Lusztig–Kashiwara canon-
ical bases for upper triangular part of the enveloping algebra, and the Lusztig
canonical bases for the entire quantum groups, give strong evidence that quan-
tum groups are the shadows of a much richer structure that we might call a
‘categorified quantum group’.
Lusztig’s geometric approach produces monoidal categories associated to
quantum groups: categories of perverse sheaves. Crane and Frenkel hoped
that these categories could be given a combinatorial or algebraic formulation
revealing a Hopf category structure. Recently there has been some progress
towards fulfilling Crane and Frenkel’s hopes. In particular, these categories of
perverse sheaves have been reformulated into an algebraic language related to
the categorification of Uq+ g [206, 207]. The entire quantum group Uq sln has been
categorified by Khovanov and Lauda [208, 209], and they also gave a conjectural
categorification of the entire quantum group Uq g for every simple Lie algebra g.
100
Categorified representation theory, or ‘2-representation theory’, has taken off,

thanks largely to the foundational work of Chuang and Rouquier [210, 211].
There is much more that needs to be understood. In particular, categorifi-
cation of quantum groups at roots of unity has received only a little attention
[212], and the Hopf category structure has not been fully developed. Further-
more, these approaches have not yet obtained braided monoidal bicategories of
2-representations of categorified quantum groups. Nor have they constructed
4d TQFTs.
Freed (1994)
In 1994, Freed published an important paper [213] which exhibited how higher-
dimensional algebraic structures arise naturally from the Lagrangian formula-
tion of topological quantum field theory. Among many other things, this paper
clarified the connection between quasitriangular Hopf algebras and 3d TQFTs.
It also introduced an informal concept of ‘2-Hilbert space’ categorifying the
concept of Hilbert space. This was later made precise, at least in the finite-
dimensional case [214, 215], so it is now tempting to believe that much of the
formalism of quantum theory can be categorified. The subtleties of analysis in-
volved in understanding infinite-dimensional 2-Hilbert spaces remain challeng-
ing, with close connections to the theory of von Neumann algebras [216].
Kontsevich (1994)
In a lecture at the 1994 International Congress of Mathematicians in Zürich,
Kontsevich [217] proposed the ‘homological mirror symmetry conjecture’, which
led to a burst of work relating string theory to higher categorical structures. A
detailed discussion of this work would drastically increase the size of this paper.
So, we content ourselves with a few elementary remarks.
We have already mentioned the concept of an ‘A∞ space’: a topological space
equipped with a multiplication that is associative up to a homotopy that satisfies
the pentagon equation up to a homotopy... and so on, forever, in a manner
governed by the Stasheff polytopes [23]. This concept can be generalized to any
context that allows for a notion of homotopy between maps. In particular, it
generalizes to the world of ‘homological algebra’, which is a simplified version of
the world of homotopy theory. In homological algebra, the structure that takes
the place of a topological space is a chain complex: a sequence of abelian
groups and homomorphisms
d1 d2 d3
V0 o V1 o V2 o ···
with di di+1 = 0. In applications to physics, we focus on the case where the

Vi are vector spaces and the di are linear operators. Regardless of this, we
can define maps between chain complexes, called ‘chain maps’, and homotopies
between chain maps, called ‘chain homotopies’.
101
topological spaces chain complexes

continuous maps chain maps
homotopies chain homotopies
Analogy between homotopy theory and homological algebra
For a very readable introduction to these matters, see the book by Rotman
[220]; for a more strenuous one that goes further, try the book with the same
title by Weibel [221].
The analogy between homotopy theory and homological algebra ultimately
arises from the fact that while homotopy types can be seen as ∞-groupoids,
chain complexes can be seen as ∞-groupoids that are ‘strict’ and also ‘abelian’.
The process of turning a topological space into a chain complex, so important
in algebraic topology, thus amounts to taking a ∞-groupoid and simplifying it
by making it strict and abelian.
Since this fact is less widely appreciated than it should be, let us quickly
sketch the basic idea. Given a chain complex V , each element of V0 corresponds
to an object in the corresponding ∞-groupoid. Given objects x, y ∈ V0 , a
morphism f : x → y corresponds to an element f ∈ V1 with
d1 f + x = y.
Given morphisms f, g : x → y, a 2-morphism α : f ⇒ g corresponds to an

element α ∈ V2 with
d2 α + f = g,
and so on. The equation di di+1 = 0 then says that an (i + 1)-morphism can
only go between two i-morphisms that share the same source and target—just
as we expect in the globular approach to ∞-categories.
The analogue of an A∞ space in the world of chain complexes is called an
‘A∞ algebra’ [29, 218, 219]. More generally, one can define a structure called
an ‘A∞ category’, which has a set of objects, a chain complex hom(x, y) for
any pair of objects, and a composition map that is associative up to a chain
homotopy that satisfies the pentagon identity up to a chain homotopy... and so
on. Just as a monoid is the same as a category with one object, an A∞ algebra
is the same as an A∞ category with one object.
Kontsevich used the language of A∞ categories to formulate a conjecture
about ‘mirror symmetry’, a phenomenon already studied by string theorists.
Mirror symmetry refers to the observation that various pairs of superficially dif-
ferent string theories seem in fact to be isomorphic. In Kontsevich’s conjecture,
each of these theories is a ‘open-closed topological string theory’. We already
introduced this concept near the end of our discussion of Lawrence’s 1993 pa-
per. Recall that such a theory is designed to describe processes involving open
strings (intervals) and closed strings (circles). The basic building blocks of such
102
processes are these:
In the simple approach we discussed, the space of states of the open string is
a Frobenius algebra. The space of states of the closed string is a commutative
Frobenius algebra, typically the center of the Frobenius algebra for the open
string. In the richer approach developed by Kontsevich and subsequent authors,
notably Costello [222], states of the open string are instead described by an
A∞ category with some extra structure mimicking that of a Frobenius algebra.
The space of states of the closed string is obtained from this using a subtle
generalization of the concept of ‘center’.
To get some sense of this, let us ignore the ‘Frobenius’ aspects and simply
regard the space of states of an open string as an algebra. Multiplication in this
algebra describes the process of two open strings colliding and merging together:
The work in question generalizes this simple idea in two ways. First, it treats
an algebra as a special case of an A∞ algebra, namely one for which only the
0th vector space in its underlying chain complex is nontrivial. Second, it treats
an A∞ algebra as a special case of an A∞ category, namely an A∞ category
with just one object.
How should we understand a general A∞ category as describing the states
of an open-closed topological string? First, the different objects of the A∞ cate-
gory correspond to different boundary conditions for an open string. In physics
these boundary conditions are called ‘D-branes’, because they are thought of
as membranes in spacetime on which the open strings begin or end. The ‘D’
stands for Dirichlet, who studied boundary conditions back in the mid-1800’s.
A good introduction to D-branes from a physics perspective can be found in
Polchinski’s books [91].
For any pair of D-branes x and y, the A∞ category gives a chain complex
hom(x, y). What is the physical meaning of this? It is the space of states
for an open string that starts on the D-brane x and ends on the D-brane y.
Composition describes a process where open strings in the states g ∈ hom(x, y)
and f ∈ hom(y, z) collide and stick together to form an open string in the state
f g ∈ hom(x, z).
However, note that the space of states hom(x, y) is not a mere vector space.
It is a chain complex—so it is secretly a strict ∞-groupoid! This lets us talk
about states that are not equal, but still isomorphic. In particular, composition
in an A∞ category is associative only up to isomorphism: the states (f g)h and
103
f (gh) are not usually equal, merely isomorphic via an associator:
af,g,h : (f g)h → f (gh).
In the language of chain complexes, we write this as follows:
daf,g,h + (f g)h = f (gh).
This is just the first of an infinite list of equations that are part of the usual
definition of an A∞ category. The next one says that the associator satisfies the
pentagon identity up to d of something, and so on.
Kontsevich formulated his homological mirror symmetry conjecture as the
statement that two A∞ categories are equivalent. The conjecture remains un-
proved in general, but many special cases are known. Perhaps more importantly,
the conjecture has become part of an elaborate web of ideas relating gauge the-
ory to the ‘Langlands program’—which itself is a vast generalization of the circle
of ideas that gave birth to Wiles’ proof of Fermat’s Last Theorem. For a good
introduction to all this, see the survey by Edward Frenkel [223].
Gordon–Power–Street (1995)
In 1995, Gordon, Power and Street introduced the definition and basic theory of
‘tricategories’—or in other words, weak 3-categories [224]. Among other things,
they defined a ‘monoidal bicategory’ to be a tricategory with one object. They
then showed that a monoidal bicategory with one object is the same as a braided
monoidal category. This is a precise working-out of the categorified Eckmann–
Hilton argument sketched in our discussion of Turaev’s 1992 paper.
So, a tricategory with just one object and one morphism is the same as
a braided monoidal category. There is also, however, a notion of ‘strict 3-
category’: a tricategory where all the relevant laws hold as equations, not merely
up to equivalence. Not surprisingly, a strict 3-category with one object and one
morphism is a braided monoidal category where all the braiding, associator
and unitors are identity morphisms. This rules out the possibility of nontrivial
braiding, which occurs in categories of braids or tangles. As a consequence, not
every tricategory is equivalent to a strict 3-category.
All this stands in violent contrast to the story one dimension down, where
a generalization of Mac Lane’s coherence theorem can be used to show every
bicategory is equivalent to a strict 2-category. So, while it was already known in
some quarters [176], Gordon, Power and Street’s book made the need for weak
n-categories clear to all: in a world where all tricategories were equivalent to
strict 3-categories, there would be no knots!
Gordon, Power and Street did, however, show that every tricategory is equiv-
alent to a ‘semistrict’ 3-category, in which some but not all the laws hold as
equations. They called these semistrict 3-categories ‘Gray-categories’, since
their definition relies on John Gray’s prescient early work [225]. Constructing a
workable theory of semistrict n-categories for all n remains a major challenge.
104
Baez–Dolan (1995)
In [226], Baez and Dolan outlined a program for understanding extended TQFTs
in terms of n-categories. A key part of this is the ‘periodic table of n-categories’.
Since this only involves weak n-categories, let us drop the qualifier ‘weak’ for the
rest of this section, and take it as given. Also, just for the sake of definiteness,
let us take a globular approach to n-categories:

• • /• • D• • _*4
E• Globes

% y
So, in this section ‘2-category’ will mean ‘bicategory’ and ‘3-category’ will mean
‘tricategory’. (Recently this sort of terminology as been catching on, since the
use of Greek prefixes to name weak n-categories becomes inconvenient as the
value of n becomes large.)
We have already seen the beginning of a pattern involving these concepts:
• A category with one object is a monoid.
• A 2-category with one object is a monoidal category.
• A 3-category with one object is a monoidal 2-category.
The idea is that we can take an n-category with one object and think of it as
an (n − 1)-category by ignoring the object, renaming the morphisms ‘objects’,
renaming the 2-morphisms ‘morphisms’, and so on. Our ability to compose
morphisms in the original n-category gets reinterpreted as an ability to ‘tensor’
objects in the resulting (n − 1)-category, so we get a ‘monoidal’ (n − 1)-category.
However, we can go further: we can consider a monoidal n-category with
one object. We have already looked at two cases of this, and we can imagine
more:
• A monoidal category with one object is a commutative monoid.
• A monoidal 2-category with one object is a braided monoidal category.
• A monoidal 3-category with one object is a braided monoidal 2-category.
Here the Eckmann–Hilton argument comes into play, as explained our discussion
of Turaev’s 1992 paper. The idea is that given a monoidal n-category C with
one object, this object must be the unit for the tensor product, 1 ∈ C. We can
focus attention on hom(1, 1), which an (n − 1)-category. Given f, g ∈ hom(1, 1),
there are two ways to combine them: we can compose them, or tensor them.
105
As we have seen, we can visualize these operations as putting together little

squares in two ways: vertically, or horizontally.
g
f g
f
fg f ⊗g
These operations are related by an ‘interchange’ morphism
(f f ′ ) ⊗ (gg ′ ) → (f ⊗ g)(f ′ ⊗ g ′ ),
which is an equivalence (that is, invertible in a suitably weakened sense). This

allow us to carry out the Eckmann–Hilton argument and get a braiding on
hom(1, 1):
Bf,g : f ⊗ g → g ⊗ f.
Next, consider braided monoidal n-categories with one object. Here the
pattern seems to go like this:
• A braided monoidal category with one object is a commutative monoid.
• A braided monoidal 2-category with one object is a symmetric monoidal
category.
• A braided monoidal 3-category with one object is a sylleptic monoidal
2-category.
• A braided monoidal 4-category with one object is a sylleptic monoidal
3-category.
The idea is that given a braided monoidal n-category with one object, we can
think of it as an (n − 1)-category with three ways to combine objects, all related
by interchange equivalences. We should visualize these as the three obvious
ways of putting together little cubes: side by side, one in front of the other, and
one on top of the other.
In the first case listed above, the third operation doesn’t give anything new.
Just like a monoidal category with one object, a braided monoidal category with
one object is merely a commutative monoid. In the next case we get something
new: a braided monoidal 2-category with one object is a symmetric monoidal
category. The reason is that the third monoidal structure allows us to interpolate
between the Eckmann–Hilton argument that gives the braiding by moving f and
g around clockwise, and the argument that gives the reverse braiding by moving
them around them counterclockwise. We obtain the equation
f g g f
=

106
which characterizes a symmetric monoidal category.

In the case after this, instead of an equation, we obtain an 2-isomorphism
that describes the process of interpolating between the braiding and the reverse
braiding:
x y y x
sf,g : ⇒

The reader should endeavor to imagine these pictures as drawn in 4-dimensional

space, so that there is room to push the top strand in the left-hand picture ‘up
into the fourth dimension’, slide it behind the other strand, and then push it back
down, getting the right-hand picture. Day and Street [156] later dubbed this 2-
isomorphism sf,g the ‘syllepsis’ and formalized the theory of sylleptic monoidal
2-categories. The definition of a fully weak sylleptic monoidal 2-category was
introduced still later by Street’s student McCrudden [157].
To better understand the patterns at work here, it is useful to define a ‘k-
tuply monoidal n-category’ to be an (n + k)-category with just one j-morphism
for j < k. A chart of these appears below. This is called the ‘periodic table’,
since like Mendeleyev’s original periodic table it guides us in extrapolating the
behavior of n-categories from simple cases to more complicated ones. It is not
really ‘periodic’ in any obvious way.
n=0 n=1 n=2

k=0 sets categories 2-categories
k=1 monoids monoidal monoidal
categories 2-categories
k=2 commutative braided braided
monoids monoidal monoidal
k=3 ‘’ symmetric sylleptic
monoidal monoidal
k=4 ‘’ ‘’ symmetric
monoidal
2-categories
k=5 ‘’ ‘’ ‘’
k=6 ‘’ ‘’ ‘’
The Periodic Table:

hypothesized table of k-tuply monoidal n-categories
The periodic table should be taken with a grain of salt. For example, a
107
claim like ‘2-categories with one object and one morphism are the same as
commutative monoids’ needs to be made more precise. Its truth may depend
on whether we consider commutative monoids as forming a category, or a 2-
category, or a 3-category! This has been investigated by Cheng and Gurski
[227]. There have also been attempts to craft an approach that avoids such
subtleties [228].
But please ignore such matters for now: just stare at the table. The most no-
table feature is that the nth column of the periodic table seems to stop changing
when k reaches n + 2. Baez and Dolan called this the ‘stabilization hypothesis’.
The idea is that adding extra monoidal structures ceases to matter at this point.
Simpson later proved a version of this hypothesis in his approach to n-categories
[64]. So, let us assume the stabilization hypothesis is true, and call a k-tuply
monoidal n-category with k ≥ n + 2 a ‘stable n-category’.
In fact, stabilization is just the simplest of the many intricate patterns lurk-
ing in the periodic table. For example, the reader will note that the syllepsis
−1
sf,g : Bf,g ⇒ Bg,f
is somewhat reminiscent of the braiding itself:
Bf,g : f ⊗ g → g ⊗ f.
Indeed, this is the beginning of a pattern that continues as we zig-zag down

the table starting with monoids. To go from monoids to commutative monoids
we add the equation f g = gf . To go from commutative monoids to braided
monoidal categories we then replace this equation by an isomorphism, the braid-
ing Bf,g : f ⊗ g → g ⊗ f . But the braiding engenders another isomorphism with
−1
the same source and target: the reverse braiding Bg,f . To go from braided
monoidal categories to symmetric monoidal categories we add the equation
−1
Bf,g = Bg,f . To go from symmetric monoidal categories to sylleptic monoidal
2-categories we then replace this equation by a 2-isomorphism, the syllepsis
−1
sf,g : Bf,g ⇒ Bg,f . But this engenders another 2-isomorphism with same source
and target: the ‘reverse syllepsis’. Geometrically speaking, this is because we
can also deform the left braid to the right one here:
x y y x
⇒

by pushing the top strand down into the fourth dimension and then behind the
other strand. To go from sylleptic monoidal 2-categories to symmetric ones, we
add an equation saying the syllepsis equals the reverse syllepsis. And so on,
forever! As we zig-zag down the diagonal, we meet ways of switching between
ways of switching between... ways of switching things.
This is still just the tip of the iceberg: the patterns that arise further from the
bottom edge of the periodic table are vastly more intricate. To give just a taste
108
of their subtlety, consider the remarkable story told in Kontsevich’s 1999 paper
Operads and Motives and Deformation Quantization [229]. Kontsevich had an
amazing realization: quantization of ordinary classical mechanics problems can
be carried out in a systematic way using ideas from string theory. A thorough
and rigorous approach to this issue required proving a conjecture by Deligne.
However, early attempts to prove Deligne’s conjecture had a flaw, first noted
by Tamarkin, whose simplest manifestation—translated into the language of
n-categories—involves an operation that first appears for braided monoidal 6-
categories!
For this sort of reason, one would really like to see precisely what features
are being added as we march down any column of the periodic table. Batanin’s
approach to n-categories offers a beautiful answer based on the combinatorics
of trees [78]. Unfortunately, explaining this here would take us too far afield.
The slides of a lecture Batanin delivered in 2006 give a taste of the richness of
his work [68].
Baez and Dolan also emphasized the importance of n-categories with duals
at all levels: duals for objects, duals for morphisms, ... and so on, up to n-
morphisms. Unfortunately, they were only able to precisely define this notion
in some simple cases. For example, in our discussion of Doplicher and Roberts’
1989 paper we defined monoidal, braided monoidal, and symmetric monoidal
categories with duals—meaning duals for both objects and morphisms. We
noted that tangles in 3d space can be seen as morphisms in the free braided
monoidal category on one object. This is part of a larger pattern:
• The category of framed 1d tangles in 2d space, 1Tang1 , is the free monoidal
category with duals on one object.
• The category of framed 1d tangles in 3d space, 1Tang2 , is the free braided

monoidal category with duals on one object.
• The category of framed 1d tangles in 4d space, 1Tang3 , is the free sym-
metric monoidal category with duals on one object.
A technical point: here we are using ‘framed’ to mean ‘equipped with a triv-
ialization of the normal bundle’. This is how the word is used in homotopy
theory, as opposed to knot theory. In fact a framing in this sense determines an
orientation, so a ‘framed 1d tangle in 3d space’ is what ordinary knot theorists
would call a ‘framed oriented tangle’.
Based on these and other examples, Baez and Dolan formulated the ‘tangle
hypothesis’. This concerns a conjectured n-category nTangk where:
• objects are collections of framed points in [0, 1]k ,
• morphisms are framed 1d tangles in [0, 1]k+1 ,
• 2-morphisms are framed 2d tangles in [0, 1]k+2 ,
• and so on up to dimension (n − 1), and finally:
109
• n-morphisms are isotopy classes of framed n-dimensional tangles in [0, 1]n+k .

For short, we call the n-morphisms ‘n-tangles in (n + k) dimensions’. Figure
2 may help the reader see how simple these actually are: it shows a typical n-
tangle in (n + k) dimensions for various values of n and k. This figure is a close
relative of the periodic table. The number n is the dimension of the tangle,
while k is its codimension: that is, the number of extra dimensions of space.
The tangle hypothesis says that nTangk is the free k-tuply monoidal n-
category with duals on one object. As usual, the one object, x, is simply a
point. More precisely, x can be any point in [0, 1]k equipped with a framing
that makes it positively oriented.
Combining the stabilization hypothesis and the tangle hypothesis, we obtain
an interesting conclusion: the n-category nTangk stabilizes when k reaches n+2.
This idea is backed up by a well-known fact in topology: any two embeddings of
a compact n-dimensional manifold in Rn+k are isotopic if k ≥ n + 2. In simple
terms: when k is this large, there is enough room to untie any n-dimensional
knot!
So, we expect that when k is this large, the n-morphisms in nTangk corre-
spond to ‘abstract’ n-tangles, not embedded in any ambient space. But this is
precisely how we think of cobordisms. So, for k ≥ n + 2, we should expect that
nTangk is a stable n-category where:
• objects are compact framed 0-dimensional manifolds;
• morphisms are framed 1-dimensional cobordisms;
• 2-morphisms are framed 2-dimensional ‘cobordisms between cobordisms’,
• 3-morphisms are framed 3-dimensional ‘cobordisms between cobordisms
between cobordisms’,
and so on up to dimension n, where we take equivalence classes. Let us call this
n-category nCobn , since it is a further elaboration of the 2-category nCob2 in
our discussion of Lawrence’s 1993 paper.
The ‘cobordism hypothesis’ summarizes these ideas: it says that nCobn is
the free stable n-category with duals on one object x, namely the positively
oriented point. We have already sketched how ‘once extended’ n-dimensional
TQFTs can be treated as symmetric monoidal functors
Z : nCob2 → 2Vect.
This suggests that fully extended n-dimensional TQFTs should be something

similar, but with nCobn replacing nCob2 . Similarly, we should replace 2Vect by
some sort of n-category: something deserving the name nVect, or even better,
nHilb.
This leads to the ‘extended TQFT hypothesis’, which says that a unitary
extended TQFT is a map between stable n-categories
Z : nCobn → nHilb
110
n=0 n=1 n=2

x x
•x • /
k=0 x∗
•
•x • / •
x x
x∗ x
x x∗ x • = •
• •L •

• z •
k=1
x x∗ x x∗ x
• • •
• vv•
• 9 r v
z vv
x rr vv

•∗r •v
x x
4d
x∗ x

x x∗ x •7[77 •
• • • 7
k=2
x
•

x
•

4d 5d
x ∗ x

• •O x •

k=3 x • x∗
• x
•
x

•

4d 5d 6d
x ∗
x
x

• •O •

k=4 x • x∗
• x
•
x

•

Figure 2: Examples of n-tangles in (n + k)-dimensional space.

111
that preserves all levels of duality. Since nHilb should be a stable n-category
with duals, and nCobn should be the free such thing on one object, we should
be able to specify a unitary extended TQFT simply by choosing an object
H ∈ nHilb and saying that
Z(x) = H
where x is the positively oriented point. This is the ‘primacy of the point’ in a
very dramatic form.
What progress has there been on making these hypotheses precise and prov-
ing them? In 1998, Baez and Langford [160] came close to proving that 2Tang2 ,
the 2-category of 2-tangles in 4d space, was the free braided monoidal 2-category
with duals on one object. (In fact, they proved a similar result for oriented but
unframed 2-tangles.) In 2009, Schommer-Pries [183] came close to proving that
2Cob2 was the free symmetric monoidal 2-category with duals on one object. (In
fact, he gave a purely algebraic description of 2Cob2 as a symmetric monoidal
2-category, but not explicitly using the language of duals.)
But the really exciting development is the paper that Jacob Lurie [184]
put on the arXiv in 2009. Entitled On the Classification of Topological Field
Theories, this outlines a precise statement and proof of the cobordism hypothesis
for all n.
Lurie’s version makes use, not of n-categories, but of ‘(∞, n)-categories’.
These are ∞-categories such that every j-morphism is an equivalence for j > n.
This helps avoid the problems with duality that we mentioned in our discus-
sion of Atiyah’s 1988 paper. There are many approaches to (∞, 1)-categories,
including the ‘A∞ categories’ mentioned in our discussion of Kontsevich’s 1994
lecture. Prominent alternatives include Joyal’s ‘quasicategories’ [231], first in-
troduced in the early 1970’s under another name by Boardmann and Vogt [26],
and also Rezk’s ‘complete Segal spaces’ [232]. For a comparison of some ap-
proaches, see the survey by Bergner [234]. Another good source of material
on quasicategories is Lurie’s enormous book on higher topos theory [233]. The
study of (∞, n)-categories for higher n is still in its infancy. At this moment
Lurie’s paper is the best place to start, though he attributes the definition he
uses to Barwick, who promises a two-volume book on the subject [235].
Khovanov (1999)
In 1999, Mikhail Khovanov found a way to categorify the Jones polynomial [237].
We have already seen a way to categorify an algebra that has a basis ei for which
X ij
ei ej = mk e k
k
where the constants mij

k are natural numbers. Namely, we can think of these
numbers as dimensions of vector spaces Mkij . Then we can seek a 2-algebra with
a basis of irreducible objects E i such that
X ij
Ei ⊗ Ej = Mk ⊗ E k .
k
112
We say this 2-algebra categorifies our original algebra: or, more technically,
we say that taking the ‘Grothendieck group’ of the 2-algebra gives back our
original algebra. In this simple example, taking the Grothendieck group just
means forming a vector space with one basis element ei for each object E i in
our basis of irreducible objects.
The Jones polynomial, and other structures related to quantum groups,
present more challenging problems. Here instead of natural numbers we have
polynomials in q and q −1 . Sometimes, as in the theory of canonical bases, these
polynomials have natural number coefficients. Elsewhere, as in the Jones poly-
nomial, they have integer coefficients. How can we generalize the concept of
‘dimension’ so it can be a polynomial of this sort?
In fact, problems like this were already tackled by Emmy Noether in the late
1920’s, in her work on homological algebra [236]. We have already defined the
concept of a ‘chain complex’, but this term is used in several slightly different
ways, so now let us change our definition a bit and say that a chain complex
V is a sequence of vector spaces and linear maps
d−1 d0 d1 d2 d3
··· o V−1 o V0 o V1 o V2 o ···
with di di+1 = 0. If the vector spaces are finite-dimensional and only finitely
many are nonzero, we can define the Euler characteristic of the chain complex
by
X∞
χ(V ) = (−1)i dim(Vi ).
i=−∞
The Euler characteristic is a remarkably robust invariant: we can change the

chain complex in many ways without changing its Euler characteristic. This
explains why the number of vertices minus the number of edges plus the number
of faces is equal to 2 for every convex polyhedron!
We may think of the Euler characteristic as a generalization of ‘dimension’
which can take on arbitrary integer values. In particular, any vector space
gives a chain complex for which only V0 is nontrivial, and in this case the Euler
characteristic reduces to the ordinary dimension. But given any chain complex
V , we can ‘shift’ it to obtain a new chain complex sV with
sVi = Vi+1 ,
and we have
χ(sV ) = −χ(V ).
So, shifting a chain complex is like taking its ‘negative’.
But what about polynomials in q and q −1 ? For these, we need to generalize
vector spaces a bit further, as indicated here:
113
vector spaces natural numbers

chain complexes integers
graded polynomials in q ±1
vector spaces with natural number coefficients
graded polynomials in q ±1
chain complexes with integer coefficients
Algebraic structures and the values of their ‘dimensions’
A graded vector space W is simply a series of vector spaces Wi where i ranges

over all integers. The Hilbert–Poincaré series dimq (W ) of a graded vector
space is given by
X
∞
dimq (W ) = dim(Wi ) q i .
i=−∞
If the vector spaces Wi are finite-dimensional and only finitely many are nonzero,
dimq (W ) is a polynomial in q and q −1 with natural number coefficients. Simi-
larly, a graded chain complex W is a series of chain complexes Wi , and its
graded Euler characteristic χ(W ) is given by
X
∞
χq (W ) = χ(Wi ) q i .
i=−∞
When everything is finite enough, this is a polynomial in q and q −1 with integer

coefficients.
Khovanov found a way to assign a graded chain complex to any link in such
a way that its graded Euler characteristic is the Jones polynomial of that link,
apart from a slight change in normalizations. This new invariant can distinguish
links that have the same Jones polynomial [238]. Even better, it can be extended
to an invariant of tangles in 3d space, and also 2-tangles in 4d space!
To make this a bit more precise, note that we can think of a 2-tangle in 4d
space as a morphism α : S → T going from one tangle in 3d space, namely S,
to another, namely T . For example:
α: UU ** →
**
**
*
In its most recent incarnation, Khovanov homology makes use of a certain

monoidal category C. Its precise definition takes a bit of work [239], but its
objects are built using graded chain complexes, and its morphisms are built
using maps between these. Khovanov homology assigns to each tangle T in 3d
space an object Z(T ) ∈ C, and assigns to each 2-tangle in 4d space α : T ⇒ T ′
a morphism Z(α) : Z(T ) → Z(T ′ ).
What is especially nice is that Z is a monoidal functor. This means we
can compute the invariant of a 2-tangle by breaking it into pieces, computing
114
the invariant for each piece, and then composing and tensoring the results.
Actually, in the original construction due to Jacobsson [240] and Khovanov
[241], Z(α) was only well-defined up to a scalar multiple. But later, using the
streamlined approach introduced by Bar-Natan [239], this problem was fixed by
Clark, Morrison, and Walker [242].
So far we have been treating 2-tangles as morphisms. But in fact we know
they should be 2-morphisms. There should be a braided monoidal bicategory
2Tang2 where, roughly speaking:
• objects are collections of framed points in the square [0, 1]2 ,
• morphisms are framed oriented tangles in the cube [0, 1]3 ,
• 2-morphisms are framed oriented 2-tangles in [0, 1]4 .
The tangle hypothesis asserts that 2Tang2 is the free braided monoidal bicate-
gory with duals on one object x, namely the positively oriented point. Indeed,
a version of this claim ignoring framings is already known to be true [160].
This suggests that Khovanov homology could be defined in a way that takes
advantage of this universal property of 2Tang2 . For this we would need to see
the objects and morphisms of the category C as morphisms and 2-morphisms
of some braided monoidal bicategory with duals, say C, equipped with a special
object c. Then Khovanov homology could be seen as the essentially unique
braided monoidal functor preserving duals, say
Z : 2Tang2 → C,
with the property that
Z(x) = c.
This would be yet another triumph of ‘the primacy of the point’.
It is worth mentioning that the authors in this field have chosen to study
higher categories with duals in a manner that does not distinguish between
‘source’ and ‘target’. This makes sense, because duality allows one to convert
input to outputs and vice versa. In 1999, Jones introduced ‘planar algebras’
[243], which can thought of as a formalism for handling certain categories with
duals. In his work on Khovanov homology, Bar-Natan introduced a structure
called a ‘canopolis’ [239], which is a kind of categorified planar algebra. The
relation between these ideas and other approaches to n-category theory deserves
to be clarified and generalized to higher dimensions.
One exciting aspect of Khovanov’s homology theory is that it breathes new
life into Crane and Frenkel’s dream of understanding the special features of
smooth 4-dimensional topology in a purely combinatorial way, using categori-
fication. For example, Rasmussen [244] has used Khovanov homology to give
a purely combinatorial proof of the Milnor conjecture—a famous problem in
topology that had been solved earlier in the 1990’s using ideas from quantum
field theory, namely Donaldson theory [246]. And as the topologist Gompf later
pointed out [245], Rasmussen’s work can also be used to prove the existence of
an exotic R4 .
115
In outline, the argument goes as follows. A knot in R3 is said to be smoothly

slice if it bounds a smoothly embedded disc in R4 . It is said to be topologically
slice if it bounds a topologically embedded disc in R4 and this embedding
extends to a topological embedding of some thickening of the disc. Gompf had
shown that if there is a knot that is topologically but not smoothly slice, there
must be an exotic R4 . However, Rasmussen’s work can be used to find such a
knot!
Before this, all proofs of the existence of exotic R4 ’s had involved ideas
from quantum field theory: either Donaldson theory or its modern formula-
tion, Seiberg–Witten theory. This suggests a purely combinatorial approach to
Seiberg–Witten theory is within reach. Indeed, Ozsváth and Szabó have already
introduced a knot homology theory called ‘Heegaard Floer homology’ which has
a conjectured relationship to Seiberg-Witten theory [247]. Now that there is a
completely combinatorial description of Heegaard–Floer homology [248, 249],
one cannot help but be optimistic that some version of Crane and Frenkel’s
dream will become a reality.
In summary: the theory of n-categories is beginning to shed light on some
remarkably subtle connections between physics, topology, and geometry. Unfor-
tunately, this work has not yet led to concrete successes in elementary particle
physics or quantum gravity. But given the profound yet simple ways that n-
categories unify and clarify our thinking about mathematics and physics, we
can hope that what we have seen so far is just the beginning.
Acknowledgements
We thank the denizens of the n-Category Café, including Toby Bartels, Michael
Batanin, David Ben-Zvi, Rafael Borowiecki, Greg Egan, Alex Hoffnung, Urs
Schreiber, and Zoran Škoda, for many discussions and corrections. JB thanks
the Department of the Pure Mathematics and Mathematical Statistics at the
University of Cambridge for inviting him to give a series of lectures on this topic
in July 2004. He also thanks Derek Wise for writing up notes for a course on this
topic at U. C. Riverside during the 2004-2005 academic year. AL was partially
supported by the NSF grants DMS-0739392 and DMS-0855713.
116
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A Prehistory of n-Categorical Physics

arXiv:0908.2469v1 [hep-th] 18 Aug 2009

John C. Baez∗ Aaron Lauda†

and an output of type y:

To formalize this algebraically, we need something more than a category: at the

To depict this geometrically requires a third dimension, hinted at here by the

way of switching two particles:

representations are a special case of physical theories described as maps between

Our whole progress up to this point may be described as a gradual

From these postulates he derives formulas for the transformation of coordi-

He formalized special relativity by treating space and time as two aspects of

{1, . . . , n}, a process U would be described by a bunch of complex numbers Uji

Von Neumann (1932)

another role in quantum theory. A real-valued quantity that we can measure by

for a unique self-adjoint operator A on H. Conversely, any self-adjoint operator

generated by Lorentz transformations and translations.

gives rise to an observable A with

These are usually called positive-energy representations.

Eilenberg–Mac Lane (1945)

• for any pair of morphisms f : x → y and g : y → z, a morphism gf : x → z

• F preserves identities: for any object x ∈ C, F (1x ) = 1F (x) ;

Many of the famous invariants in algebraic topology are actually functors,

does not admit a retraction—that is, a map

such that this diagram commutes:

cannot have a homomorphism

for which π1 (r)π1 (i) is the identity, because π1 (S 1 ) = Z and π1 (D2 ) = 0.

Given functors F, G : C → D, a natural transformation α : F ⇒ G consists

• a function α mapping each object x ∈ C to a morphism αx : F (x) → G(x)

prototype for the idea of describing theories of physics as functors, in which

The diagrams were intended to represent physical processes and the

G = U(1) × SU(2) × SU(3).

What Mills and I were doing in 1954 was generalizing Maxwell’s

Mac Lane (1963)

of representations of groups. The theory of Feynman diagrams, for example,

This sort of diagram is sometimes called a ‘string diagram’; the mathematics of

• a functor called the tensor product ⊗ : M × M → M , where we write

the left unit law:

such that the following diagrams commute for all objects w, x, y, z ∈ M :

governing the associator.

• the triangle identity:

governing the left and right unitors.

objects morphisms 2-morphisms

In a bicategory we can compose morphisms as in an ordinary category, but also

g : x → y and f : y → z, we have a 2-isomorphism called the ‘associator’:

af,g,h : (f g)h → f (gh).

As in a monoidal category, this should satisfy the pentagon identity.

Monoidal Category Bicategory

and the associative law as:

But it has an enormous advantage: no ambiguity is introduced if we leave out

This is a more vivid way of writing the mathematician’s equation

because tensor products are written horizontally and composition vertically,

f (ρ(g)ψ) = ρ′ (g)f (ψ)

(ρ ⊗ ρ′ )(g) = ρ(g) ⊗ ρ(g ′ )

and the obvious braiding.

(ρ∗ (g)f )(ψ) = f (ρ(g)ψ)

for all f ∈ H ∗ , ψ ∈ H. In our string diagrams, we can use little arrows to

This requires a little explanation. A ‘closed’ diagram—one with no edges coming

or more subtly, like this: