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Design and operation of a microfabricated phonon spectrometer utilizing superconducting tunnel junctions as phonon transducers

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Published 15 April 2013 © IOP Publishing and Deutsche Physikalische Gesellschaft
, , Citation O O Otelaja et al 2013 New J. Phys. 15 043018 DOI 10.1088/1367-2630/15/4/043018

1367-2630/15/4/043018

Abstract

In order to fully understand nanoscale heat transport it is necessary to spectrally characterize phonon transmission in nanostructures. Toward this goal we have developed a microfabricated phonon spectrometer. We utilize microfabricated superconducting tunnel junction (STJ)-based phonon transducers for the emission and detection of tunable, non-thermal and spectrally resolved acoustic phonons, with frequencies ranging from ∼100 to ∼870 GHz, in silicon microstructures. We show that phonon spectroscopy with STJs offers a spectral resolution of ∼15–20 GHz, which is ∼20 times better than thermal conductance measurements, for probing nanoscale phonon transport. The STJs are Al–AlxOy–Al tunnel junctions and phonon emission and detection occurs via quasiparticle excitation and decay transitions that occur in the superconducting films. We elaborate on the design geometry and constraints of the spectrometer, the fabrication techniques and the low-noise instrumentation that are essential for successful application of this technique for nanoscale phonon studies. We discuss the spectral distribution of phonons emitted by an STJ emitter and the efficiency of their detection by an STJ detector. We demonstrate that the phonons propagate ballistically through a silicon microstructure, and that submicron spatial resolution is realizable in a design such as ours. Spectrally resolved measurements of phonon transport in nanoscale structures and nanomaterials will further the engineering and exploitation of phonons, and thus have important ramifications for nanoscale thermal transport as well as the burgeoning field of nanophononics.

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1. Introduction

1.1. Importance of nanoscale phonon spectroscopy

One of the grand challenges of nanoscience is to develop experimental tools to understand the fundamental science of heat flow at the nanoscale [1, 2]. In insulators and dielectrics, acoustic phonons are the dominant heat carriers [3, 4]. In nanostructures, as the sample's dimension or surface morphology becomes comparable to phonon characteristic lengths—wavelength, mean free path and coherence length—the interactions of phonons with these structural features lead to regimes of phonon propagation in which the effect of confinement, scattering and/or interference of phonons dominates heat transport [5, 6]. To probe these nanoscale effects on phonon transport, one needs a measurement technique that can precisely distinguish wavelength (or frequency) and position of the phonon modes. Previous studies have investigated the effects of nanoscale geometries on thermal transport using Joule-heated metal films on suspended structures [711], but because a thermal conductance measurement employs a broad spectral distribution of phonons, the frequency dependence of the phonon transport in such measurements is difficult to distinguish. Therefore, there is a strong need for a nanoscale technique that will spectroscopically measure phonon transport at hypersonic (>1 GHz) frequencies—particularly at frequencies above 100 GHz which are most relevant to heat flow [12]. Such a technique will be apt for the development of the burgeoning field of nanophononics [13, 14].

An ability to fully understand the propagation of phonons will inform the engineering and exploitation of nanostructures and nanomaterials. For instance, through careful phonon engineering the realization of more efficient thermoelectric materials and microelectronic coolers will be feasible [10, 15, 16]. Such phonon engineering strategies have been recently demonstrated with silicon phononic crystal structures, which displayed a reduction in phonon thermal conductivity in comparison to bulk crystals [17, 18]; however, the exact mechanism and frequency dependence of this reduction is not completely understood because diagnostic tools for nanoscale phonon spectroscopy were not available.

In this paper we describe a new tool for nanoscale phonon spectroscopy using microfabricated superconducting tunnel junctions (STJs)—we detail its design and principle of operation, the fabrication techniques and challenges, the instrumentation and measurement procedures, and the results of selected phonon transport measurements. Phonon spectroscopy with STJs uses a narrow, non-thermal and tunable frequency distribution of acoustic phonons to probe the phonon transport through nanostructures. STJ-based phonon spectroscopy has previously been performed extensively in macroscale samples by only a few research groups [1922]. However, with the development in recent years of advanced micro/nanofabrication techniques, the phonon spectrometer can now be fabricated at the microscale and offer exceptional spatial resolution. The microfabricated phonon spectrometer has the advantage of probing nanoscale effects such as phonon confinement [3], end-coupling diffraction [23] and surface scattering [24], with submicron spatial resolution. We have recently demonstrated a prototype microfabricated spectrometer for emission and detection of non-equilibrium phonons with frequencies ranging from 0 to ∼200 GHz [25], and have now tuned the phonon source (emitter) to emit phonons with frequency ranging from 0 to ∼870 GHz. The spectrometer comprises a pair of aluminum–aluminum oxide–aluminum (Al–AlxOy–Al) STJs serving as phonon emitter and phonon detector on opposite sides of a silicon microstructure. The spectrometer measures the rate of phonons that propagate ballistically through the microstructure. Here we discuss in full detail the design, fabrication steps, required characterization, electronics and measurement techniques involved in successfully realizing phonon spectroscopy with microscale STJ phonon transducers.

1.2. Spectrometer design

The device design for each spectrometer consists of two STJ phonon transducers—one emitter and one detector—attached on opposite sides of a mesa that is monolithically etched on a silicon substrate (see figure 1(a)). The mesas, which are ∼0.8 μm high and have widths ranging from 7 to 15 μm, allow for the isolation of a ballistic path for phonon propagation. The devices are fabricated on a 525 μm thick silicon (100) wafer and the mesa sidewalls are on the Si(111) plane. (Because the mean free path of phonons at our experimental temperature and frequencies is ≫1 mm [26], the detected phonons will also include phonons that backscatter from the bottom of the substrate.) The ballistic path along the 〈110〉 direction between emitter and detector may be blocked by etching a trench into the mesa in order to determine this contribution of backscattered phonons [25]. This phonon transport measurement platform also enables the monolithic integration of nanostructures into the mesa. Microfabrication methods make the experiments very scalable—spectrometers are fabricated in lots of 100 on 100 mm Si wafers. Each 4.5 mm square chip contains up to six spectrometers, as shown in figure 1(b).

Figure 1. Refer to the following caption and surrounding text.

Figure 1. (a) False-colored SEM micrograph of completed phonon spectrometer. The STJ emitter is fabricated with the tunnel junction lying mostly on the sidewall of the 0.8 μm high mesa structure. The width of the mesa, R = 7, 10 or 15 μm. The mesa structure isolates a ballistic path for phonon transport between emitter and detector. The detector is fabricated in double-junction SQUID geometry with a hot electron finger for the collection of ballistically propagating phonons. Finger widths, Wf , were varied (1.5, 2, 3 or 6 μm) to observe the effects of geometry on phonon transmission. A magnetic field (∼1 G) is applied perpendicular to the SQUID detector for Josephson current suppression. 0.5 μm thick silver film is deposited on the backside of the 525 μm thick silicon substrate to reduce phonon backscattering from the bottom of the substrate. (The inset is a schematic of the side view of a silicon mesa with phonon transducers.) (b) Optical microscope image of 4.5 mm2 device comprising six spectrometers.

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The phonon emitter is a single Al–AlxOy–Al tunnel junction with the majority of the junction area lying on the sidewall of the mesa. The aluminum films are designed to be thin enough (<100 nm) to ensure that the decay length of the phonons is greater than the film thickness in order to minimize phonon reabsorption (emitted phonons breaking quasiparticles within the emitter film) [27]. As will be described, we isolate narrow bands of phonon energy by modulation of the emitter voltage. Emitter junction resistance should therefore be made low enough to maximize the amount of current (and therefore phonon signal) flowing at a given modulation amplitude, while the resistance must also be large enough to inhibit overinjection of electrons through the tunnel barrier into the Al film. Such overinjection may locally suppress the superconducting gap and thereby degrade energy resolution [28]. Residual inhomogeneities in the gap are inherent in the film and may be assessed from an IV curve of the junction. Typically we use emitters having junction resistances from ∼800 to ∼5000 Ω, and we observe an inhomogeneity of about ∼60 to ∼80 μeV (∼15 to ∼20 GHz), which represents the upper limit of our energy resolution.

The detector is designed to have a double-junction (superconducting quantum interference device (SQUID) geometry) with a 'hot electron finger' extending onto the mesa sidewall to capture the incident phonons. The actual detector junctions lie on the (100) plane of the silicon substrate. The SQUID geometry enables the suppression of Josephson current via the application of a magnetic field. With the Josephson current suppressed, we can readily distinguish an incident phonon flux as an increase in the 'subgap' tunnel current due to the incident phonons breaking Cooper pairs in the Al. Incident phonons of energy ⩾ 2Δ d break Cooper pairs in the aluminum film of the detector finger, and the excited quasiparticles diffuse to the junctions and tunnel through the oxide barrier. Some of the detectors have quasiparticle traps made from thin Au films. These traps are designed to ensure that excess quasiparticle energies do not reach the junction, and they also prevent back tunneling of quasiparticles [29]. When we change the length of the detector fingers, Lf , from 10 to 20 μm (see figure 1(a)), we found no discernible difference in the detected phonon transmission signal levels. We conclude from this that the quasiparticle diffusion length is much longer than the finger length, a conclusion that agrees with diffusion lengths reported in the literature [30]. The tunnel junction therefore faithfully measures the rate of phonon arrival at the tip of the finger several microns distant. Forming the double-junctions on the flat Si(100) plane reduces their asymmetry, therefore facilitating Josephson-current suppression, simplifies fabrication [25], and offers great flexibility in spatial resolution to be achieved merely by changing the width and position of the finger.

2. Principles of operation

2.1. Phonon emission with superconducting tunnel junction

Phonon emission in STJs occurs via the excitation and decay of quasiparticles (single electrons) in superconducting films. As depicted in figures 2(a) and (b), when the emitter STJ is biased above the superconducting gap (2Δ e ) such that Ve ⩾ 2Δ e /e (Ve = Ie Rn , where Ie , Ve and Rn are the current through, voltage across and normal state tunneling resistance of the emitter junction, respectively), the Cooper pairs (paired electrons) in the first aluminum film break apart and quasiparticles tunnel through the oxide barrier into excited energy states ranging from Δ e to eVe − Δ e (all energies referenced to the Fermi level in the aluminum film at the opposing side of the junction) [19, 20]. The Al emitters in our experiments have 2Δ e of ∼400 μeV at a temperature of ∼0.3 K. These excited quasiparticles rapidly decay toward the edge of the superconducting gap, emitting phonons as they decay. Due to the singularity in the density of states at the gap edge, this 'relaxation' process typically requires only one or two decay steps before the quasiparticle energy is reduced to Δ e . This process thus emits a broad distribution of phonons of energies ranging from 0 to eVe − 2Δ e . The phonons are incoherent and to a first approximation will have both random polarization and random direction due to elastic scattering of the tunneled electrons within the Al film. The shape of this 'relaxation' phonon distribution includes a sharp cutoff at energy eVe − 2Δ e , thus allowing a small modulation of Ve to isolate a narrow portion of the spectrum that is sharply peaked at energy eVe − 2Δ e . Subsequent recombination of the quasiparticles into Cooper pairs lead to the emission of recombination phonons of energy 2Δ e . The average relaxation τrel and recombination τrec times are on the order of ∼1 ns and 30 μs respectively [31]. When the STJ emitters are attached to one end of a microstructure, relaxation and recombination phonons—both longitudinal and transverse polarizations—are emitted and ballistically propagate through the microstructure; however, only the relaxation phonons are controlled by modulation techniques for spectroscopic studies.

Figure 2. Refer to the following caption and surrounding text.

Figure 2. (a) Energy diagram illustrating phonon emission with STJs. (b) Schematic diagram of the microfabricated phonon spectrometer. The diagram shows the emitter STJ on the left is biased above its superconducting gap (2Δ e ) leading to phonon emission into the microstructure. The STJ detector on the right, which is biased below its superconducting gap and has its Josephson current suppressed, detects incident phonons with energy greater than 2Δ d . The detector will capture incident phonons within the solid angle subtended by the detector finger with respect to the emitter. (c) Calculation of approximate phonon spectrum due to first-step relaxation of tunneled electrons, for a typical Al emitter STJ biased at ∼2.1 mV to produce peak frequency $(eV_{{\rm e}} - 2\Delta _{{\rm e}} )/h = 400 \,{{\rm GHz }}$ and peak width δ ω /2π = 20 GHz . Differential portion of the spectrum produced by modulation is shaded red. Following phonon attenuation within the Al, the remaining differential spectrum is shaded in dark blue (partially obscuring red-shaded area). (d) Calculation of Ppeak /Ptot for phonon spectrum emitted by quasiparticle relaxation in typical STJ emitter, found by integrating spectrum from part (c) and adding contributions of second-step quasiparticle relaxation, quasiparticle recombination and phonon reabsorption/reemission. For comparison, we calculate also Ppeak /Ptot of Planck distribution for a slice of width $\delta \omega /2\pi = {{\rm 20}}\,{{\rm GHz}}$ around the dominant phonon frequency.

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2.2. Modeling the phonon emission spectrum

When considering spectral precision of a phonon source, a convenient figure of merit is the ratio Ppeak /Ptot of phonon power near the peak of the distribution to the total power in the measurement. For instance, a thermal conductance measurement at temperature T employs a Planck distribution of phonon modes, having power spectral density $P(\omega )\,{{\rm d}}\omega \sim \frac{{\hbar \omega ^3 \,{{\rm d}}\omega }}{{2\pi ^2 v^3 }}/({{\rm e}}^{\hbar \omega /k_{{\rm B}} T} - 1)$ , where ω is the angular frequency of the phonon and v is the phonon speed. This distribution is peaked at the so-called 'dominant phonon frequency' ωdom = 2.78kBT/ℏ , but the distribution is quite broad and therefore a slice of spectrum within δ ω around the peak contains only a small fraction of the total power. For instance, if we wish to interrogate a spectral feature at 400 GHz with 20 GHz precision, a Planck distribution at T = 6.9 K offers ωdom /2π = 400 GHz but contains only 3% of its power within ±10 GHz of this peak.

To model the phonon emission profile of the modulated STJ phonon spectrum and to estimate Ppeak /Ptot of this distribution, we must carefully consider the non-equilibrium electron–phonon interactions within the superconducting film in the emitter STJ. These include phonon attenuation due to Cooper-pair breakage [27, 31, 32] and acoustic-mismatch transmission across the Al/Si boundary [33] as well as quasiparticle diffusion and reemission of absorbed phonons. The total emitted phonon power resulting from the quasiparticle relaxation process will comprise the phonons emitted in first-step relaxation, plus any emitted in second-step relaxation, minus the fraction reabsorbed by Cooper pair breakage within the aluminum, plus the power that is reemitted following this reabsorption processes.

For voltages Ve not greatly exceeding 2Δ e , nearly all injected quasiparticles decay to energy Δ e , so that first-step relaxation dominates, the entire modulated spectral power falls at frequency (eVe - 2Δ e )/h, and if we neglect the effect of reabsorption then $\frac{{P_{{{\rm peak}}} }}{{P_{{{\rm tot}}} }} \cong 1$  [20]. For higher bias voltages, a fraction of the quasiparticles will relax first to intermediate energies before undergoing secondary relaxation to the band edge energy Δ e . The precise distribution of generated phonon energies may be found by convolution integral of the quasiparticle injection rates, densities of states and decay rates [31, 34]. For simplicity, we will assume that the phonon density of states in the Al follows a Debye model, and adopt an approximate model of phonon production rate, presented by Eisenmenger et al based on the spectrum of phonons emitted by first-step relaxation of electrons injected across a normal-state tunnel junction of resistance Rn at T = 0 K [31, 35]:

Equation (1)

This rate of phonon production per unit bandwidth $\dot N_{{\rm e}} (\omega )$ is shown in figure 2(c), and extends from ω = 0 to a sharp cutoff at ω = (eVe − 2Δ e )/ℏ . This is a good approximation for voltages Ve ≫ 2Δ e /e [31, 35]. It is evident from the shape of this distribution that a portion of the differential phonon power is produced at the peak ℏ ω = (eVe − 2Δ e ) while the remainder is produced at energies broadly distributed over the range 0 to eVe − 2Δ e . The power spectral density P(ω) dω may be found from equation (1) as $P(\omega )\,{{\rm d}}\omega = \hbar \omega \dot N_{{\rm e}} (\omega )\,{{\rm d}}\omega$ . Second-step relaxation may add up to 25% additional phonon power, mostly at frequencies well below the cutoff at (eVe - 2Δ e )/ℏ  [31].

We must also consider reabsorption and reemission of phonon energy. Attenuation of the phonon population within the superconductor will occur as phonons of energy $\hbar \omega > 2 \Delta _{{\rm e}}$ break Cooper pairs, creating fresh quasiparticles. The probability that a phonon will survive traveling a distance r within the aluminum is erph , the mean absorption length of Λ ph (ω ) being dependent on phonon energy ℏ ω and band gap energy Δ e  [31]. If we treat the phonons as point-particles traveling ballistically within the Al, then the probability of a phonon generated at a distance z from the Al/Si interface and traveling at an angle θ to the normal, to escape into the Si before reabsorption is [31, 36]

Equation (2)

Here TAlSi (θ ) is an acoustic-mismatch transmission factor for wave transmission from Al into Si. The films of some of our emitter STJs have lower and upper layer thicknesses of ∼20 and ∼79 nm respectively on the mesa sidewall (as determined by profilometry measurement and adjusted for sidewall angle). For simplicity, we treat all phonons as being generated within the lower layer at a spatially uniform rate. We assume the phonons' velocities are distributed uniformly in all directions, and that those entering the top layer may reflect from the Al/vacuum boundary, reenter the lower layer, and reach the Al/Si boundary. For phonons to emerge and travel directly across the mesa toward the detector (an angle ∼35.3° to the sidewall normal), we estimate the refraction angle within the Al using Snell's law, assuming average wave speeds $\overline {v_{{{\rm Al}}} } = 4.4 \times 10^3\,{\rm m}\,{\rm s}^{-1}$ in Al and $\overline {v_{{{\rm Si}}} } = 6.6 \times 10^3 \,{{\rm m}}\,{{\rm s}}^{ - 1}$ in Si, to be θ ∼22.7°. From reported values of the acoustic impedances of Al and Si, we estimate TAlSi to be >0.9 for such an angle and to be frequency independent [6, 31]. Kaplan et al [27] have calculated values for phonon decay time in Al as a function of phonon energy ℏ ω and bandgap energy Δ e . We multiply these by $\overline {v_{{{\rm Al}}} }$ to find Λ ph (100 GHz) ≅ 1.04 μ m , $\Lambda _{{{\rm ph}}} \left( {400\,{{\rm GHz}}} \right) \cong 0.38\,\mu {{\rm m}}$ and Λ ph (700 GHz) ≅ 0.22 μ m . While these values are greater than some reported experimental values of Λ ph in Al at energy ℏ ω = 2Δ e , they are comparable with measured values of normal-state acoustic attenuation corrected to the [3133, 3638] superconducting state. Averaging equation (2) over our full Al layer thicknesses, we estimate that in the direction pointing out toward the detector, ∼90% of phonons at ω /2π = 100 GHz will escape into the Si, ∼78% at ω /2π = 400 GHz and ∼68% at ω /2π = 700 GHz . We use these attenuation factors to modify the spectrum in equation (1), as shown in figure 2(c).

To find the total rate of absorbed phonons, we must average equation (2) over all depths and angles. At large values of θ we note that TAlSi (θ ) will be ≪1, regardless of phonon frequency, and for angles above about 45°, TAlSi (θ ) will be zero due to total internal reflection within the Al [31, 33]. Transmission coefficients 〈TAlSi〉 averaged over all angles and phonon polarizations have been calculated by Kaplan, from which we estimate TAlSi ∼0.44 assuming the three phonon polarizations to be equally populated [33]. Thus at any frequency ω > 2 Δ e /ℏ , at least 56% of all phonons produced are liable to be reabsorbed within the Al. We can approximate the additional frequency dependence by multiplying this TAlSi by the average of equation (2) over the full Al layer thickness and all angles less than the critical angle. Therefore among all phonons at all angles we estimate that ∼61% are reabsorbed at ω /2π = 100 GHz , ∼67% at ω /2π = 400 GHz and ∼71% at ω /2π = 700 GHz . For each bias voltage Ve , we apply these proportions to the spectrum of equation (1) and integrate to find the total reabsorbed power.

By conservation of energy, all of this reabsorbed power must be reemitted. The quasiparticles created in the reabsorption subsequently relax and recombine to emit additional phonons of lower frequency than the ones initially absorbed. We estimate based on typical decay times and on the geometry of our STJ on the mesa sidewall that the quasiparticles do not travel far prior to reemission, so that about 80% of the power is reemitted at the same or nearby location as the original tunneling injection in the Al film on the mesa sidewall.

Taking together first-step relaxation, second-step relaxation (constituting up to ∼25% of the total relaxation phonon power) attenuation and reabsorbed/reemitted power, we find that for typical Ve values of up to a few mV, the total modulated power Ptot emitted from the emitter STJ is roughly proportional to the modulated emitter current δ Ie . The power emitted due to recombination on the other hand (see figure 2(a)) should remain fixed as Ve is varied, and for large Ve we take this to be a negligibly small fraction of the total power. Therefore the total emitted differential phonon rate is ∼δ Ie /e.

To find Ppeak /Ptot at a given peak frequency ωpeak , we take P(ωpeak )δ ω from equation (1), for a given peak width δ ω (e.g. $\delta \omega /2\pi = {{\rm 20}}\,{{\rm GHz}}$ ), attenuate this quantity according to equation (2) as described above, and divide by the total power Ptot found as described above at Ve = (ℏ ωpeak + 2Δ e )/e. The result of this calculation for our typical emitter film thicknesses appears in figure 2(d). For a peak width δ ω /2π = 20 GHz , at a peak frequency of ω /2π = 100 GHz , Ppeak /Ptot is ∼50%. This diminishes to ∼32% at peak ω /2π = 400 GHz , and further at higher peak frequencies. As shown in figure 2(d), the values of Ppeak /Ptot from the STJ-emitted phonon spectrum compare very favorably to a Planck distribution, exceeding it by more than an order of magnitude for ω /2π > 300 GHz . This analysis demonstrates that aluminum STJs made of films a few tens of nm thick will emit narrow spectral distributions of acoustic phonons into Si at frequencies up to several hundred GHz.

Phonon emission from aluminum STJs has been reported elsewhere at frequencies up to ∼2 THz, but Ppeak /Ptot is likely to be very small at such a peak frequency even if the films are made very thin [39]. The wavelength in Al at 700 GHz is ∼6 nm while the granularity in the Al film and the roughness at the Al/Si interface are most likely a few nm; hence, for ω /2π above ∼700 GHz, we expect to see the spectrum further modified by the effects of elastic scattering of phonons within the Al film [36], inelastic phonon scattering at the Al/Si boundary [40] and modification of phonon spectra due to excess injected quasiparticle population in the Al film [28]. All such effects are liable to become more severe as Ve and ω are increased.

2.3. Phonon detection with superconducting tunnel junction

The phonons incident on the detector are registered as an increase in the tunnel current through the detector junctions. The STJ detector is biased below its superconducting gap with voltage Vd < 2Δ d /e (figure 2(b)). Phonons incident on the detector finger with energy greater than or equal to 2Δ d will break Cooper pairs in the detector films, and the quasiparticles will diffuse until a portion reaches the detector junction and tunnel through. The STJ detectors are made from aluminum films with superconducting gap 2Δ d ∼360 μeV (corresponding to ∼90 GHz), and in essence these detectors act as high pass filters of acoustic phonons with cut-off frequency ∼90 GHz. A lock-in detector selects only the modulated portion δ Id of the detector current, corresponding to the modulated emitter phonons that strike the detector. The phonon spectrum therefore comprises phonons of frequencies between ∼90 GHz and (eVe - 2 Δ e )/h, with a sharp peak at frequency $ (eV_{{\rm e}} - 2 \Delta _{{\rm e}} )/h$ . Because the modulated emitter phonon power is proportional to δ Ie , the measured differential transfer function δ Id /δ Ie tells us the fraction of this spectrum that is transmitted from the emitter through the sample to the detector.

2.4. Modeling the detector behavior

We may use quasiparticle–phonon interactions to model and quantify the phonon detector behavior. For a differential rate $\dot n_{{{\rm ph}},{{\rm d}}}$ of phonons of frequency ω striking the detector finger, we expect the average differential rate of phonon-induced quasiparticle generation $\dot n_{{{\rm QP,ph}}}$ to be

Equation (3)

In equations (3), TSiAl is the acoustic transmission factor for phonons transiting from Si into Al, which we estimate from acoustic impedances to be >0.9 over all incidence angles [33]. The fraction of phonons αabs (ω ) absorbed in the finger will be approximately αabs (ω ) = 1 - e2dph(ω) . In our detector fingers, the thickness d in the direction of phonon incidence is 140–205 nm, thus we expect αabs (ω ) to equal at least 0.2 for ω /2π = 100 GHz , and at least 0.8 for ω /2π = 700 GHz . In our devices, the diminishing fraction Ppeak /Ptot as peak frequency is increased (figure 2(d)) motivates us to treat αabs as independent of peak frequency and having value αabs ∼0.25 . In the signal of a typical spectrometer transmitting through bulk Si, we see a modulated signal that is consistent with this assumption and with the detector response behavior of equations (3).

To find $\dot n_{{{\rm QP,ph}}}$ and thereby the phonon arrival rate $\dot n_{{{\rm ph}},{{\rm d}}}$ from the measured differential detector tunnel current δ Id , we must account for quasiparticle loss processes in the detector. The primary loss process comprises diffusion of the quasiparticles into the attached wiring leads, followed by recombination into Cooper pairs [31, 41, 42]. Using conventional theories of tunneling rate and quasiparticle recombination, we may express a nondimensional efficiency factor $\left\{ {{{\rm Eff}}} \right\} = \delta I/e\dot n_{{{\rm QP,ph}}}$ for each detector (see appendix B) [31, 40, 42]:

Equation (4)

where Rn is the normal-state tunneling resistance of the junction, N0 is the normal density of states at the Fermi level (1.75 × 1010 μ m- 3 eV- 1 in Al) [31], and Wtr and dtr are respectively the average total width and thickness of the wiring trace connected to the detector STJ. The factor D≅ 20 cm2 s- 1 is the diffusion constant for quasiparticles in Al, and τrec ∼30 μ s is the average quasiparticle recombination time in Al at a temperature of 0.3 K [30, 31, 41, 43, 44]. In our detectors $\left\{ {{{\rm Eff}}} \right\}$ is typically ∼0.1.

3. Fabrication techniques and challenges

Figure 3(a) illustrates the step-by-step fabrication of the mesas and transducers. The mesas are formed by shallow depth anisotropic etching of silicon using KOH (50% KOH, 48 °C, 4 min) with a low-stress silicon nitride etch mask. We found that standard RCA cleaning of the wafers prior to etching is crucial to obtaining smooth surfaces. The smoothness of the (100) and (111) planes is necessary to enable deposition of continuous Al films, and to minimize phonon scattering from rough surfaces. Simultaneous magnetic stirring and ultrasonication during the KOH etch helps to improve the smoothness of the etched surfaces. The trenches are simultaneously formed, where needed, on the mesas. Neither hydrochloric acid nor surfactants was added.

Figure 3. Refer to the following caption and surrounding text.

Figure 3. (a) Schematic of fabrication steps for STJ phonon transducers and mesas (steps i–x). (i) Low stress silicon nitride (∼80 nm) grown on a silicon substrate, (ii) silicon nitride etched with CHF3/O2 to define mesa, (iii) anisotropic etching in KOH masked by SixNy, (iv) nitride etched in BOE, (v) bilayer of LOR and photoresist spun on mesas and emitter pattern and wiring trace is exposed on the top resist layer, (vi) exposed pattern developed, forming resist bridge across mesa sidewall, (vii) double angle evaporation and oxidation of aluminum followed by lift-off. Tilt axis (T) is indicated. (viii) Bilayer of LOR and photoresist is re-spun and detector is patterned. (ix) Detector pattern is developed forming Dolan bridges. Tilt axis (T) is indicated. (x) Double angle evaporation and oxidation of detector film shown after lift-off. (b) Schematic diagram of double-angle evaporation. T indicates the tilt axis pointing into the page. θ1 and θ2 are the first and second deposition angles forming aluminum films numbered 1 and 2. The first film is oxidized (orange colored region) before the second evaporation. The junction overlap width $= d\left( {{{\rm tan}}\,\theta _1 + {{\rm tan}}\,\theta _2 } \right) - W_1$ . (c) Schematic of overlapping films of the wiring traces with overlap ${\rm width} = W_2 - d\left( {{{\rm tan}}\,\theta _1 + {{\rm tan}}\,\theta _2 } \right)$ . W2 should be made wide enough to ensure overlap.

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We fabricate the emitter tunnel junctions on the sidewall of the mesa using double-angle evaporation as shown in figure 3(b). A bilayer of S1818 photoresist (Rohm and Haas Inc.) and lift-off resist (LOR) (Microchem Inc.) is spun onto the fabricated mesas and the emitter geometry, wiring trace and bond pads are photolithographically patterned into the resist. The depth of field of our photolithography tool (±2.42 μm) limits the range of mesa heights and resist thickness used to form the junctions. The patterned resist is developed in AZ MIF 300 (AZ Electronic Materials) for ∼60 s until a 'Dolan photoresist bridge' is formed with sufficient undercut (see figures 3(a) and (b)) [45]. The surfaces must be cleaned with argon or oxygen plasma prior to evaporation to prevent poor aluminum film adhesion and ageing of the tunnel junctions formed [46, 47].

As shown in figure 3(a), arrays of detector and emitter STJs are patterned and deposited. The film thicknesses in the emitter are made only a few tens of nm (lower layer is ∼20 nm thick and upper layer is ∼38–80 nm thick on the mesa sidewall), to enable phonons to escape into the Si without reabsorption, whereas the detector film thicknesses are made several hundred nm thick to maximize the absorption of incident phonons. We utilize a two-step electron-beam angle evaporation interspersed with a static oxidation procedure to form the aluminum tunnel junctions. The overlap area of the tunnel junctions is dependent on the angles at which the evaporation is done (figure 3(b)). Assuming that the height of the bridge or thickness of the LOR layer is d μm and the width of the bridge is W1 μm, the overlap width is d(tan θ1 + tan θ2 ) - W1 , where θ1 and θ2 are the two deposition angles measured from the normal to the substrate surface (see figure 3(b)). The width of the wiring traces, W2 , should be made wide enough that the double angle evaporation forms a single overlapping metal trace (figure 3(c)). We found that the best quality films were obtained at evaporation rates ∼4.5–5 Å s−1. The films were evaporated at base pressures ranging from 2 × 10–7 to 1.2 × 10–6 Torr. The base pressures were sometimes lowered further by the initial evaporation of 50–100 nm Al in the chamber. The evaporated aluminum acts as a getter for particles in the chamber. The tunnel barrier is formed by static oxidation in between the two Al deposition steps, with an exposure parameter defined as exposure (Pa s) = pressure (Pa) × time (s)  [48]. The emitter tunnel barrier was grown in 3 Torr of oxygen for 60 min resulting in emitter resistances ∼1.5 kΩ. The detector tunnel barrier was grown at 300 mTorr for ∼70 min resulting in resistances of ∼200 Ω. Figure 4(a) shows the exposure parameter plotted against the area-specific resistances of the tunnel junctions. This guide can be used to estimate the exposure parameters that will produce emitters or detectors with desired junction resistances. The plot was fitted to a power law (f(x) = x0.59 ) with an adjusted-R2 value of 0.997. The area of each junction was calculated from scanning electron micrograph (SEM) inspection and can be estimated prior to fabrication based on the overlap area calculations discussed above. The post-evaporation processing includes metal lift-off, dicing of the wafer into 4.5 mm2 chips, and the evaporation of ∼500 nm thick silver on the backside. Silver has been shown to be a good absorber of phonons [26]; hence, the addition of silver reduces the backscattered signals from the bottom of the chip. The junctions are very sensitive to static discharge, and therefore, proper grounding is essential at all times.

Figure 4. Refer to the following caption and surrounding text.

Figure 4. (a) Plot of exposure parameter with the area specific resistance of 11 different tunnel junctions fabricated with five different exposure parameters. Power law fit (f(x) = x0.59 , adj-R2 = 0.997). (b) Plots showing the effect of oxygen doping on aluminum films. We measured the room temperature resistivity of aluminum evaporated in varying oxygen pressures starting with a film deposited at base pressure of 4.4 × 10–7 Torr (no oxygen).

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The base pressure at which the Al films are deposited is important as it may affect their room-temperature resistivity, which in turn affects their critical temperature Tc and superconducting gap. Such variations in the superconducting gaps of aluminum films with respect to their room-temperature resistivity have been reported to be due to oxygen doping [49, 50]. In figure 4(b), we show the dependence of the room-temperature resistivity of thin aluminum films of identical dimensions on increased oxygen partial pressure during evaporation. The dimensions of the films were patterned by photolithography and the first film was evaporated at a base pressure of 0.44 μTorr. By increasing the base pressure due to the continuous flow of oxygen into the chamber, we show that the resistivity of the films varies with base pressure. The typical transition temperature, Tc , for films evaporated at base pressures of 27 and 35 μTorr was measured to be 1.75 and 1.89 K respectively (increased Tc compared to a pure Al film with Tc ∼1.12 K).

4. Instrumentation, measurement technique and characterization of spectrometer

4.1. Low temperature apparatus

The apparatus for the low temperature phonon transport experiments includes a He-3-cryostat with a custom designed sample stage immersed in a liquid helium Dewar. The fridge wiring consists of twisted pair lines with room temperature pi-filters (Tusonix 4701 EMI) enclosed in a brass block Faraday cage, allowing up to ∼90 dB attenuation at frequencies >100 MHz. The cold stage filters are 'tapeworm' type low-pass filters [29], but the extent of cold stage filtering is limited by the space in our vacuum can. The fridge is cooled down to a base temperature of 0.3 K and the sample is held in vacuum. The thermometer at the He-3 stage is a Cernox™ RTD (Lakeshore Cryotronics) and a silicon diode thermometer (DT-470-SD-12A, Lakeshore Cryotronics) monitors temperature at the 1 K pot. Attempts are made to minimize the coupling of noise from the thermometry wiring into the measurement wiring. Metal film resistors are used in all bias networks, as this type of resistor is known to exhibit superior temperature stability and reduced 1/f noise. As shown in figure 5(a), the chips containing the spectrometers are wire-bonded onto the gold plated copper sample stage. The backside of the chips must be properly anchored to the sample stage by thermalizing with Apiezon N grease or silver paint. A 5000 turn superconducting magnetic coil is attached to the top of the sample box for Josephson current suppression as shown in figures 5(b)(d). Once the fridge is immersed in the helium Dewar, we ensure proper grounding of the fridge and equipment rack. We place rubber pads underneath the wheels of the Dewar to reduce mechanical vibration.

Figure 5. Refer to the following caption and surrounding text.

Figure 5. (a) Photograph of 4.5 mm2 chip wire-bonded onto a gold plated sample stage of our He-3 cryostat. (b) Interior photograph of the sample box cover showing magnetic coil (∼5000 turns of superconducting wire wound around a copper core). (c) Schematic of the coil assembly. The coil allows for the application of a magnetic field perpendicular to the chip in order to suppress the Josephson current in the detector. Up to ∼2.5 G of magnetic field may be applied by passing current through the coil. (d) Sample cover mounted on the sample stage prior to cool down.

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4.2. Dc characterization of superconducting tunnel junctions emitters and detectors

The dc characteristics of emitter and detector tunnel junctions are determined from current-biased current–voltage (IV) measurements at ∼0.3 K. Figure 6(a) shows how we estimate the superconducting gap from the IV behavior. From the IV curves we also calculate the normal state resistance, Rn , of the junctions. In figure 6(b), we show the current-biased IV curves in the subgap regime for four SQUID detectors with the current normalized by their normal state resistances for comparison. The red plot shows significant rounding-off, which is due to poor filtering on that particular signal line, allowing stray voltage noise to add a random perturbation to the junction voltage. In figure 6(c), we show the resistance-normalized IV curves for four emitters with normal state resistance values of 212, 935, 2250 and 5559 Ω. This plot illustrates several possible problems in emitter performance. In the 212 Ω emitter (red plot), we observe 'back bending' of the gap rise step at Ve = 2Δ e /e. This is a signature of quasiparticle overinjection, which appears consistently in emitter STJs of Rn < 700 Ω , leading to local suppression of the superconducting gap Δ e and poor phonon energy resolution. In the 2250 Ω junction (magenta plot), the IV curve shows a signature of being partially shorted (this could occur either at their formation or during processing) which will add an uncontrolled thermal phonon population to the junction's emission. The black and blue curves indicate a limitation on emitter energy resolution. For an ideal STJ, the 'gap rise' step at V = 2Δ /e should be infinitely sharp, but in practice, we observe a breadth of ∼60–80 μV (∼15–20 GHz). This behavior most likely indicates that the superconductor's gap Δ e varies within the junction by ∼60–80 μeV (corresponding to a ∼15–20 GHz imprecision in emitted phonon frequency).

Figure 6. Refer to the following caption and surrounding text.

Figure 6. (a) Typical current biased I–V measurement of an emitter. The band gap of the junction 2Δ e is ∼400 μeV. (b) I × RnV curves of four detectors taken without Josephson current suppression and focusing on the subgap regime. All the detectors are SQUIDs. Detector resistances range from 167 Ω (black), 213 Ω (magenta), 849 Ω (red), to 817 Ω (blue). Poorly filtered lines lead to rounding-off of gap rise as shown in the red plot. (c) I × RnV curves of four emitters focusing on the subgap region. Emitter resistances range from 935 Ω (black), 2250 Ω (magenta), 212 Ω (red), to 5556 Ω (blue). The plot illustrates how to identify common emitter problems: magenta-colored plot shows a partly shorted device. The red plot exhibits severe 'back bending' of the gap rise due to overinjection and local suppression of superconducting gap, commonly seen in the case of low emitter resistance. The gap-rise width (black and blue plots) indicates inhomogeneity in superconducting gap 2Δ e , which limits the energy resolution for phonon spectroscopy.

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4.3. Josephson current suppression

Josephson current (or supercurrent) in the detector must be suppressed, so that the detector may be voltage-biased and its quasiparticle tunneling current clearly distinguished. To do so, we apply a magnetic field perpendicular to the SQUID loop, using a small superconducting coil mounted as close as possible to the top of the chip to minimize vibration-coupled flux noise. For our coil geometry (see figure 5(c)), we calculate (using Biot–Savart law) the axial magnetic field to be 1.27 G mA−1. The heat load resulting from typical coil current is ⩽2 μW. The maximum supercurrent in the SQUID detector junction, assuming perfect symmetry, is given as $I_{{\rm c}} (\Phi ) = 2I_{{\rm c}} (0)\left| {\cos \left( {\frac{{\pi \Phi }}{{\Phi _0 }}} \right)} \right|$ , where Φ 0 , Φ and Ic (0) are the flux quantum (2.07 × 10–15 Wb), applied flux and critical current at zero magnetic field respectively [51]. By applying a magnetic flux proportional to $\frac{{n{{\rm \Phi }}_0 }}{2}$ , where n is an odd integer, the supercurrent should be fully suppressed. We typically employ the minimum effective flux (equivalent to n = 1 ), in order to minimize flux trapping. In practice, we find that the supercurrent is not always fully suppressed, probably due to asymmetry between the two junctions. Figure 7(a) illustrates our technique for determining the detector bias point for phonon transport studies. The detector voltage is swept in the subgap regime between ∼ − 300 and ∼300 μV. At each voltage step, the coil current is swept from 0 to 2 mA and the tunnel current is measured at each step. In the three-dimensional plot in figure 7(a), the current measured per detector bias voltage and per coil current is shown. We set the voltage bias point of the detector to ∼Δ d /e (∼180 μV) and coil current to ∼1 mA, where the minimum critical current is obtained. The measured zero-voltage and zero B-field supercurrent for the detector (Rn = 116 Ω ) in figure 7(a) is ∼1.2 μA (z-axis) and is closely predicted by the Ambegaokar–Baratoff expression for T ∼ 0 K, $I_{{{\rm c0}}} = \frac{{\pi \Delta }}{{2eR_{{\rm n}} }}$  [51]. By applying a magnetic field (∼1 G) at the bias point, the supercurrent is suppressed to ∼1 nA.

Figure 7. Refer to the following caption and surrounding text.

Figure 7. (a) Detector Josephson current suppression. The subgap tunnel current (before Josephson current suppression) is measured at each coil current from 0 to 2 mA as the bias voltage is swept from −300 to 300 μV. For operation in spectrometer, the detector is typically biased in the subgap region (∼180 μV) and at an external magnetic field (∼1 G) (1.27 G mA−1) where the critical current is mostly suppressed. The plot also shows the periodic nature of the critical current with applied magnetic field. (b) Plot of $\beta _L \left( {\frac{{2LI_{{{\rm c0}}} }}{{{{\rm \Phi }}_0 }}} \right)$ versus the ratio of the minimum suppressed critical current to the calculated critical current (Ic (min)/Ic0) at T = 310 mK for several SQUID designs. Junctions formed on a flat surface are represented by solid symbols, while the open symbols represent junctions formed on the sidewall; loop areas vary from ⩽2 μm2 (squares), to ∼10 μm2 (circles), to ∼120 μm2 (triangles), and to ∼180 μm2 (diamonds). (c) Temperature dependent subgap tunnel current (after Josephson current suppression) measured at Vd = Δ d /e (180 μV) as a function of temperature (red plot) and estimated tunnel current based on BCS prediction (blue plot).

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The extent to which the supercurrent in the SQUID detectors may be suppressed is dependent on two geometric properties: self-induced flux and junction symmetry. The self-induced flux is proportional to the self-inductance, L , of the SQUID loop, which we estimate based on the inductance of a rectangular loop [52]. The more closely identical the two junctions are, the more closely the current flowing through them may be made to cancel. In figure 7(b), we plot the ratio of the minimum obtainable critical current to the maximum zero voltage critical current (Ic (min)/Ic0 ) versus the parameter $ \beta _L = \frac{{2LI_{{\rm c}} }}{{\Phi _0 }}$ , the ratio of the self-induced flux to the flux quantum. Each symbol in figure 7(b) represents a unique SQUID design based on the location of the junction and the loop area: junctions formed on a flat surface are represented by solid symbols, while the open symbols represent junctions formed on the sidewall; loop areas vary from ⩽2 μm2 (squares), to ∼10 μm2 (circles), to ∼120 μm2 (triangles) and to ∼180 μm2 (diamonds). Smaller loop areas and larger junction resistances lead to smaller values of βL and in general to better supercurrent suppression; however, for the SQUID detectors formed on the sidewall, we observe a large variation in suppression for devices with similar βL . This is likely due to junction asymmetry. For devices formed on the flat (100) surface, supercurrent suppression is more consistent and exceeds ∼3 orders of magnitude for βL < 2 × 10- 3 , indicating more symmetric junction formation. We also note a tradeoff in detector design: while Josephson critical current scales inversely with normal-state tunnel resistance $I_{{{\rm c0}}} = \frac{{\pi \Delta }}{{2eR_{{\rm n}} }}$ , detector efficiency (equation (4)) also scales inversely with Rn . In practice we find that a loop area of ∼2 μm2 and detector resistance Rn ∼200 –300 Ω enable both suppression of Ic to levels smaller than thermal quasiparticle tunneling current, as well as detector efficiencies of ∼0.1 that permit readily measurable spectrometer signals.

With the supercurrent suppressed, we measured the subgap tunnel current due to thermally excited quasiparticles at detector voltage Vd = Δ d /e and at different temperatures (∼0.3–0.4 K) as shown in figure 7(c) (red plot). We compare the results to the Bardeen, Cooper and Schrieffer (BCS) approximation of the subgap tunnel current for an S–I–S junction (blue plot) [53]. The measurement shows exponential dependence of subgap current on temperature, as predicted by BCS theory. The deviation between the data and prediction may be due to our inability to fully suppress the supercurrent and to possible inaccuracies of our cold stage thermometer at temperatures below ∼0.34 K.

4.4. Modulated phonon transport measurements

The schematic of our phonon transport experiments is shown in figure 8(a). For phonon emission (Ve ⩾ 2Δ e /e) , the emitter is current biased by applying a dc voltage, $V_{{\rm b}} = \frac{{V_{{\rm e}} }}{{R_{{\rm n}} }}R_{{\rm b}}$ through bias resistor Rn ∼500 kΩ , where Ve (=Ie Rn ) is the voltage across the emitter junction and Rn is the normal state resistance of the emitter junction. All the device wiring comprises filtered twisted-pair lines, and shielded coaxial cables are used for all connections. The dc current through the emitter junction is stepped from Ie = ∼0.35 –2 μA, which corresponds to emitter voltages Ve = ∼0.35 –2 mV for a junction resistance of $R_{{{\rm n}} } = 1\,{{\rm k}}\Omega$ . In addition to the dc current applied to the junction, an ac modulation current $\delta I_{{{\rm e }}} \sim20\,{{\rm nA}}_{{{\rm rms}}}$ is applied by adding an ac modulation δ Vb to the dc level Vb through a unity-gain isolation amplifier (Burr Brown ISO124P) and 100 ×  voltage divider; the output is independent of frequency between 4 and 1000 Hz and exhibits noise of ∼10− 6 V Hz- 1/2 . The typical modulation frequencies for our measurements range between 7 and 11 Hz.

Figure 8. Refer to the following caption and surrounding text.

Figure 8. (a) Schematic of phonon spectroscopy measurements. (b) Steady state detector current, Id . (c) Differential transfer function δ Id /δ Ie , representing the fraction of emitted phonon flux that reaches the detector. The emitter tunnel junction turns on above Ve = 2Δ e and emits detectable relaxation phonons only above Ve = (2Δ e + 2Δ d )/e. For Ve = (2Δ e + 4Δ d )/e, the emitted relaxation phonons may break multiple Cooper pairs in the detector. The peak at ∼4 mV represents resonant backscattering of oxygen impurities in the Si, typically seen at ∼870 GHz [55]. (d) Voltage-biased detector IV curves with varying emitter voltages and partially-suppressed Josephson current. (e) Differential conductance calculated from the IV measurements in (d) (colors are the same as in (d)). (f) Equivalent circuit model of detector as a current source. The dc current Id and modulated current δ Id follow the incident flux of phonons.

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For phonon detection, the detector is voltage biased in the subgap regime (Vd ∼Δ d /e) with the Josephson current suppressed. The detector signal comprises a steady state plus a modulated component, as indicated in figure 8(a). The steady state dc detector current Id = ∼1 –2.5 nA for emitter voltages $V_{{\rm e}} = 0.35$ –5 mV as shown in figure 8(b). For dc detector tunnel currents Id up to 1.5 times the unperturbed (thermal) level of the steady state detector current, we treat τrec as being constant and therefore equation (4) as being valid and { Eff} being fixed [42, 54]. We checked this assumption by raising the device temperature until Id rose by a factor of 3, and observed very small change in the differential transfer function δ Id /δ Ie . Thus for Id < 1.5 times its thermal level, we can safely assume that the detector response remains linear with incident phonon flux. (We note that for Id > 1.5 times its thermal level, the detector response may be nonlinear with the incident phonon flux.) In our devices τrec may be limited by magnetic flux trapped in the Al detector film as well as by quasiparticle population [42].

The modulated ac detector current (also differential response or differential transfer function) of our detector (figure 8(c)), which represents the modulated portion of the incident phonons, is isolated via a low-noise current pre-amplifier (DL 1211) and a lock-in amplifier (SRS 830) over a range from 0 to ∼1 pArms. As shown in figure 8(c), the emitter tunnel junction turns on at emitter voltage above Ve = 2Δ e /e: the step in detector response at Ve = (2Δ e + 2Δ d )/e occurs because the emitted relaxation phonons (peak energy = eVe - 2Δ e ) above this voltage are energetic enough to break Cooper pairs in the detector (gap energy 2Δ d , i.e. ∼90 GHz). When Ve = (2Δ e + 4Δ d )/e, we observe a further change in detected signal level, as the emitted relaxation phonons acquire enough energy to break multiple Cooper pairs in the detector (see also equation (3)). We have also considered the effect of microwave Josephson radiation on the detector signal [30]. In one spectrometer, we biased the emitter at Ve = 0  V and modulated the Josephson branch of the emitter IV curve. We observed zero detector response. We conclude that our measurement is not influenced by Josephson radiation or inductive coupling of the emitter Josephson current into the detector.

The peak frequency of the emitted relaxation phonon distribution is related to the emitter bias voltage as (eVe - 2Δ e )/h. The feature in figure 8(c) at Ve ∼4 mV is believed to be due to backscattering by oxygen impurities in the silicon. This peak was observed at ∼870 GHz in past studies of STJ phonon spectroscopy [55, 56]. While this behavior confirms that our aluminum STJ-based spectrometer emits a strong and tunable signal well above 800 GHz, we note that at such high frequencies (figure 2(d)), we estimate only ∼20% of the total phonon power to be at the peak frequency of $(eV_{{\rm e}} - 2\Delta _{{\rm e}} )/h$ .

In figure 8(d), we present voltage-biased IV curves of a detector recorded while varying the emitter voltage from 0 to ∼5 mV. (We note that in this detector we were unable to suppress Josephson current below ∼5 nA.) For emitter voltage $V_{{\rm e}} = 0 \,{{\rm V}}$ , the subgap current at detector voltage Vd = Δ d /e (180 μV) is exactly the same as that shown in figure 7(c) at a temperature of ∼313 mK. As a larger and larger phonon flux is transmitted to the detector, the total quasiparticle density in the detector increases well beyond the thermal level, and the detector current rises. In figure 8(e), we calculate the differential conductance (dI/dV) from the subgap IV measurements of figure 8(d). The conductance of the detector remains essentially the same as emitter voltage is varied. At the typical bias point of Vd = Δ d /e, conductance G remains fixed at ∼5 × 10–6 Ω−1. The only difference is in the total current level.

These measurements motivate a simplified equivalent circuit model for our STJ phonon detector, shown in figure 8(f). The phonon detector is modeled as a current source in parallel with a resistance $ 1/G$ . The dc current Id and modulated current δ Id follow the incident flux of phonons. The detector is in series with the current amplifier (input impedance RAMP and current through IAMP ) and line resistance RLINE . The bias point on the detector is maintained by an isolated voltage source (Stanford Research SIM928, output through a 105 voltage divider) across the entire network. Typical values for RLINE and RAMP are ∼70 Ω and 2 kΩ respectively (RAMP is the manufacturer's specification). This model, and the measurements of figures 8(d) and (e), makes clear that the STJ maintains a steady bias throughout our measurement range—even if Id rises by 1 nA, the bias across the STJ will change by only a few μV. Similarly, the current through the amplifier, accurately registers the modulated current δ Id through the detector. Modulated amplifier current δ IAMP equals δ Id /(1 + G(RLINE + RAMP )) , which is only ∼1% different from δ Id for typical values of RLINE , RAMP and G .

5. Results of phonon spectroscopy measurements

5.1. Energy resolution and sensitivity

The energy resolution of our measurement is limited by noise, by the band gap inhomogeneity of the emitter STJ and by the modulation amplitude. Voltage noise across the emitter STJ adds random fluctuations to bias voltage Ve , while inhomogeneity in the emitter gap Δ e likewise reduces precision of phonon energies. In practice, we assess these effects based on the width of the gap rise in the emitter IV curve (figure 6(c)), typically ∼60–80 μeV. The modulation current δ Ie applied to the emitter may also reduce energy resolution by adding a voltage oscillation of peak amplitude $2\sqrt 2 R_{{\rm n}} \delta I_{{\rm e}}$ to the emitter voltage Ve = Ie Rn . For typical emitter junction resistance Rn ∼ 800 Ω and δ Ie ∼ 20 nArms , this modulation envelope is only ∼40 μeV, and therefore the bandgap inhomogeneity imposes the limit on energy resolution: ∼60–80 μeV—corresponding to a frequency resolution ∼15–20 GHz.

The sensitivity of the measurement is limited by detector noise, which may comprise electrical pick up noise, vibrational pickup in wiring and amplifier noise, as well as fundamental contributions such as Johnson noise in wiring and shot noise in the tunnel junction. Figure 9 shows a typical noise spectrum of the detector, exhibiting peaks in the spectrum at 60 Hz and its multiples due to power-line noise pickup, as well as an unexplained resonance at ∼600 Hz. Wiring and apparatus to minimize noise are discussed in the section on instrumentation. Based on detector noise spectra such as figure 9, we typically choose modulation frequencies between 3 and 12 Hz, adding line-frequency notch filters and low-pass filters at the input of the preamplifier and lock-in amplifier to avoid amplifier overload. The lowest noise level obtained at modulation frequency of 11 Hz was ∼ 60 fA Hz- 1/2 . We note that a tunnel junction passing a dc current of 1 nA should exhibit a shot noise of ∼ 18 fA Hz- 1/2 (assuming a Fano factor of 1), so our experimental noise is not far above the shot noise level. To reduce uncertainty in a spectral measurement, we typically repeat it 25 times and average the results. Considering the typical detector efficiencies $\left\{ {{{\rm Eff}}} \right\}$ ∼0.1 (equation (4)) as well as acoustic-transmission and absorption factors TSiAl and αabs (see equation (3)), we estimate the noise equivalent power (NEP) for phonon detection to be ∼ 10− 15 W Hz− 1/2 , or ∼2 × 107 phonons of energy $\sim 2 \Delta _{{\rm d}}$ per second per $\sqrt {{{\rm Hz}}}$ . A comparative analysis of similar low temperature thermal detectors found similar sensitivities [57].

Figure 9. Refer to the following caption and surrounding text.

Figure 9. Detector noise spectrum. Modulation frequency for transport measurements ranges from 3 to 12 Hz.

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5.2. Ballistic phonon propagation

The ballistic nature of phonon transport is evidenced by comparing the differential detector response (δ Id /δ Ie ) of spectrometers with varying mesa widths, detector finger widths, blocked ballistic path and offset line-of-sight between emitters and detectors (figures 10(a)–(d)). To enable measurements made with different detectors to be compared equivalently, we divide each measured value of δ Id /δ Ie by $\left\{ {{{\rm Eff}}} \right\}$ for that detector to obtain the phonon transmission signal. Following equations (3) and (4), we expect the resulting scaled value to equal $T_{{{\rm SiAl}}} \,\alpha _{{{\rm abs}}} \frac{{2e\delta \dot n_{{{\rm ph,d}}} }}{{\delta I_{{\rm e}} }}$ for 2Δ d ⩽ ℏ ω < 4Δ d and $T_{{{\rm SiAl}}} \,\alpha _{{{\rm abs}}} \frac{{2e\delta \dot n_{{{\rm ph,d}}} }}{{\delta I_{{\rm e}} }}(\hbar \omega /2{{\rm \Delta }}_{{\rm d}} - 1)$ for ℏ ω ⩾ 4Δ d . Since TSiAl and αabs are expected to be roughly the same from one detector to another, we do not rescale the data for these factors. We note that the quasiparticle diffusion length $\sqrt {D\tau _{{{\rm rec}}} }$ is of order 100 μ m , so that phonons reflected from the bottom of the Si chip and striking the wiring leads far from the junction or the mesa may also contribute to a measured 'background' signal level that is also subject to the same efficiency $\left\{ {{{\rm Eff}}} \right\}$ as the signal resulting from phonons striking the detector finger [40]. The rate of ballistic phonons striking the detector finger, as measured by the differential detector response, is proportional to

Equation (5)

where Ae is the fraction of emitter STJ visible from the detector, cos  θ is a Lambert law phonon emission distribution, Afoc (θ ,ϕ ) is the phonon focusing factor, TAlSi and TSiAl are acoustic transmission factors described previously, and dΩ d is the solid angle subtended by the detector with respect to the emitter STJ [31, 5861]. Figure 10(a) shows the phonon transmission signal between emitter and detector formed on different widths of mesas (7 μm (blue) and 10 μm (red)) with 6 μm detector finger widths. As the mesa width increases from 7 to 10 μm, the solid angle Ω d subtended by the detector with respect to the emitter decreases; hence, the differential detector signal decreases as expected. We further verified the ballistic phonon transmission by varying the width, Wf , of the detector fingers. For a 10 μm mesa, we show the phonon transmission signal for a 6 μm wide (red plot) and a 3 μm wide (magenta plot) detector finger (figure 10(b)). The wider finger will subtend a larger solid angle; hence, the detector signal is larger as expected for the 6 μm wide detector finger shown in figure 10(b). In figure 10(c), we blocked the ballistic path between the emitter and detector by etching a trench into the mesa. The mesa width and detector finger widths are 10 and 6 μm respectively for both the bulk (open circles) and trench (hatched circles). The latter measurement reveals a significant portion of the transmitted phonon signals that are due to backscattering from the bottom of the chip ('background signal'). The difference between the trench transmission and the transmission through the mesa represents the dynamic range of our measurements. In figure 10(d), we compare the phonon transmission signals for emitters and detectors that have a straight line-of-sight along the mesa width (along the 〈110〉 crystal direction, solid green plot) and with emitter and detectors offset with line-of-sight by ∼50° (near to the 〈100〉 crystal direction, open green plot). A slightly higher detector signal level is observed for the offset geometry. For this geometry, the ballistic signal is affected by phonon focusing—the attenuation or enhancement of phonon propagation in the preferred direction in an anisotropic crystal such as silicon [60]. In silicon crystals, the phonon focusing factor is ∼2 times higher in the 〈100〉 direction than in the 〈110〉 direction [61]. These measurements evince the sensitivity of our phonon spectrometer to submicron variations in device geometry. We point out, however, that the measured differential response of the detector must be scaled by the efficiency factor $\left\{ {{{\rm Eff}}} \right\}$ in order to compare measurements from different detectors. In figure 11, we replot the results in figure 10(a) (phonon transmission through different mesa widths) with the unscaled detector response δ Id /δ Ie , and we show that with typical detector efficiency factors { Eff} ∼0.1, there is an order of magnitude difference between the scaled and unscaled signals.

Figure 10. Refer to the following caption and surrounding text.

Figure 10. (a) Ballistic phonon transport measurements on different mesa widths. Detector signal level decreases as mesa width increases from 7 μm (blue) to 10 μm (red). The detector finger width is 6 μm in all cases. (b) Ballistic phonon transmission with varying detector finger width. Detector signal collected by the 6 μm detector finger (red) is higher compared to the 3 μm detector finger (magenta). Mesa width is 10 μm in both cases. (c) Plots comparing phonon transmission through a mesa with (hatched red circles) and without (open red circles) a trench etched into the mesa. The trench blocks the line of sight between the emitter and detector. Mesa width is 10 μm and detector finger is 6 μm wide in both cases. (d) Ballistic phonon transport measurement with varying angle between emitter and detector. In the solid green circle plot, the emitter and detector have a straight line of sight, but in the open green circle plot the emitter and detector are offset by ∼50°. The mesa width is 7 μm and detector finger width is 3 μm in both cases. (Plots in this figure are not restricted to the region where the detector response is linear with incident phonon flux, i.e. for portions of the plot Id > 1.5 times its thermal level.)

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Figure 11. Refer to the following caption and surrounding text.

Figure 11. Comparison of the modulated detector response scaled by detector efficiency factor $\left\{ {{{\rm Eff}}} \right\}$ (scatter plots and left axis) with the unscaled detector response (line plots and right axis) for plots in figure 10(a). The measured detector response must be scaled by the efficiency factor in order to ascertain the phonon transmission signal. (Plots in this figure are not restricted to the region where the detector response is linear with incident phonon flux, i.e. for portions of the plot Id >  1.5 times its thermal level.)

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6. Conclusion

We have designed and microfabricated a phonon spectrometer utilizing STJ transducers for the emission and detection of hypersonic (100 to ∼870 GHz) acoustic phonons in silicon microstructures. We model the phonon emission profile of the modulated STJ phonon spectrum considering the electron–phonon interactions within the superconductor films of the emitter STJ, and we also model the phonon detector behavior by considering quasiparticle–phonon interactions. Our energy resolution of ∼60–80 μeV, corresponding to a frequency resolution of ∼15–20 GHz, is about 20 times better than the energy resolution obtainable from conventional thermal transport measurements, which rely on a Planck distribution of phonons. We have demonstrated that with a phonon detection noise equivalent power, NEP, of 10- 15 W Hz- 1/2 , the sensitivity of our STJ phonon detectors is comparable to similar low temperature thermal detectors that are available. The design of our spectrometer—comprising a silicon mesa with STJs on the sides—serves as a good platform for phonon transport studies. The ballistic phonon transmission through the mesa alone can be distinguished from backscattering from the substrate by subtracting the mesa-with-trench phonon transmission signal from the mesa-without-trench signal—a method which eliminates the need for more complicated suspended structures as is typical for thermal conductance measurements. The silicon mesa platform is adaptable to studies of phonon transmission through nanostructures or nanomaterials by etching or depositing these into the ballistic path defined by the mesa. Finally, we have evinced spectrally resolved ballistic phonon transport in microstructures with submicron spatial resolution. Our STJ-based spectrometer provides a state-of-the-art tool for examining nanoscale effects on phonon transport.

Acknowledgments

The authors thank R B Van Dover, S Baker and Cornell LASSP for loan of key equipment. We thank R B Van Dover, R Pohl and K Schwab for helpful discussions and thank N J Yoshida, J Chang and A Lin for help with apparatus and procedures. The work was supported in part by the National Science Foundation under agreement no. DMR-1149036, and in part by the Cornell Center for Materials Research (CCMR) with funding from the Materials Research Science and Engineering Center program of the National Science Foundation (cooperative agreement no. DMR 1120296). MA was fully funded and OOO was partially funded through support of the Energy Materials Center at Cornell (EMC2), an Energy Frontier Research Center funded by the US Department of Energy, Office of Science, Office of Basic Energy Science under award no. DE-SC0001086. This work was performed in part at the Cornell Nanoscale Facility, a member of the National Nanotechnology Infrastructure Network, which is supported by the National Science Foundation (grant no. ECS-0335765). The authors declare no competing financial interests.

Appendix A.: Numerical example of phonon emission rate

As a numerical example, a typical aluminum STJ having normal-state tunnel resistance $R_{{\rm n}} = 1000\, \Omega$ and biased at ∼2.1 mV to produce a peak at 400 GHz, with a ±10 GHz modulation, will produce ∼4 nW of total phonon power and about ∼0.4 nW of modulated phonon power. Because of the geometry of our spectrometer, only about 0.1% of this modulated power, or ∼0.4 pW, will participate in the measurement. Of this about ∼32%, i.e. roughly $5 \times 10^8 \,{{\rm phonons}}\,{\mathop{\rm s}\nolimits} ^{ - 1}$ , will be carried by the peak phonons in the 20 GHz band around 400 GHz; the remainder of the power (roughly 1–$3 \times 10^9 \,{{\rm phonons}}\,{{\rm s}}^{ - {{\rm 1}}}$ ) is carried by phonons of energy lower than the peak. In contrast, a thermal source peaked at 400 GHz and emitting the same experimental power (∼0.4 nW) will emit a similar fraction (0.4 pW) in the proper direction to participate in the experiment, but only ∼3% of this, or roughly $5 \times 10^7 \,{{\rm phonons}}\,{{\rm s}}^{ - {{\rm 1}}}$ , will be carried by phonons within ±10 GHz of the peak. Of the remaining power, roughly half will be carried by phonons of frequency >410 GHz and half by phonons of frequency <390 GHz.

Appendix B.: Estimating the detector efficiency

The measured differential tunnel current δ I in our detector will be proportional to the change in nearby quasiparticle density δ NQP  [31, 40, 42]:

Equation (B.1)

where Rn is the normal-state tunneling resistance of the junction, N0 is the normal density of states at the Fermi level (1.75 × 1010 μ m- 3 eV- 1 in Al), and the last factor reduces to 1.15 at our detector bias voltage V = Δ d /e [31]. Equation (3) of the main text presents the differential rate of quasiparticle generation $\dot n_{{{\rm QP,ph}}}$ as a function of differential rate $\dot n_{{{\rm ph,d}}}$ of phonons incident on the detector. From this $\dot n_{{{\rm QP,ph}}}$ , we can determine the differential change in quasiparticle density δ NQP by the steady-state assumption that the rate of quasiparticles generated must balance all quasiparticle loss rates. The primary loss process comprises diffusion of the quasiparticles into the attached wiring leads, followed by recombination into Cooper pairs [41]. We will assume that the tunneling itself does not contribute significantly to quasiparticle loss. For quasiparticles diffusing into a volume vol , the recombination loss rate is [31, 42]

Equation (B.2)

The recombination time τrec is strongly sensitive to the total quasiparticle density NQP = NQP,th + NQP,dc + δ NQP , where NQP,th is the thermally-activated quasiparticle density, NQP,dc is the quasiparticle density due to the full rate of incident phonons and δ NQP is due to modulated incident phonons. However, as long as NQP,dc + δ NQPNQP,th , we may treat τrec as constant [54]. At a temperature of 0.3 K, τrec is roughly 30 μs [31, 43, 44]. To check the dependence of detector response on NQP , we repeated one of our spectral measurements at a temperature of 0.36 K, at which Id was three times its value at 0.3 K. We found that the detector response was degraded by only ∼10% compared to the 0.3 K measurements. Thus, we expect that restricting Id to only 1.5 times its unperturbed (thermal) value should maintain the condition NQP,dc + δ NQPNQP,th , and therefore maintain a consistent detector sensitivity. We note that the ∼10% reduction upon raising the temperature to 0.36 K is less than what would be predicted by the theory of Rothwarf and Taylor [54], suggesting that in our devices τrec is less temperature-dependent than this theory. One possible explanation is that magnetic flux trapped in the Al detector film contributes to the quasiparticle recombination rate in our detectors [42]. In some cases, cycling our devices above Tc resulted in variations of a few per cent in the measured phonon transmission signal, which is consistent with the presence of detector efficiency variations due to trapped flux.

In considering $\dot n_{{{\rm QP,rec}}}$ , the volume vol primarily comprises the wiring trace attached to the finger, so we have ${{\rm vol}} \cong W_{{{\rm tr}}} d_{{{\rm tr}}} \sqrt {D\tau _{{{\rm rec}}} }$ , where Wtr and dtr are respectively the average total width and thickness of the trace, which in our devices are respectively 3.2 μ m and 530–580 nm, and $\sqrt {D\tau _{{{\rm rec}}} }$ is the diffusion length of the quasiparticles. For diffusion constant D = 20 cm2 s−1, this length is $\sim {{\rm 25}}0\,\mu {{\rm m}}$  [30, 41]. Thus the recombination rate found from equation (B.2) is $\dot n_{{{\rm QP,rec}}} = - \delta N_{{{\rm QP}}} W_{{{\rm tr}}} d_{{{\rm tr}}} \sqrt {D/\tau _{{{\rm rec}}} }$ . In steady-state we take the total rate of change of quasiparticle density to be zero, thus $\dot n_{{{\rm QP,ph}}} + \dot n_{{{\rm QP,rec}}} = 0$ , and we find

Equation (B.3)

Thus the tunnel current may be related to the rate of quasiparticle generation by incident phonons found from equation (3) of the main text:

Equation (B.4)

From equation (B.4) we may define a nondimensional efficiency factor { Eff} for each detector as the ratio of measurable current δ I to charge production rate $e\dot n_{{{\rm QP,ph}}}$ :

Equation (B.5)

We note that the relatively large magnitude of quasiparticle diffusion length $\sqrt {D\tau _{{{\rm rec}}} }$ (of order 100 μ m ) means that phonons reflected from the bottom of the Si chip and striking the wiring leads very far from the junction or the mesa may generate quasiparticles that register as a tunneling current at the detector STJ, therefore contributing to the measured backscatter signal level. It is interesting to think about whether we could reduce or eliminate the measured background level (which represents a source of experimental uncertainty) by redesign of the detector or wiring traces. However, we note from equations (B.2)–(B.5) that changes in the wiring trace dimensions may not achieve this goal: if we reduce the width Wtr of the wiring traces in order to diminish the intercepted phonon flux, we also reduce the volume vol occupied by the quasiparticles and thereby increase the tunneling efficiency $\left\{ {{{\rm Eff}}} \right\}$ for both the quasiparticles formed in the finger and those formed in the leads.

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10.1088/1367-2630/15/4/043018