Show Summary Details
More options …

Zeitschrift für Naturforschung A

A Journal of Physical Sciences

Editor-in-Chief: Holthaus, Martin

Editorial Board Member: Fetecau, Corina / Kiefer, Claus / Röpke, Gerd / Steeb, Willi-Hans

12 Issues per year

IMPACT FACTOR 2016: 1.432

Cite Score 2016: 1.35

SCImago Journal Rank (SJR) 2016: 0.368
Source Normalized Impact per Paper (SNIP) 2016: 0.594

Online
ISSN
1865-7109
See all formats and pricing
More options …

Nanoscale Thermal Transfer – An Invitation to Fluctuation Electrodynamics

Carsten Henkel
Published Online: 2016-12-17 | DOI: https://doi.org/10.1515/zna-2016-0372

Abstract

An electromagnetic theory of thermal radiation is outlined, based on the fluctuation electrodynamics of Rytov and co-workers. We discuss the basic concepts and the status of different approximations. The physical content is illustrated with a few examples on near-field heat transfer.

1 Introduction: Motivation

The papers in this special issue illustrate that for the management of heat and energy transfer in nanotechnology, we are awaiting fundamental challenges. It may suffice to mention Moore’s law, according to which on a scale of 5–10 years, the microscopic structure of semiconductor junctions will become relevant. We will have to cope with statistical and thermal fluctuations that are significant compared to nominally specified (mean) values. At the same time, it is likely that the design of novel materials with tailored properties in photon and phonon transport opens up new avenues like ‘thermal computing’, and raises the efficiency of thermoelectric and thermophotovoltaic devices.

In this paper, we intend to provide a gentle introduction into a statistical description of thermal radiation and radiative heat transfer that may guide the engineer from the microscale down to the nanoscale. The key tool is already more than 60 years old; it has been dubbed ‘fluctuation electrodynamics’ and was developed in the 1950s by physicists in the former Soviet Union: Rytov [1], Ginzburg [2], Dzyaloshinskii et al. [3], [4], just to mention a few. A selection of more recent reviews are [5], [6], [7], [8], [9], [10], [11], [12], [13]. The main idea may be explained by analogy to Brownian motion: instead of characterizing the motion of a particle by its thermal energy, one introduces microscopic trajectories that are perturbed by randomly fluctuating forces. The forces arise from the interaction to a thermal environment, for example between a liquid and a colloidal particle immersed in it. Since the seminal work of Perrin, Langevin, and Einstein [14], it has become clear that an ensemble of such fluctuating trajectories can recover (or ‘unravel’) thermal equilibrium statistics and may even describe systems driven out of equilibrium, for example by temperature gradients or external forces.

The radiation field coupled to matter at finite temperature may be described in a similar way. Fluctuation electrodynamics thus takes serious Planck’s idea that thermal equilibrium for blackbody radiation arises from the exchange of energy with the ‘cavity walls’. The latter realise a macroscopic material system whose ‘internal temperature’ imposes the steady-state properties of the radiation field. The field is indeed unable to thermalise by itself (excluding extreme situations like the relativistic electron–positron plasma during the first minutes of the universe). The equations of fluctuation electrodynamics are involving, in an essential way, the response of the material: dispersion and absorption. It is well known since the 19th century that this can be modelled using material parameters like permittivity, permeability, and conductivity. The main insight of the current trend towards the nanoscale is that this ‘macroscopic description’ can be extended in a quite natural way to scales smaller than the typical wavelengths of thermal radiation (around 1 μm, say). There is a relatively large window of scales (the realm of nanotechnology) where an atomistic description of matter is not yet needed – it may be dubbed the ‘mesoscopic range’. Let us compute as a rough estimate the ratio between the radiation and matter degrees of freedom. Fixing a cube of size L3 of a medium with refractive index n, say the number of photonic modes in a frequency band is

$ρraddω∼ω2L3(nc)3dω,$(1)

while the microscopic degrees of freedom are given by the atomic number density ϱ/m

$ρmatdω∼ϱL3mωDdω.$(2)

We have assumed for simplicity that the phonon modes are distributed evenly up to the Debye frequency ωD . Using estimates for a cubic crystal, the photonic mode density is much less by a small factor ~(v/c)(a/λ)2≪1, where v is the speed of sound, a the unit cell size, and λ the photonic wavelength. This illustrates that matter can indeed play the role of a macroscopic reservoir for the radiation field.

For the purposes of this review, we focus on the simplest macroscopic electrodynamics where the material response is assumed to happen locally. Ohm’s law, for example, is written as j=σE, where all quantities are evaluated at the same (r, ω). In metals, this approximation is expected to work on spatial scales longer than the mean free path of charge carriers and the Fermi wavelength. To the same level of approximation, interfaces between materials are assumed to be ‘sharp’. The permittivity ε(r, ω), for example, varies like a step function across the boundary. This is accompanied by suitable boundary conditions for the macroscopic fields. Descriptions that go beyond this picture would use spatial dispersion (non-local response) and genuine response functions for the interface region like a surface current or a surface polarisation [15], [16], [17]. A discussion of how to formulate fluctuation electrodynamics in non-local media can be found in [18], [19], [20], [21].

The last approximation that is commonly applied is the linearity of the medium response. For material parameters like ε, linear response is written into their definition. The physics beyond this regime is very rich and contains effects like rectification of thermal noise, generation of higher harmonics, temperature-dependent material parameters, hysteresis, and so on. Some of these aspects are already studied in nanoscale thermal devices, in the proposals for thermal diodes or rectifiers, for example. Others are being explored theoretically [22], [23]. From a mathematical viewpoint, the linear approximation provides a simple link between the probability distributions of the Langevin forces (matter-related sources of the field) and of the field, leading eventually to Gaussian statistics, which are completely characterised by mean values and the correlation spectra.

A number of experiments have recently been performed in nanoscale heat transfer; one of the oldest we are aware of is from the Dransfeld group [24]. This was using a typical scanning setup, involving a sharp probe and a sample. Strong enhancements of heat transfer in the near field have been demonstrated in similar settings [25], [26], [27]. Experiments in a planar geometry, closer to theorists’ favourite (Fig. 1), have been reported in [28], [29], [30]. Further enhancement of heat exchange and its spatial modulation were observed with coated or ‘island-decorated’ samples [31], [32]. We shall comment on a few other recent references later in this paper. As a side remark, we mention pump–probe techniques where thermal dynamics become accessible on short time scales, comparable to the equilibration time between electron and phonons [33]. They may provide a complementary approach to a better understanding of heat transport, where the transient response of matter is in the focus.

Figure 1:

Typical setting of a heat transfer experiment on the nanoscale. The blue and red lines give the temperature profile in the two bodies (left: hot, right: cold). The heat current q is carried by conduction in the body and by the difference of thermal emission in the gap. The blue and red shadings illustrate the evanescent waves in the near field of the surfaces; they provide additional ‘tunnelling channels’ for heat transfer. This can lead to an amplification by a few orders of magnitude above the Stefan–Boltzmann law.

The sketch in Figure 1 is meant to fix some ideas: two bodies are described as extended objects with sharp boundaries. Inside, a local temperature can be defined, although its variation may be weak despite a sizeable heat current. The bodies are separated by a vacuum gap across which photons are ‘propagating’ and ‘tunnelling’ to transport energy. Direct contact, which would lead to phononic and electronic transport, can be excluded by working at distances above a few angstroms.

2.1 Maxwell–Langevin Equations

Within the approximations outlined in Section 1, the radiation field in a mesoscopic medium is described by the following set of macroscopic Maxwell equations:

$∇⋅εE=ρ,∇×H−iωεE=j,∇⋅μH=0,∇×E−iωμH=0.$(3)

Here, ε=ε(r, ω) and μ describe the electric and magnetic response of the medium. The charge and current densities can be represented by (‘external’) polarisation and magnetisation fields

$ρ=−∇⋅P, j=−iωP+∇×M.$(4)

In the case of non-magnetic and local conductors, the key response function is the conductivity σ, and (3) and (4) apply with ε=ε0/ω, μ=μ0, and M=0. In some applications to thermal radiation in metals, a ‘background’ dielectric response (due to the ionic cores) is taken into account, replacing ε0 by εb inside the conductor.

The boundary conditions for the fields at a smooth interface follow from the Maxwell equations (3) when understood in the sense of distribution theory: a jump in the normal component of the displacement field εE, for example, is given by the surface charge density, i.e. a δ-sheet of ρ(r), localised at the interface. Similar techniques may be applied when the response functions vary rapidly near the interface on the spatial scales relevant for the fields. One may introduce a ‘macroscopic’ model for the interface where the material equations are extrapolated from the two bulk media to the surface. The deviations from this extrapolation are called ‘excess charges’ or currents, and are represented by δ-distributions localised on the (macroscopic, idealised) interface. For details, see [15], [16].

The key idea of fluctuation electrodynamics, developed by Rytov and co-workers [1], is that the sources ρ, j are random or fluctuating fields. This is why we shall call (3) the Maxwell–Langevin equations. If the sources are non-zero on average, their values would be interpreted as ‘external’ charges as in the conventional macroscopic electrodynamics. The fluctuations around the average arise from the thermal motion of carriers in the medium and are therefore determined by its temperature and its ‘oscillator strength’ (or density of states). Within the macroscopic scheme, as explained above, it is natural to apply local thermodynamic equilibrium statistics to describe the charge and current fluctuations.

2.2 Source (Current) Fluctuations

To illustrate these concepts, let us take as an example the current fluctuations in a conductor and assume for simplicity that the mean value 〈j(r, ω)〉=0. The fluctuations of the current define its noise spectrum, as follows:

$〈jm∗(r′, ω′)jn(r, ω)〉=2πδ(ω′−ω)〈jm(r′)jn(r)〉ω.$(5)

The writing 〈…〉ω on the rhs is adopted to avoid additional notation, although it is, strictly speaking, slightly abusive. This quantity provides the (power) spectral density, i.e. the Fourier transform of the autocorrelation function (Wiener–Khintchine formula)

$〈jm(r′, t′)jn(r, t)〉=∫−∞∞dω2πeiω(t′ − t)〈jm(r′)jn(r)〉ω.$(6)

Note that the factor 2π in (5) is tied to this convention for the Fourier transform and that such a definition makes sense in a stationary situation only (correlations only depend on the time difference). The complex Fourier expansion used here is appropriate for ‘quantum detectors’ like a photomultiplier. Glauber’s theory of these detectors works with the Hermitian field operator E(r, t) and shows that the signal involves the convolution of E with a complex exponential e−iωt , where ω>0 is the threshold frequency (work function). In classical electronics, the signal is convolved with a real-valued reference or circuit response. In that case, an autocorrelation function based on a symmetrised operator product is relevant. Its spectrum is symmetric and is equal to the average of its ‘quantum cousin’ in (6) over positive and negative frequencies.

Thermal equilibrium and local macroscopic electrodynamics yield the following relation between spectrum and conductivity:

$〈jm(r′)jn(r)〉ω=2ℏωReσ(ω)eℏω/kBT−1δmnδ(r′−r)≡Sj(r, ω)δmnδ(r′−r),$(7)

where both T=T(r) and σ may vary spatially (local temperature). The real part (the conductance) of the complex admittance σ is related to ohmic dissipation, hence the name fluctuation–dissipation (FD) relation for (7). This formula is the quantum extension of Johnson–Nyquist noise; see the seminal paper by Callen and Welton on the FD theorem [34] and the review by Ginsburg [2]. The current spectrum Sj (r, ω) is plotted in Figure 2 inside a Drude conductor, σ(ω)=σDC/(1+iωτ). The plot covers both positive and negative frequencies: note the asymmetry of the spectrum. This is generic for operator products that are not in symmetric order and hence a typical quantum effect. Indeed, with our convention for the spectral density, quantum (or vacuum) fluctuations become visible at negative frequencies. The scheme required to detect these would be based not on absorption but on emission (spontaneous and stimulated); see, for example, [35]. The quantum current fluctuations in a Drude conductor show a broad Debye-like peak at ω~–1/τ, where τ is the electronic relaxation time. (The mean free path is vF τ with vF the Fermi velocity.) Another convention is to take a symmetrised correlation function so that the spectrum becomes symmetric (grey dashed lines).

Figure 2:

Spectral density Sj (ω) of current fluctuations inside a Drude conductor (7) for positive and negative frequencies. The spectrum is normalised to its low-frequency limit at T=300 K [see (8)]; the temperature dependence of σDC is neglected for simplicity. The frequency is normalised to the Drude relaxation rate 1/τ (we took ℏ/τ=10 meV). The dashed lines give the spectrum for the symmetrised autocorrelation of the current.

At lower frequencies, the current noise is white and given by the Nyquist formula

$〈jm(r′)jn(r)〉ω≈2kBTσDCδmnδ(r′−r).$(8)

It is interesting to note that for typical Drude parameters and room temperature, the ‘thermal energy–time uncertainty product’, kB /ℏ, is not far from unity. In other words, good conductors show strongly damped electrical currents at room temperature.

2.3 Field Correlations

The thermal radiation field generated by the Rytov–Langevin equations (3) depends linearly on its sources. Solutions to the homogeneous equations can be used to describe radiation incident from infinity [36]. Some authors prefer to cancel these solutions by allowing for a non-zero imaginary part in ε(r, ω) in all space (see, e.g. [37]). We follow this route and get for bodies in local equilibrium the following expression for the electric field correlation function [defined as in (5), summation over double indices]:

$〈Ek(r)El(r′)〉ω=∫d3sGkm∗(r, s, ω)Glm(r′, s, ω)Sj(s, ω).$(9)

Here, Gkl (r, s, ω) is the electromagnetic Green tensor, which gives the electric and magnetic fields radiated by a point current source located at s:

$Ek(r, ω)=∫d3sGkl(r, s, ω)jl(s, ω),$(10)

$Hk(r, ω)=−iωμ(r, ω)ϵklm∂∂rl∫d3sGmn(r, s, ω)jn(s, ω).$(11)

As a function of r and the index k, Gkl satisfies the boundary conditions at macroscopic interfaces. Many different conventions for prefactors in (10) and (11) appear in the literature. Explicit expressions are available in free space [38], in a planar interface [39], [40], in multi-layer structures [41], and in spherical [42] and cylindrical [43] geometries. The book by Chew [44] provides an exhaustive discussion, including numerical techniques from computational electrodynamics. Let us simply note that in practice, it is possible to avoid the volume integration in (9) and get instead an integral over the surfaces of the bodies [9], [45], [46]. (The price to pay is the restriction to spatially constant temperatures inside the bodies.) For objects with a highly symmetric shape, an expansion into orthogonal vector field modes is possible (‘principle of sufficient symmetry’), reducing (9) to a summation over products of mode functions. It is now possible to compute Green functions in quite arbitrary geometries using finite-element techniques called ‘discrete-dipole approximation’ [47], [48], [49]. These techniques permit naturally to implement the concept of local, thermal current sources; examples of recent achievements can be found in [7], [50], [51].

From a correlation function like 〈Ek (r)El (r′)〉ω , it is easy to compute the electromagnetic energy density (take k=l and r=r′), correlations for the magnetic field (take the curl with respect to both r and r′), and the Poynting vector, for example. All quantities naturally appear in the spectral domain and can be split into contributions originating from the different objects. In this way, heat currents from body A to B and back can be determined by computing the flux of the Poynting vector across a surface that separates A and B [see (20)]. The elements of the electromagnetic stress tensor [38], plotted locally, permit visualising forces (momentum transfer) between objects, providing physical insight into dispersion and Casimir forces; see [52] for examples. A seminal example of this problem is Lifshitz’ derivation of the van der Waals force between macroscopic bodies [3], [53].

2.4 Quantum Field Theory

The Maxwell–Langevin approach sketched so far provides the equations of motion of the field (operators), and relevant averages (correlation functions) are obtained from those of the source currents. The question of what would be the Hamiltonian of this theory has been raised. The quantisation scheme for the macroscopic Maxwell equations developed by the Knöll and Welsch group in Jena [37], [54] gives an answer in line with the ideas of Rytov and co-workers [1]: the information contained in the macroscopic medium response functions (dielectric function ε,...) is sufficient. We adopt here a slight rewriting of the Jena equations and keep working with the current operators j(r, ω). The basic idea is to upgrade each Fourier component to a dynamic variable. The medium+field Hamiltonian is then given by the sources alone

$HMF=∫0∞dω2π∫d3rj∗(r, ω)⋅j(r, ω)2Reσ(ω).$(12)

It is interesting that the power density of Joule absorption provides the energy (12) of medium and field. The Heisenberg equations of motion yield the seemingly trivial time evolution ~e−iωt provided the commutator of the currents is chosen as

$[jm(r′, ω′), jn∗(r, ω)]=4πℏωReσ(ω)δmnδ(r′−r)δ(ω′−ω).$(13)

The relation j(r, ω)=j*(r, –ω) is imposed as initial condition and is preserved during the evolution. The local thermal equilibrium ensemble is generated by a density operator proportional to the exponential of HMF, taking the local temperature field T(r) under the spatial integral in (12). The FD relation with its correlation function (7) then follows.

The observable electric field is given by the source current convolved with the Green function (10), and this relation can be understood as a generalised mode expansion. In this way, the Maxwell–Langevin equations are satisfied. The coupling to external sources (atoms or molecules) proceeds in the usual way by adding interaction terms to HMF. A multipolar coupling scheme is quite natural; for details, see the book by Buhmann [54]. An important consistency check is to recover the Pauli–Jordan commutator between the electric and magnetic fields:

$[Ei(r, t), Bj(r′, t)]=−iℏε0ϵijk∂∂xkδ(r−r′).$(14)

This is achieved by using (13) for the source currents, the Kramers–Kronig relations for the Green function, and the following identity [55], [56]:

$∫d3sRe(σ)Gik∗(r, s)Gjk(r′, s)=−ReGij(r, r′),$(15)

which follows from the classical macroscopic Maxwell equations. The frequency arguments were suppressed for simplicity, and the negative real part appearing on the rhs is due to the convention adopted here for the Green tensor [compared to the energy conservation law (19) below].

2.5 Discussion

Let us finish this basic introduction with a remark on physical interpretation. A cursory scan through the literature yields different answers as to the actual status of the current fluctuations appearing in (3) and (7). One may consider them as ‘mathematical artefacts’ whose only merit is to reproduce, in thermal equilibrium, the spectrum of the radiation field [57], similar to thermostatting Langevin forces in molecular dynamics. It has been checked that equilibrium is indeed recovered if all bodies are at the same temperature: the field then follows Planck’s blackbody spectrum, modified by the bodies’ emissivity and enhanced significantly when near-field components become detectable at subwavelength distances [57]. Right from its inception, however, fluctuation electrodynamics have been applied to non-equilibrium settings, too [4], where this reference situation is no longer available. One is then inclined to adopt the viewpoint that Rytov’s fluctuating currents are a model for ‘actual thermal fluctuations’ of matter observables that are relevant as electromagnetic sources – an approximate model, of course, as in the formulation presented above, the local macroscopic theory has been applied. However, the model has some reasonable footing given the large number of microscopic constituents that make up a condensed matter system on a mesoscopic length scale of, say, a few nanometres [see (2)]. The key assumption is, of course, that on this scale, the concept of a local temperature T(r) makes sense, and that the electromagnetically relevant quantities have fluctuations that are equilibrated at this temperature. Such a model is eventually required in any description, as described insightfully by Barton [58]:

One could, of course, try to pursue the further question of how the Langevin forces themselves might be kept functioning as envisaged; to explore this would then entail hypothesising some thermostats controlling the forces, and calculating the forces instead of making assumptions about them a priori; and so on, potentially ad infinitum. In the end, one must necessarily settle for control at some level through thermostats that are external to the system, in the sense that they impose temperatures by fiat, through dynamics that do not enter the calculation and are not spelled out.

The following formula may illustrate the problem of local equilibrium:

$jtot(r, ω)=σE(r, ω)+j(r, ω).$(16)

This sum (sometimes called the ‘Rytov split’) gives the current as the sum of Ohm’s law (‘induced current’), while the second term gives the fluctuations around this value (‘fluctuating current’, sometimes marked with subscript fl). In a non-equilibrium setting like in heat transfer, the field E(r, ω) is, in general, not in a state of local equilibrium because it contains radiation from distant sources at different temperatures. The current is then also not in local equilibrium, because its induced part is simply the linear (and local) response of the medium to this non-equilibrium field. However, the fluctuations around it (fluctuating current) are characterised by the local temperature. This has been recognised as the key assumption of fluctuation electrodynamics, for example by Barton [58], [59] who comments in [58] on Loomis and Maris [60]:

[...] who do not assume local thermal equilibrium, but only that in each half-space the noise is appropriate to the temperature of its thermostat.

It is, of course, another problem how local thermal equilibrium is established in the material, and whether this happens fast enough to talk of a single temperature for all frequencies of the local current noise. A practically minded engineer might want to use an equation of motion for the temperature distribution T(r) using a combination of heat conduction and heating/cooling by absorption/emission of thermal radiation. It seems quite obvious that such a ‘self-consistent’ model requires a separation of time scales between the ‘fast’ electromagnetic fluctuations and the ‘slow’ evolution of the temperature field. Needless to say that for many experiments on near-field radiative transfer, it is sufficient to assume a stationary temperature profile, which may even be spatially flat (across each body) if the thermal conductivity is large enough. For a counter-example on short time scales, see the pump–probe experiments quoted above [33].

As a quick estimate, let us consider a typical nanoscale heat flux that aims at competing with the solar constant, q=1360 W/m2. Thermal conductivities of condensed matter are in the range κ=1…102 W/Km at room temperature. On the spatial scale L of a microdevice, one thus estimates a temperature difference

$ΔT=LΔTL∼Lqκ∼10−5…10−3 K×(L/1 μm),$(17)

which is probably below the precision of typical thermometry. In a nanojunction, the temperature profile is thus ‘step-like’ (see Fig. 1). The radiation field in the gap between the bodies has no well-defined temperature, of course. This is similar to radiation in the upper atmosphere, which is a superposition of solar and terrestrial sources.

3 Examples

We now present a few examples of results from fluctuation electrodynamics. Two view points will be stressed: the differences to radiative transfer, which describes thermal radiation on larger scales [61], and challenges beyond the mesoscopic theory presented so far.

3.1 Typical Features on the Nanoscale

Thermal radiation is a classical topic in astrophysics that determines the luminosity and the inner structure of stars. The cosmic microwave background also provides an example where the blackbody spectrum is a very accurate description.

As one works with condensed matter sources, deviations from the Planck spectrum become apparent. Some of these are well known and are described by the classical Kirchhoff concepts of emissivity and absorption (‘grey bodies’) [61]. As one enters the near field, the changes in the spectrum become dramatic because surface resonances may appear in the range of thermal frequencies. Polar materials provide a generic example where optical phonons hybridise with the electromagnetic field into surface phonon polaritons. These resonances enhance the electromagnetic energy density, and they also polarise the radiation field. An example is shown in Figure 3 for a planar body: at the wavelength 11.36 μm, silicon carbide has surface resonances (because the permittivity is negative), which show up as a prominent ‘shoulder’ at distances dλ/10. At even shorter distances, the energy density follows a law 1/d2, which can be explained by the electrostatic fields of fluctuating dipole moments. The surface resonance appears as an enhanced amplitude of this power tail.

Figure 3:

Electric energy density spectrum as a function of distance (top panel) and degree of polarisation (bottom panel). Parameters for SiC: ε(11.36 μm)=–7.6+0.4 i and ε(9.1 μm)=1.8+4.0 i; for glass ε(500 nm)=2.25+10−3 i. The energy densities follow a power law 1/d2 at short distances. They have been normalised to the Planck spectrum. Figure adapted from ([11], fig. 8).

The polarisation of the near field has been quantified in the plot by the following degree of polarisation:

$P=|ℰ∥−ℰzz|ℰzz+2ℰ∥, 0≤P≤1,$(18)

where εzz (ε||) is proportional to the spectrum of the electric field perpendicular (parallel) to the surface, and by symmetry ${〈{E}_{x}^{2}\left(r\right)〉}_{\omega }={〈{E}_{y}^{2}\left(r\right)〉}_{\omega }.$ Note the strong polarisation when the surface resonance is excited. In the limits of short and large distances, a partial polarisation P≈1/4 arises, but for different reasons: at short distances, it follows from electrostatics that the normal field (Ez ) has a spectrum twice as large as the parallel field. At large distances, the polarisation is due to reflection and emission at glancing angles from the surface, similar to the Brewster effect. This partial polarisation is actually an artefact of assuming a planar source of infinite extent. For a discussion of the thermal emission of a spherical source, see, for example [64], [65].

Let us mention that to measure the energy density, a local measurement with a small pick-up antenna is needed. This can be achieved with sharp tips that scatter the near field [27]. The scattered signal is dominated by the immediate environment of the local probe: a lateral resolution comparable to the distance d would be typical. In this way, the thermal radiation may provide a scanning image of local electromagnetic properties and resolve metallic nano-objects on a dielectric substrate, for example.

3.2 Energy Balance and Radiative Heat Transfer

The radiative heat flux between two objects is described in electromagnetism by the Poynting vector. In our context, it has a natural spectral representation and involves the Planck spectrum as an essential factor, due to the assumption of local thermal equilibrium of the radiation sources. The symmetry of photons propagating to the left and the right is broken by the difference in temperature. As the bodies are approached to subwavelength distances, the heat flux is enhanced due to near-field coupling.

Let us start with the energy conservation law for the Maxwell–Langevin equations. In a stationary situation (the energy density is constant), one gets

$∇⋅Re〈E×H〉ω+ωImε(ω)〈E2〉ω=−Re〈j⋅E〉ω$(19)

for each frequency ω. The first term on the lhs involves the radiative emission spectrum (average Poynting vector 〈Sω ); the second term, the absorption in the medium. (With our sign convention, passive media have ω Im ε(ω)≥0.) On the rhs, we find the power transferred to the fields by the mechanical motion of the sources [38]. Equation (19) has been commented upon in a discussion of energy conservation in a dissipative system [66]: the electromagnetic energy lost by absorption in the medium is ‘replenished’ by the sources in the same medium, as long as local equilibrium holds. This observation illustrates the general idea of treating dissipative quantum systems with Langevin dynamics: the fluctuating forces maintain the system energy at its equilibrium value. By the same token, they prevent the commutators of the system observables from decaying.

In a typical setting of heat transfer between two bodies, we may evaluate (19) in a vacuum gap between them and get ∇·〈Sω =0. In a planar geometry, this means that the normal flux is constant. We quote here the formula that results from evaluating the mixed electric and magnetic correlation function between two planar bodies, ‘hot’ and ‘cold’, at distance d [60], [67]:

$〈Sz〉ω=[Φ(ω, Th)−Φ(ω, Tc)]∑p∫kdk(ω/c)2AhpAcp|eikzd1−rhprcpe2ikzd|2,$(20)

$Φ(ω, T)=ℏω34π3c21eℏω/kBT−1,$(21)

where Φ(ω, T) is the flux of a blackbody. Two polarisations p are summed over, and the quantity

$0≤Abp={1−|rbp|2if k≤ω/c2Imrbpif k>ω/c$(22)

is proportional to the absorption of a wave incident on body b=h, c where it is reflected with amplitude rbp=rbp(k, ω). Finally, ${k}_{z}=\sqrt{{\left(\omega /c\right)}^{2}-{k}^{2}}$ is the normal component of the wavevector in vacuum. For k>ω/c, it is purely imaginary (evanescent wave, Im kz ≥0), the heat then transports by photon tunnelling from the bulk of one body to the other. The denominator $1-{r}_{hp}{r}_{cp}{\text{e}}^{2\text{i}{k}_{z}d}$ in (20) describes the multiple reflection of waves between the two interfaces; its zeros correspond to ‘cavity modes’ and surface resonances. In practice, due to imperfect reflections, the modes are broadened but are still visible in the spectrum. See Figure 4 for an example, taken from the seminal paper by Polder and van Hove [67].

Figure 4:

Spectrum of heat flux between two metallic slabs, contribution of propagating waves [0≤kω/c in (20)]. Upper pair of curves: electric field in the plane of incidence (p- or TM-polarisation), lower pair: electric field perpendicular to plane of incidence (s or TE). Solid lines: distance d=10 μm, dashed lines: d→∞. The peaks arise when d is an integer multiple of λ/2 (standing waves). The bodies are at temperatures Th =315 K=Tc +1 K, and have identical conductivities described by a Drude model with σDC=3.9 MS/m and damping time τ=7.1 fs. Figure taken from [67], fig. 4. ©American Physical Society. Reprinted with permission.

The enhancement of the heat flux in the near field is mainly due to the contribution of evanescent waves (photon tunnelling). Taking the limit of purely non-absorbing bodies, one finds that the k-integral in (20) is limited to k/c, where n is the refractive index: in this case, the Planck spectrum (Stefan–Boltzmann law) still holds with a modified prefactor (angle-averaged emissivity). Figure 4 illustrates that the propagating modes that dominate in this range mainly redistribute the emission spectrum by forming cavity resonances. The qualitatively new power laws seen in Figure 3 emerge from deeply evanescent waves, kω/c. These correspond, on the one hand, to surface polaritons with typically k~1/d where d is the distance, and on the other, to diffusion ‘modes’, k~(μ0σDCω)1/2, where (μ0σDC)−1 is the magnetic diffusion constant in a conductor. From an analysis of the two polarisations, these contributions have been pointed out already by Polder and van Hove [67].

4 Challenges

For recent theoretical developments in fluctuation electrodynamics, see other papers in this issue. We conclude this discussion with two challenges that may lead to a refined theory. By analogy to the unexplained perihelion shift of Mercury about 100 years ago, we suggest to name the following observations ‘anomalies’.

4.1 Casimir Force and Conductivity

The first anomaly has been dubbed the ‘plasma vs. Drude controversy’ in the community working on dispersion forces and the Casimir effect [68], [69]. We focus here on an experiment, performed by Chen et al. [70], that does not seem to have received much attention in this discussion. In the setup, an atomic force microscope is measuring the force between a sphere (diameter ~200 μm) and a planar body. The body is a semiconductor membrane (silicon) and gets irradiated with a train of laser pulses that increases the carrier density. The density is practically stationary over the ~5-ms duration of the pulses, and the force is measured with a lock-in amplifier at the repetition rate. From the viewpoint of macroscopic electrodynamics, the irradiation changes the conductivity by a few orders of magnitude from quite small (intrinsic, related to defects, ~10 S/m) to high (~1 MS/m), while the temperature increase is negligible.

The calculated electromagnetic force agrees with the experimental data within 10%, but only for the laser-doped material. A statistically significant difference to the theory (about 50% at distance d≈100 nm) is observed for the intrinsic Si membrane. The deviation can be pinpointed to the behaviour of the material conductivity across the thermal spectrum (see [71], [72] for a discussion). Technically speaking, frequency integrals are performed as summations over the imaginary axis (Matsubara sum), using the analytical continuation of response functions. The anomaly arises from the zero-frequency term in the sum. The current discussion about this thermal Casimir anomaly has led to experiments with magnetic materials (nickel) where the deviation to theory amounts to a few orders of magnitude [73]. An explanation for this anomaly would turn precision measurements of forces in the submicrometre range into tighter constraints on gravity on small scales (including extensions of the standard model like the ‘fifth force’ and additional dimensions) [74], [75], [76].

4.2 Heat Transfer Below 10 nm

Two recent experiments on heat transfer between sharp, metallised tips and a substrate, held at different temperatures, suggest a similar anomaly when compared to fluctuation electrodynamics [77], [78]. The setups differ mainly in the size of the tips (radius of curvature). The group of Reddy and Meyhofer found good agreement with theoretical calculations done by Kim et al. [77], within uncertainties related to surface roughness. Distances down to a few nanometres were reached (tip radius ≈450 nm). Kloppstech et al. observed a heat flux much larger than theory (computed by Biehs and Rodriguez) [78], using a sharper tip (radius ≈30 nm) in a similar range of distances. The data show an onset of the ‘giant heat flux’ at d≈5…6 nm, with a roughly linear increase as d is reduced. According to calculations, a significant fraction of the heat flux should originate from the ‘shaft’ of the tip (conical shape with ~300 nm height). It has been checked that the difference between experiment and theory is not reduced when a non-local conductivity (spatial dispersion) is applied [78]. The solution to this anomaly in short-range heat transfer is currently under investigation [79], [80]. Molecular dynamics simulations like those performed by Domingues et al. [81] and Chiloyan et al. [82] are likely to provide a versatile tool to ‘bridge the gap’ between the mesoscopic and microscopic scales and to include, for example, heat transport from phonons [83].

5 Conclusion

We hope that the present introduction to fluctuation electrodynamics provides a convenient ‘entry point’ into this powerful method. As it happens with any other physical theory, it must be used having its limitations in mind. Some of these are quite natural and related to convenient approximations like a local response. The limits set by experiments that give different results are pointing towards new challenges and may open up directions for further development.

References

• [1]

S. M. Rytov, Y. A. Kravtsov, and V. I. Tatarskii, Elements of Random Fields (Principles of Statistical Radiophysics vol. 3), Springer, Berlin 1989. Google Scholar

• [2]

W. L. Ginsburg, Fortschr. Phys. 1, 51 (1953) [Uspekhi Fiz. Nauk 46, 348 (1952)]. Google Scholar

• [3]

I. E. Dzyaloshinskii, E. M. Lifshitz, and L. P. Pitaevskii, Sov. Physics Usp. 4, 153 (1961). Google Scholar

• [4]

E. M. Lifshitz and L. P. Pitaevskii, Statistical Physics (Part 2), 2nd ed. (Landau and Lifshitz, Course of Theoretical Physics vol. 9), Pergamon, Oxford 1980. Google Scholar

• [5]

P. Ben-Abdallah and S. A. Biehs, AIP Adv. 5, 053502 (2015). Google Scholar

• [6]

B. Song, A. Fiorino, E. Meyhofer, and P. Reddy, AIP Adv. 5, 053503 (2015). Google Scholar

• [7]

X. Liu, L. Wang, and Z. M. Zhang, Nanoscale Microscale Thermophys. Eng. 19, 98 (2015). Google Scholar

• [8]

A. C. Jones, B. T. O’Callahan, H. U. Yang, and M. B. Raschke, Progr. Surf. Sci. 88, 349 (2013). Google Scholar

• [9]

I. A. Dorofeyev and E. A. Vinogradov, Phys. Rep. 504, 75 (2011). Google Scholar

• [10]

A. Narayanaswamy, S. Shen, L. Hu, X. Chen, and G. Chen, Appl. Phys. A 96, 357 (2009). Google Scholar

• [11]

J. J. Greffet and C. Henkel, Contemp. Phys. 48, 183 (2007). Google Scholar

• [12]

C. Henkel, Nanometer-Scale Electromagnetic Field Fluctuations, in: Handbook of Theoretical and Computational Nanotechnology, vol. 7, chap. 25 (Eds. M. Rieth and W. Schommers), American Scientific Publishers, Stevenson Ranch, CA 2006, p. 463. Google Scholar

• [13]

C. H. Henry and R. F. Kazarinov, Rev. Mod. Phys. 68, 801 (1996). Google Scholar

• [14]

A. Einstein, Ann. Phys. (Leipzig) 322, 549 (1905), former citation: Ann. d. Physik, Vierte Folge 17, 549. Google Scholar

• [15]

F. Garca-Moliner and F. Flores, Introduction to the Theory of Solid Surfaces, Cambridge University Press, Cambridge 1979. Google Scholar

• [16]

D. Bedeaux and J. Vlieger, Optical Properties of Surfaces, World Scientific, Singapore 2004. Google Scholar

• [17]

A. Liebsch, Phys. Rev. Lett. 71, 145 (1993), comment: P. J. Feibelman, Phys. Rev. Lett. 72, 788 and reply 789 (1994). Google Scholar

• [18]

N. Asger Mortensen, S. Raza, M. Wubs, T. Søndergaard, and S. I. Bozhevolnyi, Nat. Commun. 5, 3809 (2014). Google Scholar

• [19]

S. Y. Buhmann, D. T. Butcher, and S. Scheel, New J. Phys. 14, 083034 (2012). Google Scholar

• [20]

L. P. Pitaevskii, Las. Phys. 19, 632 (2009). Google Scholar

• [21]

Y. S. Barash and V. L. Ginzburg, Sov. Phys. Usp. 18, 305 (1975). Google Scholar

• [22]

S. Scheel and D. G. Welsch, Phys. Rev. Lett. 96, 073601 (2006). Google Scholar

• [23]

H. Soo and M. Krüger, Europhys. Lett. 115, 41002 (2016). Google Scholar

• [24]

J. B. Xu, K. Lauger, R. Moller, K. Dransfeld, and I. H. Wilson, J. Appl. Phys. 76, 7209 (1994). Google Scholar

• [25]

E. Rousseau, A. Siria, G. Jourdan, S. Volz, F. Comin, et al., Nat. Photon. 3, 514 (2009). Google Scholar

• [26]

S. Shen, A. Narayanaswamy, and G. Chen, Nano Lett. 9, 2909 (2009). Google Scholar

• [27]

A. Babuty, K. Joulain, P. O. Chapuis, J. J. Greffet, and Y. De Wilde, Phys. Rev. Lett. 110, 146103 (2013). Google Scholar

• [28]

R. S. Ottens, V. Quetschke, S. Wise, A. A. Alemi, R. Lundock, et al., Phys. Rev. Lett. 107, 014301 (2011). Google Scholar

• [29]

T. Kralik, P. Hanzelka, M. Zobac, V. Musilova, T. Fort, et al., Phys. Rev. Lett. 109, 224302 (2012). Google Scholar

• [30]

R. St-Gelais, L. Zhu, S. Fan, and M. Lipson, Nat. Nanotech. 11, 515 (2016). Google Scholar

• [31]

L. Worbes, D. Hellmann, and A. Kittel, Phys. Rev. Lett. 110, 134302 (2013). Google Scholar

• [32]

B. Song, Y. Ganjeh, S. Sadat, D. Thompson, A. Fiorino, et al., Nat. Nanotech. 10, 253 (2015). Google Scholar

• [33]

R. Shayduk, H. Navirian, W. Leitenberger, J. Goldshteyn, I. Vrejoiu, et al., New J. Phys. 13, 093032 (2011). Google Scholar

• [34]

H. B. Callen and T. A. Welton, Phys. Rev. 83, 34 (1951). Google Scholar

• [35]

K. Usami, Y. Nambu, B. S. Shi, A. Tomita, and K. Nakamura, Phys. Rev. Lett. 92, 113601 (2004). Google Scholar

• [36]

O. Di Stefano, S. Savasta, and R. Girlanda, Phys. Rev. A 61, 02 3803 (2000). Google Scholar

• [37]

L. Knöll, S. Scheel, and D. G. Welsch, QED in Dispersing and Absorbing Dielectric Media, in: Coherence and Statistics of Photons and Atoms, chap. 1 (Ed. J. Perina), John Wiley & Sons, Inc., New York 2001, pp. 1–64. Google Scholar

• [38]

J. D. Jackson, Classical Electrodynamics, 2nd ed., Wiley & Sons, New York 1975. Google Scholar

• [39]

J. M. Wylie and J. E. Sipe, Phys. Rev. A 30, 1185 (1984). Google Scholar

• [40]

J. E. Sipe, J. Opt. Soc. Am. B 4, 481 (1987). Google Scholar

• [41]

M. S. Tomaš, Phys. Rev. A 51, 2545 (1995). Google Scholar

• [42]

H. Chew, J. Chem. Phys. 87, 1355 (1987). Google Scholar

• [43]

V. A. Parsegian, Van der Waals Forces – A Handbook for Biologists, Chemists, Engineers, and Physicists, Cambridge University Press, New York 2006. Google Scholar

• [44]

W. C. Chew, Waves and Fields in Inhomogeneous Media Electromagnetic Waves, Wiley-IEEE Press, Piscataway, NJ 1999. Google Scholar

• [45]

M. Levin and S. Rytov, Sov. Phys. JETP 38, 688 (1974) [ZhETF 65, 1382 (1974)]. Google Scholar

• [46]

M. T. H. Reid, A. W. Rodriguez, and S. G. Johnson, Proc. IEEE 101, 531 (2013). Google Scholar

• [47]

E. M. Purcell and C. R. Pennypacker, Astrophys. J. 186, 705 (1973). Google Scholar

• [48]

B. T. Draine, Astrophys. J. 333, 848 (1988). Google Scholar

• [49]

A. Lakhtakia, Opt. Commun. 79, 1 (1990). Google Scholar

• [50]

A. G. Polimeridis, M. T. H. Reid, W. Jin, S. G. Johnson, J. K. White, et al., Phys. Rev. B 92, 134202 (2015). Google Scholar

• [51]

S. Edalatpour and M. Francoeur, Phys. Rev. B 94, 045406 (2016). Google Scholar

• [52]

A. Rodriguez, M. Ibanescu, D. Iannuzzi, F. Capasso, J. D. Joannopoulos, et al., Phys. Rev. Lett. 99, 080401 (2007). Google Scholar

• [53]

E. M. Lifshitz, Soviet Phys. JETP 2, 73 (1956) [ZhETF 29, 94 (1955)]. Google Scholar

• [54]

S. Y. Buhmann, Dispersion Forces I – Macroscopic Quantum Electrodynamics and Ground-State Casimir, Casimir–Polder and van der Waals Forces (Springer Tracts in Modern Physics vol. 247), Springer, Heidelberg 2012. Google Scholar

• [55]

W. Eckhardt, Phys. Rev. A 29, 1991 (1984). Google Scholar

• [56]

S. Scheel, L. Knöll, and D. G. Welsch, Phys. Rev. A 58, 700 (1998). Google Scholar

• [57]

G. S. Agarwal, Phys. Rev. A 11, 230 (1975). Google Scholar

• [58]

G. Barton, J. Phys. Condens. Matt. 27, 214005 (2015). Google Scholar

• [59]

G. Barton, J. Stat. Phys., (2016). doi: .

• [60]

J. J. Loomis and H. J. Maris, Phys. Rev. B 50, 18517 (1994). Google Scholar

• [61]

• [62]

T. Setälä, M. Kaivola, and A. T. Friberg, Phys. Rev. Lett. 88, 123902 (2002).Google Scholar

• [63]

J. Ellis, A. Dogariu, S. Ponomarenko, and E. Wolf, Opt. Commun. 248, 333 (2005).Google Scholar

• [64]

G. S. Agarwal, G. Gbur, and E. Wolf, Opt. Lett. 29, 459 (2004). Google Scholar

• [65]

J. R. Zurita-Sánchez, J. Opt. Soc. Am. A 33, 118 (2016). Google Scholar

• [66]

X. L. Li, G. W. Ford, and R. F. O’Connell, Phys. Rev. E 48, 1547 (1993), comment: I. R. Senitzky, Phys. Rev. E 51, 5166 and reply 5169 (1995). Google Scholar

• [67]

D. Polder and M. van Hove, Phys. Rev. B 4, 3303 (1971). Google Scholar

• [68]

D. A. R. Dalvit, P. W. Milonni, D. Roberts, and F. da Rosa, Eds., Casimir Physics (Lecture Notes in Physics vol. 834), Springer, Berlin 2011. Google Scholar

• [69]

W. Simpson and U. Leonhardt, Eds., Forces of the Quantum Vacuum – An Introduction to Casimir Physics, World Scientific, Singapore 2015. Google Scholar

• [70]

F. Chen, G. L. Klimchitskaya, V. M. Mostepanenko, and U. Mohideen, Opt. Express 15, 4823 (2007). Google Scholar

• [71]

F. Chen, G. L. Klimchitskaya, V. M. Mostepanenko, and U. Mohideen, Phys. Rev. B 76, 035338 (2007). Google Scholar

• [72]

V. B. Svetovoy, Phys. Rev. Lett. 101, 163603 (2008), erratum: Phys. Rev. Lett. 102, 219903 (2009). Google Scholar

• [73]

G. Bimonte, D. Lopez, and R. S. Decca, Phys. Rev. B 93, 184434 (2016). Google Scholar

• [74]

R. S. Decca, D. López, E. Fischbach, G. L. Klimchitskaya, D. E. Krause, et al., Phys. Rev. D 75, 077101 (2007). Google Scholar

• [75]

R. S. Decca, D. López, E. Fischbach, G. L. Klimchitskaya, D. E. Krause, et al., Eur. Phys. J. C 51, 963 (2007). Google Scholar

• [76]

Y. J. Chen, W. K. Tham, D. E. Krause, D. López, E. Fischbach, et al., Phys. Rev. Lett. 116, 221102 (2016). Google Scholar

• [77]

K. Kim, B. Song, V. Fernández-Hurtado, W. Lee, W. Jeong, et al., Nature 528, 387 (2015). Google Scholar

• [78]

K. Kloppstech, N. Könne, S. A. Biehs, A. W. Rodriguez, L. Worbes, et al., arXiv:1510.06311 (2015). Google Scholar

• [79]

S. Wang, J. Ng, M. Xiao, and C. T. Chan, Sci. Adv. 2, e1501485 (2016). Google Scholar

• [80]

J. S. Wang and J. Peng, A Microscopic Theory for Ultra-Near-Field Radiation, arXiv:1607.02840 (2016). Google Scholar

• [81]

G. Domingues, S. Volz, K. Joulain, and J. J. Greffet, Phys. Rev. Lett. 94, 085901 (2005). Google Scholar

• [82]

V. Chiloyan, J. Garg, K. Esfarjani, and G. Chen, Nat. Commun. 6, 6755 (2015). Google Scholar

• [83]

N. Mingo, Green’s Function Methods for Phonon Transport Through Nano-Contacts, in: Thermal Nanosystems and NanoMaterials, chap. 3 (Ed. S. Volz), Topics in Applied Physics vol. 118, Springer, Berlin 2009, pp. 63–94. Google Scholar

Accepted: 2016-11-10

Published Online: 2016-12-17

Published in Print: 2017-02-01

Funding Source: Deutsche Forschungsgemeinschaft

Award identifier / Grant number: Schm-1049/7-1

Funding: Deutsche Forschungsgemeinschaft through the DIP program (Grant Number: Schm-1049/7-1).

Citation Information: Zeitschrift für Naturforschung A, ISSN (Online) 1865-7109, ISSN (Print) 0932-0784,

Export Citation

©2017 Walter de Gruyter GmbH, Berlin/Boston.