Hyperbolic metamaterials (HMM) are of great interest due to their ability to break the diffraction limit for imaging and enhance near-field radiative heat transfer. Here we demonstrate that an annular, transparent HMM enables selective heating of a sub-wavelength plasmonic nanowire by controlling the angular mode number of a plasmonic resonance. A nanowire emitter, surrounded by an HMM, appears dark to incoming radiation from an adjacent nanowire emitter unless the second emitter is surrounded by an identical lens such that the wavelength and angular mode of the plasmonic resonance match. Our result can find applications in radiative thermal management.
© 2015 Optical Society of America
Engineering thermal radiation is of importance for a number of technologies, including infrared imaging, energy conversion, thermal insulation, thermal signature control, and thermal management . Recent works have demonstrated that far-field spectral and angular characteristics of thermal radiation can be controlled using photonic crystals [2–4] and metamaterials [5–8]. These structures can also enable near-field resonant surface modes to propagate into the far-field using gratings  and antennas  to out-couple surface modes. In the near field, radiative heat transfer can be greatly enhanced due to the presence of evanescent waves [11, 12]. These enhancements have recently been demonstrated experimentally [13–16]. Thermal radiation into the far-field can also be enhanced in a thermal extraction scheme in which an impedance-matched extraction device allows the propagation of internally reflected modes .
Recently, hyperbolic metamaterials (HMMs) have been under intense investigation for their potential to control thermal radiation. HMMs possess dielectric constants of opposite sign along different axes and hence allow the propagation of high momentum modes within the HMM due to the hyperbolic dispersion . HMMs can be fabricated in practice as a multilayer stack with alternating materials of opposite sign of dielectric constant. HMMs were originally of interest for their potential to project images with resolution below the diffraction limit into the far-field, as proposed theoretically [19, 20] and later demonstrated experimentally . For thermal radiation, HMMs have been studied for their potential to enhance near-field heat transfer [22–25] as well as control the spectral and angular distribution of far-field radiation [26–29].
Here, we examine how HMMs modify the far-field thermal emission spectrum of nanostructures. We find that a lossy plasmonic nanowire surrounded by a transparent, annular HMM lens yields thermal emission that primarily occurs only at a specific wavelength and angular mode number and greatly exceeds that of the nanowire alone. This angular mode resonance enables highly selective radiative heating because only nanowires that are surrounded by identical HMM lenses can exchange radiation.
2. Absorption of nanowire and annular HMM lens
We begin by considering a lossy nanowire core of radius a that is in optical contact with a lens medium in vacuum, as shown in the inset of Fig. 1(a). The system is assumed to be infinite in the z direction with the polarization such that E ⊥ z. The magnetic field from an incident plane wave can be expressed as30] and am and bm are the coefficients of the Hankel function of the scattered field in outermost vacuum (ρ > b) and the Bessel function of the transmitted field of core (ρ < a), respectively. and are the coefficients for the Bessel and Hankel function of the field in the jth layer, and kj denotes the wave vector of each jth layer up to the core. k0 denotes the wave vector in vacuum. The coefficients am, bm, and can be solved by matching boundary conditions of tangential fields at the boundary of each layer using the Transfer Matrix Method (TMM) in cylindrical coordinates [30–32]. In this method, the continuity of the tangential components of the E and H fields at the boundary of each layer derived from Eq. 1 can be written in matrix form and the transfer matrix of the whole system can be calculated by multiplying the transfer matrices of individual interfaces.
The absorption efficiency Qabs can then be expressed asEq. (1) according to Mie theory [33,34]. By Kirchoff’s law, the absorptivity equals the emissivity for each direction and wavelength , and hence Qabs can be interpreted as the emissivity. Note that the emissivity can exceed unity for subwavelength objects because the absorption cross-section can be larger than the geometric cross-section .
The HMM lens consists of an alternating layered structure of dielectric and metal leading to anisotropic permittivity along the radial and tangential direction. These anisotropic dielectric constants can be expressed using effective medium theory (EMT) according to (ερ,εθ) = (εmεd/((1 − f)εm + fεd), fεm + (1 − f)εd) where f is the volume fraction occupied by the metal and εm, εd are the respective metal and dielectric permittivities . We assume the lens to be lossless and define Qabs with respect to the core radius a. Neglecting loss in the lens means that the lens cannot exchange thermal radiation with the core, and thus its contribution to heat transfer can be neglected.
First, we consider the emissive properties of only the nanowire core of radius a in vacuum without any lens. Figure 1(a) shows the computed emissivity Qabs for the nanowire core with a permittivity of −1.05+0.01i and a = 0.1λ where λ is the wavelength of the incident field. We choose the core permittivity close to the ideal plasmonic resonance condition to demonstrate our result but other negative real permittivity values can be chosen with similar results. We assume a typical wavelength λ = 10 μm, corresponding to the maximum of the blackbody spectrum around 290 K, giving a = 1 μm and yielding an emissivity of 0.5. The maximum emissivities for the nanowire core decrease with increasing size parameter as the absorption efficiency scales as 1/a. Note that plasmonic resonances do occur at specific sizes for a given permittivity for a nanowire core , but tuning the angular mode number of the resonance requires changing the permittivity of the nanowire.
Now, consider the nanowire surrounded by a transparent material called “lens” as shown in inset of Fig. 1(a). The transparent lens is assumed to be lossless such that it cannot exchange radiation with the core. The total thickness of the core and lens b considered in Fig. 1(a) ranges from 1 − 7 μm with corresponding size parameters k0b shown. We assume a vacuum gap of width λ/200 (50 nm for λ = 10 μm) exists to prevent heat conduction, although this assumption does not affect our conclusions. The addition of this lens with a lossless metal of permittivity ε = −1.05 or a dielectric of permittivity ε = 10 results in a lower emissivity Qabs than the bare core (Core-Vacuum) case. This reduction in emissivity can be attributed to the impedance mismatch between the lens and vacuum that reflects some modes before they reach the absorptive core.
Next, consider the nanowire surrounded by a transparent HMM lens. We compute am in this case using either EMT or considering each individual layer of the HMM with a transfer matrix. For the EMT-HMM case, we scale m in Eq. (1) of the HMM layer  to . For the layer by layer case, the thickness of each metal-dielectric bi-layer is chosen to be λ/400 (25 nm for λ = 10 μm). We examine both the metal-dielectric (TMM-md) and dielectric-metal (TMM-dm) structures such that the first layer adjacent to the core is a metal or a dielectric, respectively. For the EMT-HMM case, we take the optical constants to be (ερ,εθ) = (10, −0.025) according to EMT. For the TMM calculation, we take (εm,εd) = (−5.1,3.4) with f = 0.4, giving the same values of (εm,εd) = (10,−0.025) as EMT.
This calculation is plotted in Fig. 1(a). In contrast to decrease of emissivity Qabs with the metal and dielectric lens, the emissivity Qabs with the HMM lens exhibits strong peaks as the size parameter increases for both the EMT and TMM calculations. The emissivity Qabs peaks in the TMM-md and TMM-dm cases are in close proximity to the right and left of the EMT-HMM peaks, respectively, and converge to the EMT result as the layer thickness decreases. Thus, by placing a HMM lens of the right size at one of these peaks around the core, the emissivity can be increased by about three times compared to the same bare nanowire core. Larger enhancements of 4-5 times relative to the bare core can be achieved at larger core sizes for the same loss of the core. Enhancements greater than 50 times that of a larger bare core can be achieved if the loss of the core is optimized but the required small loss is not realistic for any available plasmonic materials and thus is not considered further.
3. Angular-mode specific resonances
To understand the origin of these peaks, we examine the decomposition of absorptivity from the EMT-HMM case in Fig. 1(a) into partial absorptivity for modes m = 1 to m = 6 as shown in Fig. 1(b). The m = 1 and m = 2 cases do not have resonant peaks for the given size range but modes m = 3 to m = 6 each have a specific resonance at different size parameters k0b. These resonant size parameters correspond to the same peak positions in Fig. 1(a) and achieve emissivity close to the well-known single channel limit . At a given size parameter, most of the total absorption cross-section is due to a single resonant angular mode.
Further, assuming a Drude model for optical properties, this resonance yields by far the largest emissivity over a considerable range of wavelength. We examine the wavelength dependence of the enhancement in thermal emission that can be achieved using the HMM lens using a Drude model given by , where ω is the frequency and ωp is the plasmon frequency. For the core, γ = 0.0035ωp and λp = 2πc/ωp = 7 μm. The metal in the HMM is assumed to have a Drude dispersion that is lossless (γ = 0) and λp = 4.05 μm. These parameters yield the same permittivities as used in Fig. 1 at a wavelength λ = 10 μm as shown in the inset in Fig. 2(a). The partial emissivity Qabs,m versus wavelength for size parameter k0b = 1.8, at the resonance for m = 4, is plotted in Fig. 2(a). At a particular wavelength, the emissivity is nearly entirely due to a single angular mode; for example, the resonant peak at 10 μm is nearly completely due to m = 4 mode, with a small additional contribution from m = 3 but not from m = 5.
We now compare the overlap of these resonances for identical nanowires surrounded by HMM lenses of different size parameters by multiplying the partial emissivity Qabs,m from Fig. 2(a) for two different size parameters, k0b = 2.6 and k0b = 1.8, for modes m = 3,4,5. As shown in Fig. 2(b), there is negligible overlap between the partial emissivity of the two cases over the full range of the blackbody spectrum at 290 K. Although not plotted, negligible overlap also occurs for higher order modes m > 6. Physically, this small overlap indicates that little of the emitted radiation from a core lens system of size k0b = 1.8 will be absorbed by a core lens system of a size parameter k0b = 2.6 and vice versa.
We thus arrive at the principal result of our study. Nanowires surrounded by HMM lenses interact with radiation primarily at a particular wavelength and angular mode with absorptivity that can reach the single channel limit. Therefore, radiation emitted by a nanowire with a certain HMM lens can only minimally exchange radiative heat with other identical nanowires surrounded by lenses of different size parameters. Unlike other selective heating schemes based on plasmonic resonances [38–42], the selective resonance identified here is based both on wavelength and angular mode number, enabling high selectivity. This effect is harder to realize with the plasmonic resonances of the bare nanowire alone because achieving similar mode selectivity close to the single channel limit requires tuning both size parameter or material permittivity of the nanowires, while all material properties remain fixed with our core-lens system.
4. Origin of angular selectivity
We investigate the origin of the angular selectivity by comparing the observed resonance with previous applications of curvilinear HMMs as hyperlenses [19, 21]. Hyperlenses are used to convert high angular momentum, evanescent modes to propagating modes using conservation of angular momentum as the mode propagates radially outward. The mode becomes propagating inside the HMM lens when size parameter k0b ≥ m. However, k0b is 1.8 for the m = 4 mode on resonance in Fig. 1(a), indicating that the excitation in vacuum is actually evanescent. This observation indicates that the HMM lens here is modifying the plasmon resonance of the core similar to the mechanism of enhancement in Ng et al.  rather than converting evanescent and propagating waves. We confirm that the resonance is plasmonic in nature by noticing that little absorption is observed for the polarization for which E||z.
The origin of the selectivity is also not solely due to the hyperbolic dispersion. HMMs are typically of interest because the hyperbolic dispersion occurs over a broad spectral range, as is the case here. However, Fig. 2(a) and the inset shows that the mode selectivity only occurs around the εθ close to zero region of the HMM dispersion, making the selectivity narrowband. The angular selectivity thus requires the anisotropic properties of the HMM but also the epsilon-near-zero (ENZ) region of the dispersion along the θ direction.
Next, we examine the angular mode selectivity using the well-known single channel limit for absorption and scattering. Physically, the single-channel limit is achieved when radiative damping and absorptive loss both contribute equally to the absorption efficiency of the mode [44–46]. Mathematically, from Eq. (2) the maximum partial absorption cross-section occurs  when Re(am) = 1/2 and Im(am) = 0, yielding Qabs,m = 1/(2k0a). For example, when a = 0.1λ, the limit for partial emissivity is Qabs,m ≈ 0.796 as indicated in Fig. 1(b). Figure 2(c) plots the real and imaginary part of the coefficient am in Eq. (1) for mode m = 4 demonstrating that this mode meets the conditions required to reach the single-channel limit for k0b = 1.8. Likewise, modes m = 5 and m = 6 reach the single-channel limit in Fig. 1(b) and satisfy the same conditions for am at their respective resonant size parameters. However, due to the wavelength-dependence of permittivity, the requirements of the single-channel limit for a fixed size parameter can be met for a single angular mode but are unlikely to be satisfied for other angular modes, as in Fig. 2(a). This sensitivity of the angular resonance to the conditions of the single-channel limit contributes to the mode selectivity.
We further investigate this modal selectivity by examining the resonant mode profiles using the TMM calculation. We reconstruct the field profile of |Hz| in wave 2D for each mode |m| with incident plane wave direction defined in Fig. 1(a). Although am is symmetric for positive and negative m, we must account for the phase factors exp(imϕ) to accurately plot the spatial profile. Figures 3(a)–3(d) show the 2D plots of |Hz| corresponding to three different size parameters in Fig. 1(a) for the TMM-md case. The field magnitude |Hz| at resonant size parameters k0b ≈ 1.1 (|m| = 3) and k0b ≈ 1.9 (|m| = 4) are plotted in Figs. 3(a) and 3(b), respectively. In Figs. 3(c) and 3(d), we plot |Hz| for modes |m| = 3 and |m| = 4, respectively, for an intermediate size parameter k0 ≈ 1.62 that is off resonance. We observe from Figs. 3(a) and 3(b) that the lobe patterns at the resonant mode number are highly-confined within the HMM lens. In contrast, in Figs. 3(c) and 3(d) the modes are not confined. Additionally, the fields magnitudes |Hz| in Figs. 3(a) and 3(b) are higher than in Figs. 3(c) and 3(d) by a factor between 3 to 4. The strong, localized field intensities in Figs. 3(a) and 3(b) highlights the modal selectivity of the resonances at specific size parameters.
We can gain further insight into the origin of the thermal emission spectrum by examining the bulk behavior of an equivalent planar structure. We use the planar Transfer Matrix Method (pTMM) to simulate the equivalent bulk HMM structure on a semi-infinite metallic substrate of the same permittivity of −1.05+0.01i as the core in Fig. 1. The HMM has the same bi-layer thickness of λ/400 and material arrangement, including the air-gap, as the TMM-md case of the HMM lens calculation in Fig. 1. We relate the wave vector component parallel to the vacuum-HMM interface k║ in the planar case to m in the cylindrical case by approximating the mode to lie within the HMM  such that k║ = m/reff where reff = (a + b)/2. The penetration of the modes through the HMM to the absorbing layer can be observed by the non-zero imaginary part of the Fresnel reflection coefficient Im(Rp) which describes the absorption of the incident evanescent field .
We plot log[Im(Rp)] obtained from pTMM against the normalized parallel wave vector k║/k0 and number of HMM bi-layers N in Fig. 4(a). As N increases, the position of maximum Im(Rp) decreases from the metal-vacuum surface plasmon condition of k║/k0 ≈ 4.6 to around k║/k0 ≈ 3, decreasing the high parallel momentum for plasmonic resonance when the HMM is present. As m is a measure of the angular momentum, the above relationship k║ = m/reff indicates that the angular momentum for the mode is reduced, for a fixed effective radius reff, when k║ is decreased. We also plot log[Im(Rp)] versus the converted effective m and size parameter k0b in Fig. 4(b) and overlay the positions of the resonances of the cylindrical case in Fig. 1 onto Fig. 4(b). The resonant peaks in the cylindrical case closely follow the prediction of the planar case, allowing us to conclude that both resonances are of the same nature.
From this planar analysis, we can understand the relationship between the size parameter and mode number of the resonances in Fig. 1(b). After approximately 50 bi-layers, the parallel momentum required to excite the resonance becomes nearly constant as in Fig. 4(a). From the relation k║ = 2m/(a + b), if k║ is constant as b increases m must also increase, leading to the nearly linear increase of the mode number with size parameter as in Fig. 1(b).
We now examine the optical properties of the HMM lens and core that will allow the selectivity by studying how the partial emissivity of a mode depends on the permittivity of the HMM lens. Figure 4(c) plots the partial emissivity for the m = 4 mode (k0b ≈ 1.8 for EMT-HMM case in Fig. 1(a)) as ερ and εθ varies. From Fig. 4(c), the largest enhancement occurs in the region of ερ > 5 and a negative but close to zero value of εθ. The enhancement for these permittivity values can be explained by the dispersion relation in the HMM , , and noting that small and negative εθ, with kθ/k0 ≈ 3 and ερ = 10 for example, causes kρ to be very small and imaginary and allows the field to extend to the inner absorbing core. The sensitivity of the mode selective plasmonic resonances to the HMM parameters is unlike typical broadband enhancement effects of HMMs [22, 23].
5. Effects of lossy HMMs lens
Finally, we consider the effect of loss in the HMM lens. Physically, loss causes the lens to also play a role in radiative transfer. Since the temperature of the lens is not fixed, HMM lens will equilibrate to a temperature close to that of the heated core, allowing us to consider the core-lens structure as a single object for the purposes of analyzing radiative emission. We incorporate loss by modifying the Drude dispersion of the metal to have γ = 0.001ωp so that (εm,εd) = (−5.1 + .015i,3.4) at 10 μm which is close to the lowest loss with negative real permittivity in the mid-infrared range of materials such as 4H-SiC . The partial emissivity Qabs,m is now defined with respect to the size of the whole structure b. As shown in Fig. 5, adding loss decreases the peak absorptivity around 10 μm for m = 4 compared to the lossless case of Fig. 2(a). Also, with loss the m = 4 mode is no longer as dominant a resonance compared to adjacent m = 3,5 modes in the wavelength range shown. We conclude that loss reduces the angular mode and wavelength selectivity for selective heating and thus that fully exploiting the thermal HMM lens requires low-loss plasmonic materials in the infrared. Recently, hexagonal Boron Nitride has been demonstrated as a low-loss material in the mid-infrared range with a hyperbolic dispersion [49, 50], potentially allowing the layered HMM lens to be replaced with a single material.
In summary, we theoretically demonstrated a new approach to selective radiative heating based on tuning angular mode resonances with HMM lenses. This approach enables selectivity for thermal radiative exchange due to the requirement that both wavelength and angular mode number of the emitter and absorber match. Our result could have applications in radiative thermal management.
The authors thank J. D. Caldwell for useful discussions. This work is part of the ‘Light-Material Interactions in Energy Conversion’ Energy Frontier Research Center funded by the U.S. Department of Energy, Office of Science, Office of Basic Energy Sciences under Award Number DE-SC0001293. D.D. gratefully acknowledges fellowship support from the Agency for Science, Technology and Research (Singapore).
1. B. Liu, J. Shi, K. Liew, and S. Shen, “Near-field radiative heat transfer for Si based metamaterials,” Opt. Commun. 314, 57–65 (2014). [CrossRef]
2. J.-H. Lee, Y.-S. Kim, K. Constant, and K.-M. Ho, “Woodpile metallic photonic crystals fabricated by using soft lithography for tailored thermal emission,” Adv. Mater. 19, 791–794 (2007). [CrossRef]
4. Y. X. Yeng, M. Ghebrebrhan, P. Bermel, W. R. Chan, J. D. Joannopoulos, M. Soljačić, and I. Celanovic, “Enabling high-temperature nanophotonics for energy applications,” Proc. Natl. Acad. Sci. USA 109, 2280–2285 (2012). [CrossRef] [PubMed]
5. X. Liu, T. Tyler, T. Starr, A. F. Starr, N. M. Jokerst, and W. J. Padilla, “Taming the blackbody with infrared metamaterials as selective thermal emitters,” Phys. Rev. Lett. 107, 045901 (2011). [CrossRef] [PubMed]
6. P. Bermel, M. Ghebrebrhan, M. Harradon, Y. X. Yeng, I. Celanovic, J. D. Joannopoulos, and M. Soljačić, “Tailoring photonic metamaterial resonances for thermal radiation,” Nanoscale Res. Lett. . 6, 549 (2011). [CrossRef] [PubMed]
7. N. Mattiucci, G. D’Aguanno, A. Alù, C. Argyropoulos, J. V. Foreman, and M. J. Bloemer, “Taming the thermal emissivity of metals: A metamaterial approach,” Appl. Phys. Lett. 100, 201109 (2012). [CrossRef]
8. H. Wang and L. Wang, “Perfect selective metamaterial solar absorbers,” Opt. Express 21, A1078–A1093 (2013). [CrossRef]
10. J. A. Schuller, T. Taubner, and M. L. Brongersma, “Optical antenna thermal emitters,” Nature Photon. 3, 658–661 (2009). [CrossRef]
11. E. G. Cravalho, C. L. Tien, and R. P. Caren, “Effect of small spacings on radiative transfer between two dielectrics,” Journal of Heat Transfer 89, 351–358 (1967). [CrossRef]
12. D. Polder and M. Van Hove, “Theory of radiative heat transfer between closely spaced bodies,” Phys. Rev. B 4, 3303–3314 (1971). [CrossRef]
13. A. Kittel, W. Müller-Hirsch, J. Parisi, S.-A. Biehs, D. Reddig, and M. Holthaus, “Near-field heat transfer in a scanning thermal microscope,” Phys. Rev. Lett. 95, 224301 (2005). [CrossRef] [PubMed]
15. E. Rousseau, A. Siria, G. Jourdan, S. Volz, F. Comin, J. Chevrier, and J.-J. Greffet, “Radiative heat transfer at the nanoscale,” Nature Photon. 3, 514–517 (2009). [CrossRef]
16. R. S. Ottens, V. Quetschke, S. Wise, A. A. Alemi, R. Lundock, G. Mueller, D. H. Reitze, D. B. Tanner, and B. F. Whiting, “Near-field radiative heat transfer between macroscopic planar surfaces,” Phys. Rev. Lett. 107, 014301 (2011). [CrossRef] [PubMed]
20. A. Salandrino and N. Engheta, “Far-field subdiffraction optical microscopy using metamaterial crystals: Theory and simulations,” Phys. Rev. B 74, 075103 (2006). [CrossRef]
24. B. Liu and S. Shen, “Broadband near-field radiative thermal emitter/absorber based on hyperbolic metamaterials: Direct numerical simulation by the wiener chaos expansion method,” Phys. Rev. B 87, 115403 (2013). [CrossRef]
25. Y. Guo, S. Molesky, H. Hu, C. L. Cortes, and Z. Jacob, “Thermal excitation of plasmons for near-field thermophotovoltaics,” Appl. Phys. Lett. 105, 073903 (2014). [CrossRef]
26. E. E. Narimanov, H. Li, Y. A. Barnakov, T. U. Tumkur, and M. A. Noginov, “Darker than black: radiation-absorbing metamaterial,” arXiv:1109.5469 [cond-mat, physics:physics] (2011).
27. S. Molesky, C. J. Dewalt, and Z. Jacob, “High temperature epsilon-near-zero and epsilon-near-pole metamaterial emitters for thermophotovoltaics,” Opt. Express 21, A96–A110 (2013). [CrossRef] [PubMed]
28. D. Ji, H. Song, X. Zeng, H. Hu, K. Liu, N. Zhang, and Q. Gan, “Broadband absorption engineering of hyperbolic metafilm patterns,” Sci. Rep.4 (2014). [CrossRef]
29. I. Nefedov and L. Melnikov, “Super-planckian far-zone thermal emission from asymmetric hyperbolic metamaterials,” arXiv:1402.3507 [physics] (2014).
30. J. Scheuer, W. M. J. Green, G. A. DeRose, and A. Yariv, “InGaAsP annular bragg lasers: theory, applications, and modal properties,” IEEE Journal of Selected Topics in Quantum Electronics 11, 476–484 (2005). [CrossRef]
31. W. C. Chew, Waves and Fields in Inhomogenous Media (John Wiley & Sons, 1999). [CrossRef]
32. V. V. Nikolaev, G. S. Sokolovskii, and M. A. Kaliteevskii, “Bragg reflectors for cylindrical waves,” Semiconductors 33, 147–152 (1999). [CrossRef]
33. H. C. v. d. Hulst, Light Scattering by Small Particles (Dover Publications, 1981).
34. C. F. Bohren and D. R. Huffman, Absorption and Scattering of Light by Small Particles (Wiley-VCH, 1998). [CrossRef]
35. M. F. Modest, Radiative Heat Transfer (Academic, 2003).
36. B. S. Luk’yanchuk and V. Ternovsky, “Light scattering by a thin wire with a surface-plasmon resonance: Bifurcations of the poynting vector field,” Phys. Rev. B 73, 235432 (2006). [CrossRef]
37. Y. Ni, L. Gao, and C.-W. Qiu, “Achieving invisibility of homogeneous cylindrically anisotropic cylinders,” Plasmonics 5, 251–258 (2010). [CrossRef]
39. A. J. Schmidt, J. D. Alper, M. Chiesa, G. Chen, S. K. Das, and K. Hamad-Schifferli, “Probing the gold nanorodligand-solvent interface by plasmonic absorption and thermal decay,” J. Phys. Chem. C 112, 13320–13323 (2008). [CrossRef]
41. J. Huang, W. Wang, C. J. Murphy, and D. G. Cahill, “Resonant secondary light emission from plasmonic Au nanostructures at high electron temperatures created by pulsed-laser excitation,” Proc. Natl. Acad. Sci. USA 111, 906–911 (2014). [CrossRef] [PubMed]
44. R. E. Hamam, A. Karalis, J. D. Joannopoulos, and M. Soljačić, “Coupled-mode theory for general free-space resonant scattering of waves,” Phys. Rev. A 75, 053801 (2007). [CrossRef]
47. O. Kidwai, S. V. Zhukovsky, and J. E. Sipe, “Effective-medium approach to planar multilayer hyperbolic meta-materials: Strengths and limitations,” Phys. Rev. A 85, 053842 (2012). [CrossRef]
48. Y. Chen, Y. Francescato, J. D. Caldwell, V. Giannini, T. W. W. Ma, O. J. Glembocki, F. J. Bezares, T. Taubner, R. Kasica, M. Hong, and S. A. Maier, “Spectral tuning of localized surface phonon polariton resonators for low-loss mid-IR applications,” ACS Photonics 1, 718–724 (2014). [CrossRef]
49. S. Dai, Z. Fei, Q. Ma, A. S. Rodin, M. Wagner, A. S. McLeod, M. K. Liu, W. Gannett, W. Regan, K. Watanabe, T. Taniguchi, M. Thiemens, G. Dominguez, A. H. C. Neto, A. Zettl, F. Keilmann, P. Jarillo-Herrero, M. M. Fogler, and D. N. Basov, “Tunable phonon polaritons in atomically thin van der waals crystals of boron nitride,” Science 343, 1125–1129 (2014). [CrossRef] [PubMed]
50. J. D. Caldwell, A. V. Kretinin, Y. Chen, V. Giannini, M. M. Fogler, Y. Francescato, C. T. Ellis, J. G. Tischler, C. R. Woods, A. J. Giles, M. Hong, K. Watanabe, T. Taniguchi, S. A. Maier, and K. S. Novoselov, “Sub-diffractional volume-confined polaritons in the natural hyperbolic material hexagonal boron nitride,” Nat. Commun.5 (2014). [CrossRef] [PubMed]