Efficient homogenization procedure for the ... - OSA Publishing

Efficient homogenization procedure for the calculation of optical properties of 3D nanostructured composites W. Luis Mochán,1,4 Guillermo P. Ortiz,2,5 and Bernardo S. Mendoza3,6 1 Instituto

de Ciencias F´ısicas, Universidad Nacional Autónoma de México, Apdo. Postal 48-3, 62251 Cuernavaca, Morelos, México 2 Departamento de F´ısica, Facultad de Ciencias Exactas, Naturales y Agrimensura, Universidad Nacional del Nordeste - Instituto de Modelado e Innovación Tecnológica, CONICET-UNNE, Corrientes, Argentina 3 Department of Photonics, Centro de Investigaciones en Optica, Le´ on, Guanajuato, México 4 [email protected] 5 [email protected] 6 [email protected]

Abstract: We present a very efficient recursive method to calculate the effective optical response of metamaterials made up of arbitrarily shaped inclusions arranged in periodic 3D arrays. We apply it to dielectric particles embedded in a metal matrix with a lattice constant much smaller than the wavelength of the incident field, so that we may neglect retardation and factor the geometrical properties from the properties of the materials. If the conducting phase is continuous the low frequency behavior is metallic, and if the conducting paths are thin, the high frequency behavior is dielectric. Thus, extraordinary-transparency bands may develop at intermediate frequencies, whose properties may be tuned by geometrical manipulation. © 2010 Optical Society of America OCIS codes: (350.4238) Nanophotonics and photonic crystals; (160.3918) Metamaterials; (160.4236) Nanomaterials; (260.2065) Effective medium theory.

References and links 1. L. Novotny and B. Hecht, Principles of Nano-Optics (Cambridge University Press, New York, 2006). 2. J. Weiner, “The physics of light transmission through subwavelength apertures and aperture arrays,” Rep. Prog. Phys. 72, 0644011 (2009). 3. L. Chen and G. P. Wang, “Pyramid-shaped hyperlenses for three-dimensional subdiffraction optical imaging,” Opt. Express 17, 3903 (2009). 4. Z. Liu, H. Lee, Y. Xiong, C. Sun, and X. Zhang, “Far-field optical hyperlens magnifying sub-diffraction-limited objects,” Science 315, 1686 (2007). 5. A. Salandrino and N. Engheta, “Far-field subdiffraction optical microscopy using metamaterial crystal: theory and simulations,” Phys. Rev. B 74, 075103 (2006). 6. W. Cai, U. K. Chettiar, A. V. Kildishev, and V. M. Shalaev, “Nonmagnetic cloak with minimized scattering,” Appl. Phys. Lett. 91, 1111051 (2007). 7. L. Novotny, “Effective wavelength scaling for optical antennas,” Phys. Rev. Lett. 98, 2668021 (2007). 8. W. Dickson, G. Wurtz, P. Evans, D. O’Connor, R. Atkinson, R. Pollard, and A. Zayats, “Dielectric-loaded plasmonic nanoantenna arrays: A metamaterial with tunable optical properties,” Phys. Rev. B 76, 115411 (2007). 9. J. K. Gansel, M. Thiel, M. S. Rill, M. Decker, K. Bade, V. Saile, G. von Freymann, S. Linden, and M. Wegener, “Gold helix photonic metamaterial as broadband circular polarizer,” Science 325, 1513 (2009). 10. B. Hou, H. Wen, Y. Leng, and W. Wen, “Enhanced transmission of electromagnetic waves through metamaterials,” Appl. Phys. A 87, 217 (2007). 11. L. Zhou, W. Wen, C. T. Chan, and P. Sheng, “Electromagnetic-wave tunneling through negative-permittivity media with high magnetic fields,” Phys. Rev. Lett. 94, 2439051 (2005).

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Received 19 Jul 2010; revised 30 Aug 2010; accepted 3 Sep 2010; published 4 Oct 2010

11 October 2010 / Vol. 18, No. 21 / OPTICS EXPRESS 22119

12. J. B. Pendry, L. Mart´ın-Moreno, and F. J. Garcia-Vidal, “Mimicking surface plasmons with structured surfaces,” Science 305, 847 (2004). 13. P. Sheng, R. Stepleman, and P. Sanda, “Exact eigenfunctions for square-wave gratings: application to difraction and surface-plasmon calculations,” Phys. Rev. B 26, 2907 (1982). 14. H. Lochbihler and R. Depine, “Highly conducting wire gratings in the resonance region,” Appl. Opt. 32, 3459 (1993). 15. H. Lochbihler, “Surface polaritons on gold-wire gratings,” Phys. Rev. B 50, 4795 (1994). 16. H. Ghaemi, T. Thio, D. Grupp, T. Ebbesen, and H. Lezec, “Surface plasmon enhanced optical transmission through subwavelength holes,” Phys. Rev. B 58, 6779 (1998). 17. L. Mart´ın-Moreno, F. J. Garc´ıa-Vidal, H. J. Lezec, K. M. Pellerin, T. Thio, J. B. Pendry, and T. W. Ebbesen, “Theory of extraordinary optical transmission through subwavelength hole arrays,” Phys. Rev. Lett. 86, 1114 (2001). 18. D. Grupp, H. Lezec, T. Ebbesen, K. Pellerin, and T. Thio, “Surface plasmon enhanced optical transmission through subwavelength holes,” Appl. Phys. Lett 77, 1569 (2000). 19. S. Darmanyan and A. Zayats, “Light tunneling via resonant surface plasmon polariton states and the enhanced transmission of periodically nanostructured metal films: an analytical study,” Phys. Rev. B 67, 035424 (2003). 20. J. Porto, F. Garc´ıa-Vidal, and J. Pendry, “Transmission resonances on metallic gratings with narrow slits,” Phys. Rev. Lett. 83, 2845 (1999). 21. Q.-H. Park, J. H. Kang, J. W. Lee, and D. S. Kim, “Effective medium description of plasmonic metamaterials,” Opt. Express 15, 6994 (2007). 22. A. Agrawal, Z. Vardeny, and A. Nahata, “Engineering the dielectric function of plasmonic lattices,” Opt. Express 16, 9601 (2008). 23. Q. Cao and P. Lalanne, “Negative role of surface plasmon in the transmission of metallic gratings with very narrow slits,” Phys. Rev. Lett. 88, 0574031 (2002). 24. M. Treacy, “Dynamical diffraction in metallic optical gratings,” Appl. Phys. Lett. 75, 606 (1999). 25. M. Treacy, “Dynamical diffraction explanation of the anomalous transmission of light through metallic gratings,” Phys. Rev. B 66, 1951051 (2002). 26. E. Popov, M. Neviére, S. Enoch, and R. Reinisch, “Theory of light transmission through subwavelength periodic hole arrays,” Phys. Rev. B 62, 16100 (2000). 27. G. P. Ortiz, B. E. Mart´ınez-Zérega, B. S. Mendoza, and W. Mochán, “Effective dielectric response of metamaterials,” Phys. Rev. B 79, 245132 (2009). 28. R. Haydock, “The recursive solution of the Schrödinger equation,” Solid State Phys. 35, 215 (1980). 29. S. L. Alder, “Quantum theory of the dielectric constant in real solids,” Phys. Rev. 126, 413 (1962). 30. N. Wiser, “Dielectric constant with local field effects included,” Phys. Rev. 129, 62 (1963). 31. W. L. Mochán and R. G. Barrera, “Electromagnetic response of systems with spatial fluctuations. i. general formalism,” Phys. Rev. B 32, 4984 (1985). 32. That the long-wavelength response ε M is independent of the direction of q → 0 allows us to use the results of longitudinal calculations to solve optical (i.e., transverse) problems. 33. R. G. Barrera, A. Reyes-Coronado, and A. Garc´ıa-Valenzuela, “Nonlocal nature of the electrodynamic response of colloidal systems,” Phys. Rev. B 75, 184202 (2007). 34. M. Born and E. Wolf, Principles of Optics (Cambridge University Press, 1999), 7th ed. 35. K. Glazebrook, J. Brinchmann, J. Cerney, C. DeForest, D. Hunt, T. Jenness, T. Luka, R. Schwebel, and C. Soeller, “The perl data language v.2.4.4,” Available from http://pdl.perl.org. 36. K. Glazebrook and F. Economou, “Pdl: The perl data language,” Dr. Dobb’s Journal (1997). http://www.ddj.com/184410442. 37. E. Cortés, W. L. Mochán, B. S. Mendoza, and G. P. Ortiz, “Optical properties of nano-structured metamaterials,” Phys. Status Solidi B 247, 2102 (2010). 38. J. B. Keller, “Conductivity of a medium containing a dense array of perfectly conducting spheres or cylinders or nonconducting cylinders,” J. Appl. Phys. 34, 991 (1963). 39. J. B. Keller, “A theorem on the conductivity of a composite medium,” J. Math. Phys. 5, 548 (1964). 40. J. Nevard and J. B. Keller, “Reciprocal relations for effective conductivities of anisotropic media,” J. Math. Phys. 26, 2761 (1985). 41. R. Fuchs, “Theory of the optical properties of ionic crystal cubes,” Phys. Rev. B 11, 1732 (1975). 42. P. Johnson and R. Christy, “Optical constant of noble metals,” Phys. Rev. B 6, 4370 (1972). 43. W. L. Mochán and R. G. Barrera, “Intrinsic surface-induced optical anisotropies of cubic crystals: local-field effect,” Phys. Rev. Lett. 55, 1192 (1985).

1.

Introduction

Metallic films with sub-wavelength nanometric holes may display an extraordinarily large transmittance at near infrared frequencies for which the metal is opaque and light waves are not #131887 - $15.00 USD

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expected to propagate within the holes [1, 2]. The anomalous transmission and other extraordinary optical properties of nano-structured metallic films open the possibility of tailored design for many applications that include hyperlens far-field-subdiffraction imaging [3–5], cloaking [6], optical antennas [7, 8], and circular polarizers [9]. Thus, understanding the electromagnetic properties of this sometimes called plasmonic metamaterials has become important. Many different approaches have been proposed to describe the extraordinary optical properties of some structured systems [10–26]. Some resonances have been identified with surfaceplasmon-polaritons (SPP’s) that may be excited by light after being scattered by the system. In a recent work [27] the frequency-dependent complex macroscopic dielectric-response tensor εiMj (ω ) of 2D-periodic lattices of cylindrical inclusions with arbitrarily shaped cross sections embedded within metallic hosts were obtained in the local limit and enhanced transparency was obtained without invoking explicitly a SPP mechanism. In this paper we develop Haydock’s recursive scheme [28] to obtain the macroscopic dielectric response of 2D and 3D periodic metamaterials in the long wavelength limit. Our method yields a speed improvement of several orders of magnitude over that of Ref. [27], which allows previously untractable calculations for 3D structures with arbitrary geometry, such as interpenetrated inclusions made out of dispersive and dissipative components. We show that the geometry of the inclusions and of the lattice might lead to very anisotropic optical behavior and to a very generic enhanced transmittance for metal-dielectric metamaterials whenever there are only poor conducting paths across the whole sample. 2.

Theory

We consider a periodic lattice of arbitrarily shaped nanometric inclusions (b) embedded within a homogeneous material (a). We assume that each region α = a, b is large enough to have a well defined macroscopic response εα which we assume local and isotropic, but much smaller than the free wavelength λ0 = 2π c/ω with c the speed of light in vacuum and ω the frequency. The microscopic response is described by

ε (rr ) = εa − B(rr )εab

(1)

where εab ≡ εa − εb and B(rr ) = 0 or 1 is the characteristic function for the b regions, which R} the Bravais lattice of the metamaterial. The we assume periodic, B(rr + R ) = B(rr ), with {R E (rr ) may be written in reciprocal space as constitutive equation D (rr ) = ε (rr )E D G (qq) = ∑ εG G E G (qq),

(2)

G

where D (rr ) is the displacement field and E (rr ) the electric field, D G (qq) and E G (qq) the correG, q is the conserved Bloch’s vector and {G G} sponding Fourier coefficients with wavevector q +G the reciprocal lattice. Here, εG G is the G − G Fourier coefficient of ε (rr ). We consider now a longitudinal external field E ex (rr ) = −∇φ ex (rr ) and we neglect retardation within the small unit cell, so we may assume that the total electric field is longitudinal ˆG ˆ · E G, E G → E LG = G

(3)

ˆ . We remark that our final results where we simplify our notation denoting (qq + G )/|qq + G | by G ex L apply as well to transverse fields. As ∇ × E = ∇ × D = 0 and ∇ · E ex = ∇ · D L = 4πρ ex , we may identify E ex with D L which we chose as a plane wave with wavevector q without small lengthscale fluctuations, D LG =0 (qq) = 0. Substituting Eq. (3) into the longitudinal projection of Eq. (2) allows us to solve for ˆ · D L0 , (4) E L0 = qˆ η0−1 0 q #131887 - $15.00 USD

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where we first invert

ηG G ≡ Gˆ · (εG G Gˆ )

(5)

and afterwards take its 0 0 component. The macroscopic longitudinal field E LM is obtained from E L by eliminating its spatial fluctuations, i.e., E LM = E L0 . Similarly, D LM = D L0 . Thus, from Eq. (4) we identify the longitudinal projection of the macroscopic dielectric response ˆ, ˆ ξ qˆ = qˆ η0−1 ε −1 ML ≡ q 0 q

(6)

L defined through E LM = ε −1 ML · D M . A more formal derivation of this result, first obtained for natural crystals [29, 30], may be found in [31]. We should emphasize that ξ in Eq. (6) depends in general on the direction of q . Calculating ˆ qˆ · ε −1 ˆ qˆ for several propagation directions qˆ [32] we may obtain all the components ε −1 ML = q M ·q of the full inverse long-wavelength dielectric tensor ε −1 M and from it ε M . Properties such as reflectance and transmittance may then be calculated using standard formulae [33, 34]. Equation (6) is the starting point for a very efficient numerical calculation of the macroscopic response of nanostructured metamaterials, as developed in the next section.

3.

Haydock’s recursive method

To calculate efficiently the macroscopic response, we start by taking the Fourier transform of Eq. (1), εG G = εa δG G − εab BG G , where BG G ≡ BG −G G = (1/Ω)

G−G G )·rr d 3 r ei(G

(7)

v

describes the geometry of the inclusions which occupy the volume v within the unit cell of volume Ω. In particular, B0 0 = v/Ω ≡ f is the filling fraction of the inclusions. From Eq. (5) we obtain 1 1 −1 ηG−1G = [uδ − BLL = G , (8) GG ] εab G G εab G G where GG G (u) are the matrix elements of a Green’s operator Gˆ, i.e., the resolvent of the operator (u − Hˆ ), corresponding to a Hamiltonian Hˆ with elements ˆ · (B Gˆ ), HG G ≡ BLL =G GG GG

(9)

and the spectral variable u(ω ) ≡ (1 − εb (ω )/εa (ω ))−1 plays the role of energy. From Eqs. (6) and (8) we obtain ξ = G0 0 (u)/εab = 00|Gˆ(u)|00/εab , where |00 represents the state corresponding to a plane wave with wave vector q, and we denote in general the state corG. This allows the use of Haydock’s responding to a plane wave with wave vector q + G as |G recursive scheme [28] to obtain the projected Green’s function and thus the macroscopic response. We set |−1 = 0, |0 = |00, b0 = 0 and recursively define the states |n through |n ˜ = Hˆ |n − 1 = bn−1 |n − 2 + an−1 |n − 1 + bn |n.

(10)

The coefficients an and bn can be recursively obtained by demanding that the states |n be ˜ = n − 1|Hˆ |n − 1 and normalized and orthogonal to each other, yielding an−1 = n − 1|n ˜ n ˜ − a2n−1 − b2n−1 . b2n = n| From Eq. (9) we notice that the action of Hˆ includes first of all a multiplication with a ˆ δ , which is trivially done in reciprocal vector (in cartesian space) operator with elements G GG space. Afterwards we have to multiply with a scalar operator with elements BG G . However, from Eq. (7), this product is actually a convolution. Thus, a fast Fourier transform allows us to #131887 - $15.00 USD

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perform this operator in real space simply as a multiplication with the characteristic function ˆ δ . B(rr ). Finally, we transform back into reciprocal space and apply an inner product with G GG This way, by going back and forth between real and reciprocal space we can calculate the action of Hˆ performing only trivial multiplications with diagonal matrices. We implemented the calculation of Haydock’s coefficients using the Perl Data Language [35, 36]. In the basis {|n} the Hamiltonian is represented by a symmetric tridiagonal matrix with main diagonal elements {an } and sub/sup-diagonal elements {bn }. Then, we may write Gˆ−1 = u − Hˆ recursively in blocks as An Bn+1 , (11) Gn−1 = −1 T Bn+1 Gn+1 with An = (u − an ) and Bn = (−bn , 0, 0, · · ·), and where Gˆ ≡ Gˆ0 . Here we used calligraphic letters to denote any matrix except 1 × 1 matrices which are equivalent to scalars. Now we write Gn in blocks as Rn Pn Gn = , (12) Qn Sn where Rn =

1 1 = , −1 T An − b2n+1 Rn+1 An − Bn+1 Gn+1 Bn+1

(13)

obtained by inverting Eq. (11). Iterating Eq. (13) backwards we obtain G0 0 (u) = R0 given recursively through the continued fraction

ξ=

R0 u = εab εa u − a − 0

1 b21 u−a1 −

.

(14)

b2 2

b2 u−a2 − 3

..

.

which may be evaluated very efficiently. Further details of this calculation in the 2D case may be found in Ref. [37]. . The depenNotice that Haydock’s coefficients depend only on the geometry through BLL GG dence on composition and frequency is completely encoded in the complex valued spectral variable u. Thus, for a given geometry we may explore manifold compositions and frequencies without recalculating Haydock’s coefficients. Furthermore, our method is equally suited to the calculation of the properties of systems made up of non-dispersive, transparent materials as well as dispersive, opaque and dissipative media; we simply have to provide the appropriate (complex) value for u. 4.

Results

To test our calculation scheme, in Fig. 1(a) we show the macroscopic response εM calculated for a simple cubic lattice of dielectric spheres, with dielectric function εb = 4, embedded within a dielectrically inert host εa = 1 for various filling fractions f . We compare the results of our formulation based on Haydock’s recursive scheme (H) with the standard Maxwell-Garnett (MG) and Bruggeman (B) effective medium theories for 3D. As expected, the three theories coincide at low filling fractions, but not surprisingly, they differ as f increases, with our results being between MG which lies below and B which lie above. Bruggeman theory treats both a and b media on an equivalent basis, while medium a plays the role of the host and medium b that of

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2.2

2.2 3D

B

B

1.8

1.8

MG

H

M

2

2D

MG

1.6 H

1.4

1.4

1.2

1.2 (b)

(a) 1

1.6

M

2

0

0.1

0.2

0.3

0.4

Filling fraction f

0.5

0

0.1

0.2

0.3

0.4

1 0.5

Filling fraction f

Fig. 1. (a) Macroscopic dielectric function εM for a simple cubic array of spherical inclusions with response εb = 4 within a host with εa = 1 as a function of filling fraction f calculated with our Haydock’s recursive method (H), with Maxwell-Garnett’s (MG) and with Bruggeman’s formulae (B). (b) εM normal to the optical axis of a square array of square prisms with diagonals along the sides of the unit cell as a function of f for the theories H, MG and B in 2D.

inclusions in both Maxwell-Garnett theory and our’s. Thus the MG results are closer to ours than B. Similarly, in Fig. 1(b) we show the response εM normal to the optical axis of a system composed of a 2D square array of square prisms whose diagonals are oriented along the conventional unit cell sides, i.e., the prisms are rotated 45◦ with respect to the unit cell. As in the 3D case, we took εb = 4 and εa = 1. We compare our theory with the 2D MG and B formulae. As in the 3D example, the three formulations coincide in the dilute f → 0 limit. For larger f , MG yields the lowest εM and B the largest. At the percolation limit f = 0.5, at which the edges of nearby prisms touch each other, the geometry does not distinguish between inclusions and host. At this limit, our results coincide with the symmetric Bruggeman 2D formula. Thus, in this particular 2D case, B lies closer to our results than MG. Furthermore, due to the symmetry √ of this system at f = 0.5, Keller’s theorem [38–40] implies that εM = εa εb , which is exactly fulfiled by both B and H theories. In a recent paper [27] an alternative, essentially exact but much more numerically expensive method to calculate the macroscopic response of metamaterials was developed and it was applied to manifold 2D systems. The results were satisfactorily tested against previous calculations and validated through the fulfilment of Keller’s reciprocal theorem [38–40]. We have repeated all the calculations in that paper and have verified that our results agree in the long wavelength limit and are about four orders of magnitude faster. The speed enhancement is much larger in 3D. Further 2D calculations with the present formulation were presented in Ref. [37]. In Fig. 2(a) we show the normal-incidence reflectance of an isotropic semi-infinite metamaterial consisting of a simple cubic lattice of small cubical voids embedded with filling fraction f = 0.6 within a Drude conductor with plasma frequency ω p and a low dissipation parameter Γ = 0.01ω p . The low frequency reflectance is almost unity but attains a deep minimum at ω ≈ 0.51ω p and becomes very small for 0.74 < ω /ω p < 0.91, it attains a peak at 0.97ω p and

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1

2

R T T Tef Tef

R, T

0.6

low low high low high

1

M

0.8

Re[M ] Im[M ]

0.4

0

0.2

(a) 0 0.2

(b) 0.4

0.6

0.8

ω/ωp

1

1.2 0.2

0.4

0.6

0.8

1

-1 1.2

ω/ωp

Fig. 2. (a) Normal-incidence reflectance R of a semi-infinite metamaterial and transmittance T of a d = 200 nm film made of a model Drude conductor with low (Γ = 0.01ω p ) and high (Γ = 0.1ω p ) dissipation parameters with a simple cubic lattice of cubical cavities of filling fraction f = 0.6, as a function of the frequency ω , together with the corresponding transmittance Teff of an homogeneous Drude film with an effective width deff = (1 − f )d. (b) Macroscopic dielectric response of the low dissipation metamaterial. The vertical lines indicate the frequencies at which εM ≈ 1.

finally decreases at high frequencies. This behavior may be understood by looking at Fig. 2(b), which shows the corresponding macroscopic dielectric function. We remark that at low frequencies εM is negative and large corresponding to a metallic behavior, as the interstitial conducting phase percolates. Nevertheless, at larger frequencies there is a series of resonances originated from the coupled plasmon modes of the cubic inclusions [41]. Thus, the macroscopic response becomes dielectric like and approaches at some intermediate frequencies the vacuum value εV = 1. In the figure we have indicated the frequencies ω ≈ 0.80ω p and 0.86ω p for which Re[εM ] = 1 and the metamaterial would become perfectly transparent were it not for its small Im[εM ]. Similarly, at ω ≈ 0.52ω p there is a resonance below which Re[εM ] approaches 1, corresponding to a sharp minimum in the reflectance followed by a sharp maximum due to resonant absorption. Above the last resonance, at 0.94ω p , there is a narrow frequency band where Re[εM ] < 0, corresponding to a maximum in R. Finally, asymptotically [εM ] → 1 and R → 0. The figure also shows the transmittance T of a d = 200 nm wide film made of the same metamaterial. We notice that below ω p it is almost null except for a band where it takes appreciable values and which corresponds to the frequencies where the reflectance R of the corresponding semi-infinite system becomes small. The interpretation of the optical properties of a film is encumbered by the interference between the multiply reflected waves and their decay as they cross the film. Thus, T shows a very rich structure originated from resonant oscillations in the √ imaginary part of k = (ω /c) εM below ω p and Fabry-Perot resonances above ω p . Increasing dissipation diminishes both types of oscillations, as illustrated in the figure for Γ = 0.1ω p without eliminating the transmission band. Finally, as a reference we show the transmittance of low and high dissipation homogeneous Drude films of effective width deff = (1 − f )d = 80 nm, so that it contains the same amount of metal as the metamaterial. Notice that for both cases the transmittance of the metamaterial can be enhanced by about two orders of magnitude above that of the effective film. In Fig. 3(a) we show the transmittance T of a film made of a simple cubic lattice of spherical dielectric inclusions with response εb = 4 within an Au host [42]. We chose the radius as r = 0.6a, with a the lattice parameter, so the spheres overlap their neighbors. We have normalized the results to the transmittance Teff of an effective homogeneous Au film of width deff , in order to emphasize the transmittance enhancement due to the metamaterial geometry. Several enhancement peaks between one and two orders of magnitude are visible in the transmittance

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λ0 (nm) 2480

60

(a)

50

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λ0 (nm)

826

620

496 1240

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T = Tx = Ty

775

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Fig. 3. Normal-incidence transmittance Tα for α = x, y polarization vs. frequency ω for 200 nm Au films with faces normal to the z axis with an embedded lattice of dielectric inclusions, normalized to the transmittance Teff of a homogeneous Au film with the same amount of metal. (a) Simple cubic lattice of spheres of radius r = 0.6a with a the lattice parameter with εb = 4. (b) Simple orthorhombic lattice of z-oriented cylinders with radius r = 0.53ax , height h = 0.9az and dielectric response εb = 2 with lattice parameters ax = az and ay = 1.15ax .

spectrum, corresponding to the excitation of coupled multipolar plasmon resonances within the region where the metal is opaque. In Fig. 3(b) we show the normalized transmittance Tα /Teff (α = x, y) for plane polarized light normally incident on a 200 nm film lying on the xy plane made of a simple orthorhombic lattice of z-oriented dielectric cylinders with radius r = 0.53ax , height h = 0.9az dielectric function εb = 2 and lattice parameters ay = 1.15ax and az = ax within an Au host. There is a huge anisotropy, with a peak enhancement of Ty almost two orders of magnitude larger than that of Tx . For this geometry there is an overlap between neighbor cylinders along x, so the system is a better low frequency conductor along y. We should discuss some limitations of our theory. Our assumption of periodicity is not too restrictive, as the efficiency of our scheme permits calculations for complex unit cells with many inclusions. Our use of the long-wavelength approximation does impose a bound on the size a of the unit cell. Retardation corrections are expected to be of order (a/λ0 )2 and we have verified through comparisons to exact calculations in 2D [27] that for a ≤ λ0 /5 they may be considered negligible for many purposes. The interesting regions in Figs. 3(a) and 3(b) lie below ∼ 2 eV, where our macroscopic response would be accurate up to a ∼ 100 nm, a scale easily attainable with current fabrication techniques. Nevertheless, our theory does neglect the presence of transition layers of width dt ∼ a close to the metamaterial boundaries which are not well described by a local, position independent macroscopic response εM , as inclusions closer to the surface than their size would be sliced by the surface, and the effect of the surface on the microscopic fluctuations of the field extendes a distance ∼ a. Similar transition regions are present at ordinary crystalline and spatially dispersive materials and are known to produce corrections to the optical properties of semi-infinite media of the order of dt /λ0 [43]. For a freestanding subwavelength thin film of width d , the corresponding corrections can become of the order of dt /d. In our case, both corrections would be of only a few percent if the inclusions were of the order of a few nanometers. In summary, we developed a formalism that allows very efficient calculations of the macroscopic response of 3D nano-structured periodic metamaterials. We applied it to films made of various lattices of dielectric inclusions with assorted shapes within opaque metallic hosts and found an extraordinary enhancement of the transmittance as a generic property whenever

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Tα /Tef

T /Tef

40

the metamaterial is conducting at low frequencies and has dielectric like resonances at larger frequencies. Acknowledgment This work was supported DGAPA-UNAM (IN120909 (WLM)), CONACyT (48915-F (BMS) and by ANPCyT-UNNE (204 y 190-PICTO-UNNE-2007 (GPO)).

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