Abstract

In this paper we demonstrate that rigorous high-order perturbation of surfaces (HOPS) methods coupled with analytic continuation mechanisms are particularly well-suited for the assessment and design of nanoscale devices (e.g., biosensors) that operate based on surface plasmon resonances generated through the interaction of light with a periodic (metallic) grating. In this connection we explain that the characteristics of the latter are perfectly aligned with the optimal domain of applicability of HOPS schemes, as these procedures can be shown to be the methods of choice for low to moderate wavelengths of radiation and grating roughness that is representable by a few (e.g., tens of) Fourier coefficients. We argue that, in this context, the method can, for instance, produce full and precise reflectivity maps in computational times that are orders of magnitude faster than those of alternative numerical schemes (e.g., the popular “C-method,” finite differences, integral equations or finite elements). In this initial study we concentrate on the description of the basic principles that underlie the solution scheme, including those that relate to analytic continuation procedures. Within this framework, we explain how, in spite of conventional wisdom to the contrary, the resulting perturbative techniques can provide a most valuable tool for practical investigations in plasmonics. We demonstrate this with some examples that have been previously discussed in the literature (including treatments of the reflectivity and band gap structure of some simple geometries) and extend this to demonstrate the wider applicability of the proposed approach.

© 2013 Optical Society of America

1. INTRODUCTION

The interest in surface plasmons (SP), waves that propagate at the surface of a conductor [1,2], stems largely from their tight confinement of optical energy which allows for device miniaturization and sensitivities that are well beyond those achievable by current diffraction-limited technologies. The potential application of SP-based structures encompasses a variety of areas, including nanoscale photonic circuits [3], optical data storage [4], near-field imaging [5], photolithography [6], biosensing [7], etc. The research on and use of nanoplasmonic devices has seen a significant surge over the last twenty years as new phenomena have consistently been discovered and fabrication and characterization techniques have become ever more sophisticated and precise, to the point where surface roughness can be controlled to an accuracy below 1 nm (see e.g., [8] and the references therein). These advances, in turn, have provided an impetus for the development of accurate simulation tools that can resolve the effects of such small features and thus aid in the design process. A number of methodologies, both classical and specific to the relevant configurations, have been applied in this context, with varying degrees of success [912]. In this paper we show that high-order perturbation of surfaces (HOPS) methods enhanced with techniques of analytic continuation are well-suited for the analysis and design of nanoscale optical devices, and particularly so for those wherein the coupling of light to the metal is enabled through (periodic) corrugations on the surface of the latter.

The recent history of attempts at incorporating simulation work into the characterization of the devices that rely on the excitation of surface plasmon resonances (SPR) is very rich in number and variety of techniques. Indeed we find, for instance, that methods based on widely applicable numerical procedures for the modeling of scattering experiments, such as finite elements (see e.g., [1317]), finite differences (e.g., [1823]) and integral equations (e.g., [2428]) have all been used in this context. As is the case in every other application and while, as we mentioned, these methods provide a wide domain of applicability, they also display significant challenges. Finite elements and finite differences, for instance, entail the (very fine, in most cases) discretization of entire domains (with the resulting computational expense) even if, as it is typically the case, the optical characteristics of the device are piecewise homogeneous. Moreover, and as a consequence of this, the computational domain must be artificially truncated, which gives rise to the need for the selection of appropriate “absorbing (or transparent) boundary conditions,” “perfectly matched layers” or other approximate treatment (see e.g., [2931]). (Surface) integral equation based methods, on the other hand, avoid both of these impediments, as they rely on surface discretizations and their formulation intrinsically encodes the outgoing character of diffracted waves. These advantages, however, are attained at the expense of introducing full, i.e., not sparse, (impedance) matrices whose numerical inversion can render the methods uncompetitive, unless appropriate acceleration schemes are introduced (see, for instance, [24,28,32,33], and the references cited there). Furthermore, in the case of (infinite) periodic surfaces, a most significant challenge arises from the efficient evaluation of the appropriate Green function (see e.g., [34] and the references therein).

For the specific configurations provided by infinitely periodic (“rough”) surfaces, a number of alternative procedures have been devised (see e.g., the reviews in [3537]). In this connection, a particularly popular method in nanoplasmonics is the “Chandezon Method” (or “C-method”) [3840], which has been successfully applied in numerous studies (see e.g., [4149] among many others). The method, and related schemes [40,50] (such as the “Rigorous Coupled Wave Analysis” (RCWA) and the “Fourier Modal Method” (FMM) [5153]), relies on a simple coordinate transformation (or successive slicing, in the case of RCWA) that “flattens” the interfaces to simplify the imposition of boundary/transmission conditions. This, however, is done at the expense of foregoing the classical Rayleigh expansion [which is no longer representative of the field in the new coordinates; see Eq. (10) below], and it thus necessitates the derivation of new basis functions. This, in turn, can be posed as an eigenvalue problem, with the consequently high associated computational cost (see e.g., [50,54]).

Boundary perturbation schemes, such as the “small perturbation method” (SPM) [55], the “small slope approximation” (SSA) [56], or the “unified perturbation method” [57], on the other hand, typically rely on the explicitness of the Rayleigh expansion (or other representations, such as that provided by the T-matrix) to rapidly deliver scattering results. Most often, these techniques are used in low-order instances (see e.g., [5860]) which restricts their domains of validity to very small deformations of a planar interface. Attempts at expanding these domains have been pursued through the derivation of higher-order approximations [6165], but these can often result in only slight improvements. In spite of the long-held view that this is a consequence of the failure of the so-called “Rayleigh hypothesis” (RH) [36,6668] for larger deformations, it has been rigorously shown that the latter plays no role in the intrinsic limitations of these procedures, as the fields can be proved to be complex analytic functions of a parameter measuring (analytic) boundary deformations for values of this parameter that include a complex neighborhood of the entire real line [69]. As a consequence, it follows that the limitations on these higher-order attempts are solely due to the finite radius of convergence of the resulting power series, that is, they arise as a result of the insistence on straightforward summation of the latter.

The observation that the field can, in fact, be analytically extended to a neighborhood of the real line, on the other hand, suggests that, while the totality of the information on the fields as the boundary perturbation parameter is varied is encoded in its Taylor series, a direct summation of the latter is not an optimal strategy. Based on these conclusions, a suite of novel numerical procedures was devised [7073] that introduced mechanisms of analytic continuation (e.g., conformal maps [70], Padé approximation [71]) to significantly enlarge the domain of applicability of perturbative approaches. Indeed, it is now recognized that “… these results are of fundamental importance…” [74], and that the consequent algorithms constitute “… efficient numerical schemes…” that “… improve drastically the accuracy and the range of the SPM…” [36,58].

In what follows we show that the specific characteristics of these HOPS methods are well-suited to the evaluation of grating-mediated SPR generation. Indeed, as is well established, SPR arise in regimes wherein the height h of the grating coupler is small compared to the period d, and the latter is comparable to the wavelength. Moreover, rich and varied reflectivity maps can be attained with the use of gratings that are representable by only a handful (tens) of Fourier modes, thus limiting their slope. As we explain below, in this regime the HOPS methods provide unparalleled performance and thus constitute excellent candidates to guide virtual design studies.

The rest of the paper is organized as follows. First, in Section 2 we set up the problem, describe our notation, and recall some basic facts on the origins and characteristics of surface plasmon polaritons (SPP). Section 3 is devoted to a brief review of the algorithms in the context of plasmonics, including a discussion of the materials and material properties that are relevant in the visible regime. The performance of the numerical procedure in this realm is analyzed in Section 4, where a variety of simulations are presented that demonstrate the effectiveness of the proposed procedures. We conclude in Section 5 with a summary and a survey of perspectives for the approach presented herein.

2. PRELIMINARIES: MAXWELL’S EQUATIONS AND SURFACE PLASMONS

To begin, we refer to the configuration in Fig. 1 where we display a generic (metallic) grating underlying a dielectric, and where we define our notation. The grating lines extend in the z direction, and the profile is given by a function,

y=f(x),f(x+d)=f(x),
where d is the period. The plane of incidence contains the grating wavevector, and the incidence angle is denoted by θ; when applicable, the “line width” of a grating will be denoted by w.

 

Fig. 1. Generic grating configuration for SPR coupling.

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As is well known, in (“two-dimensional”) configurations such as that in Fig. 1, Maxwell’s equations decouple into two scalar problems corresponding to the transverse electric (TE) and transverse magnetic (TM) modes of polarization (see e.g., [71]). These are characterized by the alignment of the electric or magnetic fields with the grating lines, respectively. In the case of monochromatic waves oscillating at a single frequency ω,

(E⃗(x,t),H⃗(x,t))=eiωt(E(x),H(x)),x=(x,y,z),
u={Ezfor TE polarization,Hzfor TM polarization,
and discriminating the scattered and incident fields with the superscripts “scat” and “inc,” and the reflected and transmitted portions of the former, respectively, by ±, we have
Δu±,scat+(k±)2u±,scat=0inD±,
{u+,scatu,scat=uinconS,u+,scatnC02u,scatn=uincnonS,
where
C0={1for   TEk+/kfor   TM.
Here, D±={(x,y):±y>±f(x)} denotes the dielectric (+) and metal () domains, S={(x,y):y=f(x)} denotes the interface, n is the unit normal to S into the dielectric, k±=n±ω/c=2π/λ± stands for the wavevectors, λ± for the corresponding wavelengths, n±=ε± are the refractive indices, and ε±=ε±(ω) are the permittivities of each media.

The incident wave, uinc, on the other hand, will be assumed to be a monochromatic plane of unit amplitude wave,

uinc(x)=u(x,y)=eiαxiβy,
where α=k+sin(θ), β=k+cos(θ) and θ denotes the angle of incidence (see Fig. 1). The quasi-periodicity of the incident wave,
uinc(x+d,y)=eiαdeiαxiβy,
implies a similar property for the scattered field,
u±,scat(x+d,y)=eiαdu±,scat(x,y),
and this, in turn, allows for the (Rayleigh) expansion of the latter,
u±,scat(x,y)=r=Br±eiαrx±iβr±y,valid on±y>maxxR{±f(x)},
where
αr=α+rK,K=2πd,βr±=(k±)2(αr)2,ris an integer,and Im(βr±)0.
Note that the conditions in Eq. (11) imply that only a finite number of reflected waves are propagating (namely, those corresponding to indices r in the set U={r:βr+>0}), while the rest decay in the far field.

The objective, then, is to design a scheme to determine the amplitudes Br± of the reflected and transmitted waves that, in this case, will allow for the assessment and design of structures that support SPP. These are particular electromagnetic waves that propagate at the interface between a dielectric and a conductor, and are tightly confined in the direction of the normal to the surface. These surface waves arise through the coupling of the electromagnetic fields to oscillations of the conductor’s electron plasma [1], and they are closely related to (and can be deduced from) the Brewster mode [75,76].

The macroscopic study of SPP begins with the simple realization that, if the interface is flat, their dispersion relation can be explicitly derived to yield [1]

kx,SPP=±k0ε+εε++ε,
where k0=ω/c, and the real and imaginary parts determine the propagation constant and absorption, respectively. This relation, however, also shows that in such a case,
|Re(kx,SPP)|>k+,
provided Re(ε)<0 and |Re(ε)|>ε+ (as is the case with “noble metals” in the visible range of the spectrum), and thus the plasmon cannot be excited. Significant research into this phenomenon has resulted in three main techniques by which the additional momentum can be introduced, namely, the use of prism coupling to enhance the momentum of the incident light [77,78]; scattering from a topological defect on the surface [79,80]; and the use of a periodic corrugation in the metal’s surface [81], as studied herein. Specifically, in relation to the latter, the additional momentum is attained through coupling with a diffracted mode,
Re(kx,SPP)αr,
where αr is defined as in Eq. (11) [1].

In terms of the overall expansions in Eq. (10), this behavior manifests itself through a transfer of incident radiation into modes that decay in the far field, in a manner so that the entire reflected field is significantly suppressed away from the dielectric/metal interface. More precisely, in this context, the condition for the occurrence of a SPR for a given angle of incidence can be identified with the presence of a drop in the reflectivity

RrUβr+β|Br+|21,
or “normalized reflectivity,”
NRrUβr+β|Br+|2|B0+,flat|21,
as the wavelength is varied, where B0+,flat denotes the amplitude of the reflected wave corresponding to a flat interface. For instance, in the particular case wherein the set U consists of a single index r=0 (corresponding to specular reflection), such as is the case when the illumination is normal with a wavelength λ=2π/k+ that exceeds the period d, then this latter condition reduces to
NR=|B0+|2|B0+,flat|21.
More generally we define the “reflectivity at order r” for a specific structure to be |Br+|2, and its normalized version to be the ratio of this value to the corresponding one for a flat surface.

In the next section we present an efficient high-order boundary perturbation method that allows for the rapid evaluation of the quantities in Eq. (15) as functions of the grating height and the wavelength of radiation and/or incidence angle.

3. A FAST, HIGH-ORDER BOUNDARY PERTURBATION

In this section we briefly review the HOPS method that was originally introduced in [7072]; see also [73] for a more comprehensive review of subsequent improvements and applications. As indicated in Fig. 1, here we shall concentrate on “two-dimensional” configurations [71], where the lines of the gratings are assumed to have infinite extent; every development that follows, however, can be extended to fully three-dimensional (biperiodic) structures, at the expense of somewhat more complex derivations [72].

The basic observation underlying the procedure relates to the explicit nature of the solution in the case wherein the interface is entirely flat. This, in turn, suggests that this solution may be analytically continued to that of an undulated surface. More precisely, for a normalized profile f(x) satisfying

maxxR(f(x))minxR(f(x))=1,
we consider an entire family of scattering problems,
Δxu±,scat(x;h)+(k±)2u±,scat(x;h)=0inDh±,
{u+,scat(·;h)u,scat(·;h)=uinc(·;h),u+,scatn(·;h)C02u,scatn(·;h)=uincn(·;h),
on Sh=Dh+Dh, where Dh±={(x,y):±y>±hf(x)}, Sh={(x,y):y=hf(x)} has height h and, as in Eq. (10),
u±,scat(x,y;h)=r=Br±(h)eiαrx±iβr±y,valid on±y>maxxR{±hf(x)}.

As was shown in [71] (see also [69]), the coefficients Br(h) are complex analytic functions of the parameter h (putting to rest a fifty-plus year controversy as to its validity as an actual convergent series; see e.g., [63,67,74,82] and the references therein), and as such these can be expressed in the form

Br±(h)=n=0dn,r±hn,
where the series converges for small values of the parameter h. More importantly, as shown in [69], these coefficients (and the field itself) are complex analytic in a whole neighborhood of the real line; see the example in Fig. 2 (here and throughout, and whenever appropriate, the axes denote measures in nanometers).

 

Fig. 2. (a) (Normalized) reflectivity [“NR” in Eq. (16)] map for a sinusoidal silver profile with period d=650nm at normal incidence in TM polarization; (b) poles (open markers) and zeros (filled markers) of (the [32/32] Padé approximant of) the coefficient B0(h) for complex h and different values of the wavelength of radiation λ (660λ700). Note that no poles are present in a complex neighborhood of the real line (shaded area). The zeros indicated with an arrow correspond to close-to-zero reflectivity, indicating the presence of a plasmon resonance.

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Arguing as in [71] it is straightforward to show that the coefficients dn,r± satisfy the following recursion:

dn,r+dn,r=(iβ)nCn,rk=0n1q=max[kF,r(nk)F]min[kF,r+(nk)F]Cnk,rq[(iβq+)nkdk,q+(iβq)nkdk,q],
iβr+dn,r++C02iβrdn,r=Cn,r(iβ)n1[(iα)(iKr)(iβ)2]+k=0n1q=max[kF,r(nk)F]min[kF,r+(nk)F]Cnk,rq{[iK(rq)](iαq)×[(iβq+)nk1dk,q+C02(iβq)nkdk,q][(iβq+)nk1dk,q+C02(iβq)nkdk,q]},
for n0, from which their values can be readily derived. The recursion assumes that the profile f(x) can be accurately represented with a number F of Fourier modes,
f(x)=p=FFC1,pei2πpx/d,
and the quantities C,p are defined by the relation
f(x)!=p=FFC,pei2πpx/d.
In connection with the recursion (22), (23), we note that its low operation count (resulting in exceedingly fast run times) is owed to the convolution nature of the inner sums (that can therefore be effected with fast Fourier transforms), and to the relatively low orders n that suffice to attain accurate results throughout the domain of interest, if mechanisms of analytic continuation are incorporated. In relation with the latter, and following [71], we have chosen to rely on the Padé approximants that can be readily evaluated from knowledge of the Taylor coefficients dn,r± [Eq. (21)]. In this regard, throughout this paper we shall use the standard notation “[L/M]” to refer to a Padé approximant (i.e., a rational function approximation) with polynomials of degrees L and M in the numerator and denominator, respectively; we refer the interested reader to the treatise [83] for the mechanics of evaluation and fascinating properties of Padé approximation.

Clearly, the results are influenced by the dispersion model that is applicable in each instance. Figure 2(a), for example, was derived using a “classical” Lorentz model [84] that provides a characterization of the frequency-dependent permittivity in silver. Alternative models that depend (rather weakly) on the preparation of the sample include, for instance, the polynomial expressions of [85] and [86]. Models can also de derived from the well-known discretely sampled values of Johnson and Christy [87] or of Lynch and Hunter [88] (e.g., through suitable interpolation or by matching to an appropriately defined Lorentz model).

As is well-recognized [89] the appearance of classical SPR is confined to a rather well delineated subset in the parameter space consisting of the wavelength, angle of incidence, period, and the shape of the profile itself. Indeed, these grating-mediated SPR occur for small to moderate modulations that are much smaller than the wavelength of radiation which, in turn, is comparable to the period. The latter condition follows readily from the requirement that coupling occur at low (0,±1,±2,±3,) diffracted orders. Further increases in height, on the other hand, will typically conspire against efficient SPP coupling due to the concomitant changes in the phase of reradiated light [43,90]. It should be noted, however, that very deep gratings may still give rise to localized field enhancements, though this can be qualitatively distinguished as arising from waveguide and hybrid waveguide-SPP modes (see e.g., [85,9193] and the references therein).

As we show next, and by its very nature, the HOPS method presented here is particularly advantageous for simulations of classical SPR, whose sharp features and high sensitivity to changes in the low amplitude harmonics of the interface make them ideally suited for applications that seek to quantify either the surface shape or the permittivities of the constituent media [43].

4. NUMERICAL RESULTS

In this section we shall present a variety of numerical results that demonstrate the accuracy and versatility of the proposed scheme for the analysis of grating-mediated SPR generation. As is well known, these resonances occur in the TM polarization [1], so we shall henceforth constrain our experiments to such illuminations [C0=k+/k in Eq. (6)]. More precisely, we begin in Section 4.A with a demonstration of the benefits of incorporating mechanisms of analytic continuation into the HOPS methods within this context. Indeed, we show that these techniques can have a dramatic effect not only in expanding the domain of applicability of perturbative approaches [cf. Fig. 2(b)] but that they also effect an intrinsic acceleration of the convergence of the corresponding Taylor series even when the perturbations are sufficiently shallow to allow for its direct summation. Next, in Section 4.B we expand on the examples of the previous section to demonstrate the effectiveness of the approach in computing full reflectivity maps, as well as field representations. Finally, this demonstration is complemented in Section 4.C. where experiments on the (three-dimensional) dependence of the reflectivity on height, wavelength. and angle of incidence are shown to recover the appearance of “energy (photonic) band gaps.”

A. Analytic Continuation

In this section we present some results that demonstrate that analytic continuation mechanisms significantly expand on the applicability of boundary-perturbation methods and that they allow for a complete coverage of the range of SPP excitations, as described at the end of the previous section [allowing, in fact, for calculations that are well-beyond the regime of “shallow” modulations (see e.g., Fig. 5 below)]. In particular, as we show here, these procedures negate a variety of arguments that have been advanced over the last few decades (and that continue to be so) in connection with the suitability of perturbative approaches for SPP simulations (and, more generally, for simulations concerning diffractive nano-optical structures); see e.g., [41,94].

Our first example is motivated by that in [94] where it is argued that the results from perturbation theory display a behavior that ranges from differing “… from those of Maradudin’s and Rayleigh’s methods not only in magnitude but also in the angle…” for small perturbations (6 nm) to differences “… by 1 order of magnitude…” for slightly deeper profiles (10 nm). Unfortunately, it seems impossible to reproduce the specific results presented there with the description provided by the authors, as it does not appear feasible to excite a resonance at an incidence angle of 19.31° with a wave of energy 3.5 eV that illuminates a sinusoidal silver grating,

y=h2sin(2πdx),
of period d=400nm with optical constants garnered from [87] (we speculate that a typographical error may have been introduced before publication). Still, in an attempt for a more thorough assessment of the behavior of perturbative methodologies within this regime, and particularly of the case where analytic continuation schemes are integrated into these, we have chosen to retain the sinusoidal nature of the silver grating as well as its period. We find, however, that the constraints of choosing an instance wherein (i) a resonance through coupling to the 1 order occurs in the vicinity of the incidence angle used in [94], and (ii) measured data can be found in the work of Johnson and Christy [87], leaves little choice but to select an incidence with an energy of 2.26 eV (corresponding to a wavelength λ=548.60nm), for which the data in [87] is
n=0.060+3.586i.
Then, from Eq. (14) (with r=1), we get that for a flat interface a resonance occurs at an angle of 19.2829° (19.31°).

Under these conditions, in Fig. 3 we present, as in [94], approximations to the square of the absolute value |B1+|2 of the 1st order Rayleigh coefficient B1+ [Eq. (10)] as garnered from the perturbative treatment described in Section 3 for heights h=4, 6 and 10 nm [Figs. 3(a)3(c), respectively]. As in [94], the figure shows a clear deterioration of the results if a series with N=5 terms is used for increasing heights. In fact, the results corresponding to the deepest grating [Fig. 3(c)] show that such an approximation can differ qualitatively from more accurate representations corresponding to approximations of orders N=7, 9, and 11, which clearly converge in this case. More importantly, perhaps, the results also show that the limitations alluded to in [94] are solely due to the insistence on approximating the analytic continuation of B1+=B1+(h) through a Taylor series. Indeed, as the figure shows, a simple change to Padé approximation of order [2/2] (which uses exactly the same information as the Taylor series of order N=5) provides results with several digits of accuracy. Indeed, for instance, Fig. 4 shows the maximum relative error in the value of |B1+|2 for h=10nm over the interval [19.2°,19.4°] for Taylor and Padé approximations of orders N and [(N1)/2,(N1)/2], respectively, when these are compared to the results from Padé approximants of order [12/12] (N=25), and it shows that the results of the [2/2] approximants have a maximum error of 3.3×104, guaranteeing three-digit accuracy; similar calculations show that the accuracy of the [2/2] approximant for h=4nm and h=6nm is five and four digits, respectively. These results constitute an instance of the property of analytic continuation mechanisms that we alluded to above, namely, the fact that these procedures accelerate convergence to the exact solution. This is clearly exemplified in Fig. 4 which shows, for instance, that a Taylor approximation of order N=13 still incurs an error of about 4.9×103 while the corresponding [6/6] Padé approximant displays close to full double precision accuracy.

 

Fig. 3. (a) Square of the absolute value |B1+|2 of the 1st order Rayleigh coefficient B1+ in Eq. (10) (“reflectivity at order 1”) for a sinusoidal silver grating of pitch d=400nm and height h=4nm, illuminated with a plane wave of energy 2.26 eV, corresponding to a wavelength λ=548.60nm, as a function of incidence angle; (b) same as in (a) but with height h=6nm; (c) same as in (a) but with height h=10nm.

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Fig. 4. Logarithm (base 10) of the maximum relative error in |B1+|2 (circles) and best linear fit (dashed line) corresponding to incidences in the interval [19.2°,19.4°] for the case in Fig. 3(c) (h=10nm). The optical constants used in the simulations are from [87].

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To complement the above, for our second example we retain the sinusoidal nature of the (silver) grating, but we explore a domain that lies well beyond that where low-order perturbative approaches are applicable. In fact, as we show, the fields corresponding to this configuration do not admit a high-order representation in terms of Taylor series, and Padé approximation (or other analytic continuation mechanisms) are essential to attain convergence. The configuration is taken from the work of Chen et al. [85] and is depicted in Fig. 5(a): a deep silver structure with period d=258nm and height h=124nm, illuminated at an angle of 30°; the permittivity of the metal is assumed to be given by the polynomial model on p. 1574 of [85]. Figure 5(b) shows the computed reflectivity at order zero, |B0+|2, as a function of the wavelength throughout the visible range as computed from [32/32] Padé approximants; a dashed line marks the wavelength λ=387nm below which the 1st order is propagating. A numerical convergence analysis suggests that the displayed results have at least three digits of accuracy over the entire range, and at least four digits outside the interval from 405 to 415 nm.

 

Fig. 5. (a) Configuration corresponding to a deep silver grating of height 124 nm and period 258 nm illuminated with a plane wave with an incidence angle of 30°; (b) reflectivity at order 0, |B0|2 [Eq. (10)], as a function of wavelength for a polynomial permittivity [85] as computed from a [32/32] Padé approximant. The dashed vertical line marks the wavelength below which the 1st Rayleigh mode becomes propagating.

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In Fig. 6 we display results corresponding to the configuration in Fig. 5(a) as a function of the height h for wavelengths λ=410nm and λ=495nm. In this connection it is worth noting that, in contrast with other simulation methodologies, the evaluation of the scattering amplitudes at different heights within the boundary perturbation framework is immediate from knowledge of the terms of the Taylor series or of its analytic continuation (as this simply entails a most straightforward and inexpensive evaluation of the approximation at specified points). Figures 6(a) and 6(c) show the reflectivity at order zero as a function of the height h: each figure displays various truncated Taylor series approximations as indicated, together with the corresponding Padé approximants (solid lines). In each case, the value hp corresponds to the absolute value of the closest (noncanceling) pole to the origin as approximated by the corresponding pole of the [32/32] Padé approximant (N=65). These values can be garnered from Figs. 6(b) and 6(d), where we show the complete set of poles and zeros in the complex h plane of this latter approximation: the circles and crosses (square and stars) mark the location of the zeros and poles, respectively, that do (do not) cancel out. As it follows from these figures, the value of hp is strongly dependent on the frequency, and it provides a good approximation to the radius of convergence of the Taylor series. In particular, we see that in each case the series approximation is unable to capture the behavior for deep modulations such as that in Fig. 5(a), while the analytically continued representations (in the form of Padé approximants) do provide us with meaningful results.

 

Fig. 6. (a) Zeroth order reflectivity as a function of height h for the example in Fig. 5 at a wavelength λ=410nm; the solid lines display the values of the Padé approximations corresponding to each value of N for which the Taylor partial sums are shown. The vertical line at h=hp depicts the absolute value of the closest (noncanceling) pole to the origin as approximated by the corresponding pole of the [32/32] Padé approximant (N=65). (b) Poles and zeros in the complex h plane of the [32/32] Padé approximant for the configuration in (a); the circles and crosses (square and stars) mark the location of the zeros and poles, respectively, that do (do not) cancel out. (c) Same as (a) but for a wavelength λ=495nm. (d) Same as (b) for the configuration in (c).

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Our final example in this section, entails a configuration that is far from sinusoidal as it corresponds to a (semi-) elliptical grating, which we assume to be optically described by the dispersion relation that follows from (cubic-spline) interpolation of the results in [88]. A Fourier representation with F=25 [Eq. (24)] was garnered from the exact description of the top portion of an ellipse with major axis of length w=400nm and periodized with period d=630nm; the Fourier coefficients were summed in the Cesàro sense (i.e., using Fejér means) to minimize the oscillatory behavior at corners (see e.g., [95]).

Figure 7(a) shows the configuration to scale, with an amplitude h=20nm; the maximum slope, on the other hand is approximately 32 for this height, and 50 for h=30nm. The results in (b)–(d) correspond to slight changes in the incidence angle, from θ=0° (normal incidence), to θ=0.2° to θ=1° for heights of 20 and 30 nm. The results were obtained from [5/5] Padé approximants (N=11) and the error is no larger than 6×107 and 5×104 for h=20nm and h=30nm, respectively. The corresponding Taylor sums display minimum errors of 25% and 100%, respectively, with N=11; increasing this to N=65 shows that, in fact, the series diverges for wavelengths in the vicinity of the resonances for heights that exceed approximately 15 nm.

 

Fig. 7. (Total) Reflectivity for an elliptical grating configuration, as shown in (a) (period d=630nm, linewidth w=400nm), for different incidence angles. The vertical lines mark the wavelength below which the 1st order becomes propagating.

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It is worth noting, however, that in this case the amplitudes are sufficiently small so as to allow for a comparison with the resonances as predicted by Eq. (12) corresponding to a flat interface. Figure 8 provides a graphical display of these predictions. Indeed, for a grating of period d, Eq. (12) can be rewritten as

r=[±ε+εε++εsin(θ)]dλ,
where r is an integer. Figure 8 displays the right- and left-hand sides of Eq. (28), and the intersections of the curves with the horizontal lines at heights r=±1 mark the wavelengths at which resonances would occur for a flat interface, if these could be excited (compare with Fig. 7).

 

Fig. 8. Left- (horizontal lines) and right- (curves) hand sides of Eq. (28). Their intersections, marked by vertical lines, define the resonant wavelengths that would be excited for a silver interface of zero height and period d=630, as depicted in (a), under the dispersion derived from (cubic-spline) interpolation of the results in [88].

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B. Reflection Data, Reflectivity Maps, and Electromagnetic Fields

In this section we expand on the results above and provide examples of entire “reflectivity maps” as functions of both wavelength and height. As we stated, the latter evaluation of the dependence of the reflectivity on the depth of the structure entails almost no computational effort within the HOPS method we review here, as this only demands the straightforward calculation of a Padé approximant (or other approximations to the analytic continuation of a Taylor series) at a number of points. In addition, we provide examples of actual field distributions as computed from the HOPS method that exemplify the “surface” character and subwavelength localization of the SPP.

To begin, in Fig. 9 we present the total reflectivity maps and an instance of the spatial distribution of the reflected (transverse) magnetic field for a configuration as in Fig. 5(a); similar results for the geometry in Fig. 7(a) are presented in Fig. 10.

 

Fig. 9. Results for a sinusoidal profile of period d=258nm illuminated with an incidence angle θ=30°. (a) Total reflectivity as a function of height and wavelength; the thick dashed vertical line marks the wavelength below which the 1st order becomes propagating. (b) Spatial distribution of the (transverse) reflected magnetic field, |u+,scat|=|Hz+,scat| [Eq. (3)], corresponding to the intersection of the dotted–dashed lines in (a): height h=20nm and wavelength λ=410nm.

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Fig. 10. Same as Fig. 9 for the configuration of Fig. 7(a) with incidence angle θ=1° [Fig. 7(d)]. (a) Reflectivity map; (b) spatial (transverse) magnetic field distribution corresponding to the intersection of the dotted–dashed lines in (a): height h=30nm and wavelength λ=640nm.

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In both cases the results correspond to simulations that rely on the HOPS method using Padé approximants of order [16/16] and spanning heights up to 125 nm. In the case of the elliptical grating, we further consider “negative” heights corresponding to elliptical grooves [rather than elliptical bumps as in Fig. 7(a)]. As in the examples above, we use the polynomial dispersion relation [85] for the sinusoidal profile, and the interpolated data from [88] in the case of the elliptical geometry. We note that, in the case of the sinusoidal structure, strong coupling to an SPP occurs for wavelengths that are slightly above that which marks the propagation of the 1st order mode and only for small to moderate modulations of less than 60 nm (h/d0.23), with truly negligible reflection for heights below 30 nm (h/d0.12). Indeed, Fig. 9(b) shows that for a height h=20nm and a wavelength λ=410nm, a six-fold enhancement of the field strength is attained in a subwavelength region near the interface, with nearly vanishing reflected fields a wavelength away. Similar results are obtained for the elliptical profile, where the period is significantly larger and so are the heights for which SPPs can be excited. Interestingly, for deeper profiles the longer wavelength branch that is present for small heights, and that can be identified with one that would be present on a flat interface [Fig. 8(d)], ceases to exist, as the resonant character of the structure begins to differ significantly from that which is derived from approximation by a planar surface. As shown in Fig. 10(b), at a height of 30 nm and a wavelength of 640 nm the local enhancement of the field is approximately ten-fold.

C. Photonic Energy Gaps (“band gaps”)

In this final section we present results that are similar in nature to those described in Section 4.B but where we interchange the variations of the height at a fixed incidence angle for variations of the incidence angle at a fixed height. In this manner, we are able to investigate the existence and dependencies of “band gaps” in the spectrum for SPP excitation (see [96] and the references therein). As with our previous examples, again here our results can be interpreted as countering the belief [41] that perturbation treatments lead to a “… limited range of applicability…” as “… shown by da Silva et al. [94]…” (see the first example in Section 4.A) and that perturbation methods that are “… applicable to larger amplitude gratings… must be in question…”.

For comparison purposes, the examples we present follow closely those in [41], where the effect of a second harmonic [F=2 in Eq. (24)] on the reflectivity of purely sinusoidal gratings is analyzed with the method of Chandezon et al. [39]. More precisely, we shall consider profiles of the form

f(x)=asin(Kx)+bsin(2Kx),(K=2πd),
where a and b are constant, and d=634nm (see Fig. 11). Specifically, in our first set of experiments, in Figs. 12 and 13, we consider a purely sinusoidal grating with a=5nm (b=0, height h=10nm), as well as a modulated sinusoid with b=2nm (h12.1nm) (see Figs. 11(a) and 11(b), respectively). For each case, we have computed the reflectivity for wavelengths spanning the range 635–670 nm and angles between 1.5° and 1.5° appealing to the HOPS scheme described in Section 3 with N=5 terms. Figure 12 displays the results for each profile that are obtained from direct summation of the Taylor series (21) showing that, rather interestingly, convergence is attained throughout for the purely sinusoidal structure but not for a (slight) perturbation that includes a second harmonic (in fact, the Taylor series diverges in this case; see Fig. 14 below). As shown in Fig. 13, on the other hand, this latter behavior can be overcome with the use of Padé approximants of order [2/2], even though, as we have alluded to before, this uses precisely the same information as that encoded in the first five terms of the Taylor series.

 

Fig. 11. (a) A purely sinusoidal profile with amplitude a=5nm; and (b) a modulated profile with a second harmonic component as in Eq. (29) (a=5nm, b=2nm).

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Fig. 12. Reflectivities as a function of incidence angle and wavelength for: (a) the purely sinusoidal profile in Fig. 11(a); and (b) for the modulated profile Fig. 11(b). In both cases the simulations correspond to the HOPS approach using a truncated Taylor series of order N=5. The white area in the middle of the latter figure (corresponding to values that exceed the maximum in the plotted scale) indicates the region where the five-term Taylor series is already inaccurate.

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Fig. 13. Reflectivities as a function of incidence angle and wavelength for: (a) the purely sinusoidal profile in Fig. 11(a); and (b) a modulated profile with a second harmonic component [Eq. (29)] as depicted in Fig. 11(b). For the latter, the line λ=λ¯ marks the approximate location of the center of the band gap as predicted in Eq. (3b) in [96]. In both cases the simulations correspond to the HOPS approach using [2/2] Padé approximants (N=5).

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Fig. 14. (a) Numerical convergence analysis (and best linear fit) for the examples in Figs. 12(a) and 13(a); and (b) same for those in Figs. 12(b) and 13(b). Logarithm of the relative maximum error in the range of wavelengths between 635 and 670 nm, and angles between 1.5° and 1.5°. Note the slow convergence of the Taylor series in (a), and its divergence in (b).

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As in [41] we find that for a purely sinusoidal structure the intersection of the SPR generated by the orders ±1 takes place at normal incidence, with little interaction among them as evidenced by the absence of a gap; see Fig. 13(a). In contrast, and again as in [41], Fig. 13(b) shows that a small additive perturbation in the second order harmonic allows for the coupling of these modes in a manner so as to generate a band gap. An approximate formula for the central wavelength λ¯ of this gap is provided in Eq. (3b) in [96] (see also Eq. (5.52) in [41]), based on the neglect of higher order harmonics and the assumption that

|Kb|1,
in the formulation of the scattering problem as provided by the approach in [39]. This results in an expression,
(2πλ¯)2(2πλ0)2[12(Kb)2],
where λ0 denotes the SPP wavelength corresponding to a flat surface at normal incidence [Eq. (28) with r=±1 and θ=0°]. In the case of Fig. 13(b) this evaluates to λ¯653.15nm.

It is worth noting that, in contrast with the sentiment that “… the number of terms required to construct a solution to a given precision increases extremely rapidly as the depth of the grating rises…” [41], the results in Fig. 13 were obtained with only five terms and, thus, in the form of a [2/2] Padé approximant, this approximate solution should be equally amenable to an analytic study as the (truncated) 4×4T-matrix” from [39] as used in [41]. Although we shall defer this analysis to a future publication, we remark here that a numerical investigation of the errors associated with the results in this figure suggests that the maximum relative error is less than 0.1% for Fig. 13(a) and about 3% for Fig. 13(b); a numerical convergence analysis is included in Fig. 14.

Our final example is related to the above, and it corresponds to that in Fig. 8 in [41]. In Fig. 15(a) we display the reflectivity corresponding to a sinusoidal profile, such as that in Fig. 12(a), but with a much deeper height of h=60nm. To accurately deal with these depths we resort to a higher order version of the HOPS method using Padé approximants of order [8/8]. We see that, as a result of the change in height a gap opens in angle rather than in frequency. This transition is further detailed in Fig. 15(b), where the progression of reflectivities is shown from 10 to 60 nm in increments of 10 nm; the estimated maximum relative error for the entirety of the results in this latter figure is approximately 1×105; that is, we estimate that each value of the reflectivity has at least 5 digits of accuracy (and increasing with decreasing height). We note that this figure is largely in qualitative agreement with Fig. 8 in [41], though some differences appear visible. On account of our estimated accuracy, we conjecture that these differences arise as a consequence of a lack of resolution in the results in [41] that arise from the use of the method in [39]. Finally, and for comparison purposes, we also include in Fig. 16 analogous results for profiles, such as that in Fig. 12(b) for varying heights h=1060nm and constant frequency content ratio a/b=2.5 [Eq. (29)]; the estimated maximum relative error in this case is about 3×105.

 

Fig. 15. (a) Reflectivity for a sinusoidal profile [b=0 in Eq. (29)] with period d=634nm and height h=2a=60nm. (b) Reflectivities corresponding to the same profile for varying heights h=1060nm.

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Fig. 16. (a) Reflectivity for a Fourier grating [as in Eq. (29)] with period d=634nm, height h=60nm, and a/b=2.5. (b) Reflectivities corresponding to the same profile for varying heights h=1060nm.

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5. CONCLUSIONS

In this paper we have introduced a rigorous high-order perturbation method (HOPS) to bear upon the area of plasmonics. Here, we have shown that the HOPS method [71], which couples ideas of regular perturbation theory and analytic continuation, is particularly well-suited to applications in this area, which typically relate to small and periodic deviations of flat surfaces (to couple light to SPP) in a regime where the period of the resulting structure is comparable to the wavelength of visible light. We have further explained how previous results and conclusions relating to boundary perturbation methods incorrectly attributed their failure to that of the RH and/or dismissed them as limited in applicability due to an unfounded insistence on summation of the resulting Taylor polynomials. In contrast, we have demonstrated that the limitations of HOPS methods are unrelated to the RH and, moreover, that these are significantly less restrictive for the investigation of SPR than previously assumed, provided that suitable mechanisms of analytic continuation (e.g., Padé approximation) are used.

As with other popular numerical simulation schemes (e.g., the C-method [38] or methods based on frequency domain integral equations [28]) the approach we elaborate upon here resolves the scattering problem at a single, fixed frequency and, thus, its repeated use is necessary to garner the complete spectral response of a given structure. In contrast with other techniques, however, the investigation of the dependence of the reflectivity upon the height of the grating coupler can be obtained through the HOPS discussed herein, with minimal additional computational cost. Moreover, we have explained how the very nature of the formulation allows for FFT-accelerated evaluations. The combination of these factors results in a scheme that can deliver full reflectivity maps at a fraction of the computational expense associated with alternative frequency-domain procedures. Comparisons with time-domain techniques [such as the popular finite-difference time-domain (FDTD) method], on the other hand, are left for future work; however, we anticipate that the extremely fine meshes that are known to be necessary to resolve plasmonic resonances (see e.g., [97]) and the extended time intervals that would allow for accurate reconstruction of the spectral response will render these uncompetitive (and, at times, simply impracticable) when compared to the schemes discussed above.

The developments presented here suggest a number of avenues for further investigation. In particular, as we mentioned, further quantitative comparison with and verification against state-of-the-art (e.g., efficient solvers based on surface integral equations) and/or popular (e.g., FDTD) alternative techniques will be forthcoming. In addition, the specifics related to modeling of SPR supports the idea of a search for adaptive procedures (e.g., ones that may take advantage of their “localized” nature in parametric space; see, e.g., Figs. 9 and 10) that may further lower the overall computational cost. This is particularly relevant in connection with an ultimate goal of incorporating mathematical optimization techniques that will allow for rapid virtual prototyping, particularly in the context of biosensing. In this connection, and due to the simplicity, accuracy and efficiency of the approach discussed here, we fully expect that thoughtful parallelization strategies (on CPUs and, particularly, on GPUs) will enable the development of a real-time simulator for use by practitioners. Finally, as we have explained, the extension of the present developments to allow for the treatment of fully three-dimensional geometries is rather straightforward [72]; slightly more complex, though still extremely efficient, is that to multi-layer structures (see e.g., [98]). The implications of both of these extensions in the study of SPR will be the subject of future work.

ACKNOWLEDGMENTS

This work was supported in part by the National Science Foundation through grant number DMR-0941537. TWJ and SHO gratefully acknowledge support from the National Science Foundation through grant number CBET-1067681. The authors wish to thank the anonymous referees for their thoughtful and detailed comments.

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70. O. P. Bruno and F. Reitich, “Numerical solution of diffraction problems: a method of variation of boundaries,” J. Opt. Soc. Am. A 10, 1168–1175 (1993). [CrossRef]  

71. O. P. Bruno and F. Reitich, “Numerical solution of diffraction problems: a method of variation of boundaries. II. Finitely conducting gratings, Padé approximants, and singularities,” J. Opt. Soc. Am. A 10, 2307–2316 (1993). [CrossRef]  

72. O. P. Bruno and F. Reitich, “Numerical solution of diffraction problems: a method of variation of boundaries. III. Doubly periodic gratings,” J. Opt. Soc. Am. A 10, 2551–2562 (1993). [CrossRef]  

73. O. Bruno and F. Reitich, “High-order boundary perturbation methods,” in Mathematical Modeling in Optical Science, G. Bao, L. Cowsar, and W. Masters, eds., Vol. 22 of Frontiers in Applied Mathematics (SIAM, 2001), Chap. 3, pp. 71–109.

74. L. Kazandjian, “A discussion of the properties of the Rayleigh perturbative solution in diffraction theory,” Wave Motion 42, 169–176 (2005). [CrossRef]  

75. J.-J. Greffet, “Introduction to surface plasmon theory,” in Plasmonics: From Basics to Advanced Topics, S. Enoch and N. Bonod, eds., Vol. 167 of Springer Series in Optical Sciences (Springer, 2012), Chap. 4, pp. 105–148.

76. D. Maystre, “Theory of Wood’s anomalies,” in Plasmonics: From Basics to Advanced Topics, S. Enoch and N. Bonod, eds., Vol. 167 of Springer Series in Optical Sciences (Springer, 2012), Chap. 2, pp. 39–83.

77. E. Kretschmann and H. Raether, “Radiative decay of non radiative surface plasmons excited by light,” Z. Naturforsch. A 23, 2135–2136 (1968).

78. A. Otto, “Excitation of nonradiative surface plasma waves in silver by the method of frustrated total reflection,” Z. Phys. 216, 398–410 (1968). [CrossRef]  

79. B. Hecht, H. Bielefeldt, L. Novotny, Y. Inouye, and D. W. Pohl, “Local excitation, scattering, and interference of surface plasmons,” Phys. Rev. Lett. 77, 1889–1892 (1996). [CrossRef]  

80. H. Ditlbacher, J. R. Krenn, N. Felidj, B. Lamprecht, G. Schider, M. Salerno, A. Leitner, and F. R. Aussenegg, “Fluorescence imaging of surface plasmon fields,” Appl. Phys. Lett. 80, 404–406 (2002). [CrossRef]  

81. R. H. Ritchie, E. T. Arakawa, J. J. Cowan, and R. N. Hamm, “Surface-plasmon resonance effect in grating diffraction,” Phys. Rev. Lett. 21, 1530–1533 (1968). [CrossRef]  

82. J. Uretsky, “The scattering of plane waves from periodic surfaces,” Ann. Phys. 33, 400–427 (1965). [CrossRef]  

83. G. A. Baker and P. Graves-Morris, Padé Approximants, 2nd ed., Vol. 59 of Encyclopedia of Mathematics and its Applications (Cambridge University, 1996).

84. A. Rakic, A. Djurišic, J. Elazar, and M. Majewski, “Optical properties of metallic films for vertical-cavity optoelectronic devices,” Appl. Opt. 37, 5271–5283 (1998). [CrossRef]  

85. Z. Chen, I. R. Hooper, and J. R. Sambles, “Low dispersion surface plasmon–polaritons on deep silver gratings,” J. Mod. Opt. 53, 1569–1576 (2006). [CrossRef]  

86. D. J. Nash and J. R. Sambles, “Surface plasmon–polariton study of the optical dielectric function of silver,” J. Mod. Opt. 43, 81–91 (1996).

87. P. B. Johnson and R. W. Christy, “Optical constants of the noble metals,” Phys. Rev. B 6, 4370–4379 (1972). [CrossRef]  

88. D. W. Lynch and W. R. Hunter, “Comments on the optical constants of metals and an introduction to the data for several metals,” in Handbook of Optical Constants of Solids, E. D. Palik, ed. (Academic, 1985), Vol. 1, pp. 275–367.

89. J. Homola, Surface Plasmon Resonance Based Sensors, Vol. 4 of Springer Series on Chemical Sensors and Biosensors (Springer, 2006).

90. Z. Chen, “Grating coupled surface plasmons in metallic structures,” Ph.D. dissertation (University of Exeter, 2007).

91. T. López-Rios, D. Mendoza, F. J. García-Vidal, J. Sánchez-Dehesa, and B. Pannetier, “Surface shape resonances in lamellar metallic gratings,” Phys. Rev. Lett. 81, 665–668 (1998). [CrossRef]  

92. F. García-Vidal, J. Sánchez-Dehesa, A. Dechelette, E. Bustarret, T. López-Rios, T. Fournier, and B. Pannetier, “Localized surface plasmons in lamellar metallic gratings,” J. Lightwave Technol. 17, 2191–2195 (1999). [CrossRef]  

93. I. R. Hooper and J. R. Sambles, “Dispersion of surface plasmon polaritons on short-pitch metal gratings,” Phys. Rev. B 65, 165432 (2002). [CrossRef]  

94. E. P. Da Silva, G. A. Farias, and A. A. Maradudin, “Analysis of three theories of scattering of electromagnetic radiation by gratings,” J. Opt. Soc. Am. A 4, 2022–2024 (1987). [CrossRef]  

95. A. J. Jerri, ed., Advances in the Gibbs Phenomenon (Σ Sampling, 2011).

96. W. L. Barnes, T. W. Preist, S. C. Kitson, J. R. Sambles, N. P. K. Cotter, and D. J. Nash, “Photonic gaps in the dispersion of surface plasmons on gratings,” Phys. Rev. B 51, 11164–11167 (1995). [CrossRef]  

97. A. Taflove and S. Hagness, Computational Electrodynamics: The Finite-Difference Time-Domain method (Artech House, 2005).

98. A. Malcolm and D. P. Nicholls, “A field expansions method for scattering by periodic multilayered media,” J. Acoust. Soc. Am. 129, 1783–1793 (2011). [CrossRef]  

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  85. Z. Chen, I. R. Hooper, and J. R. Sambles, “Low dispersion surface plasmon–polaritons on deep silver gratings,” J. Mod. Opt. 53, 1569–1576 (2006).
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  89. J. Homola, Surface Plasmon Resonance Based Sensors, Vol. 4 of Springer Series on Chemical Sensors and Biosensors (Springer, 2006).
  90. Z. Chen, “Grating coupled surface plasmons in metallic structures,” Ph.D. dissertation (University of Exeter, 2007).
  91. T. López-Rios, D. Mendoza, F. J. García-Vidal, J. Sánchez-Dehesa, and B. Pannetier, “Surface shape resonances in lamellar metallic gratings,” Phys. Rev. Lett. 81, 665–668 (1998).
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  92. F. García-Vidal, J. Sánchez-Dehesa, A. Dechelette, E. Bustarret, T. López-Rios, T. Fournier, and B. Pannetier, “Localized surface plasmons in lamellar metallic gratings,” J. Lightwave Technol. 17, 2191–2195 (1999).
    [CrossRef]
  93. I. R. Hooper and J. R. Sambles, “Dispersion of surface plasmon polaritons on short-pitch metal gratings,” Phys. Rev. B 65, 165432 (2002).
    [CrossRef]
  94. E. P. Da Silva, G. A. Farias, and A. A. Maradudin, “Analysis of three theories of scattering of electromagnetic radiation by gratings,” J. Opt. Soc. Am. A 4, 2022–2024 (1987).
    [CrossRef]
  95. A. J. Jerri, ed., Advances in the Gibbs Phenomenon (Σ Sampling, 2011).
  96. W. L. Barnes, T. W. Preist, S. C. Kitson, J. R. Sambles, N. P. K. Cotter, and D. J. Nash, “Photonic gaps in the dispersion of surface plasmons on gratings,” Phys. Rev. B 51, 11164–11167 (1995).
    [CrossRef]
  97. A. Taflove and S. Hagness, Computational Electrodynamics: The Finite-Difference Time-Domain method (Artech House, 2005).
  98. A. Malcolm and D. P. Nicholls, “A field expansions method for scattering by periodic multilayered media,” J. Acoust. Soc. Am. 129, 1783–1793 (2011).
    [CrossRef]

2012 (3)

T. W. Johnson, Z. J. Lapin, R. Beams, N. C. Lindquist, S. G. Rodrigo, L. Novotny, and S.-H. Oh, “Highly reproducible near-field optical imaging with sub-20-nm resolution based on template-stripped gold pyramids,” ACS Nano 6, 9168–9174 (2012).
[CrossRef]

N. C. Lindquist, P. Nagpal, K. M. McPeak, D. J. Norris, and S.-H. Oh, “Engineering metallic nanostructures for plasmonics and nanophotonics,” Rep. Prog. Phys. 75, 036501 (2012).
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2011 (10)

J. M. Taboada, J. Rivero, F. Obelleiro, M. G. Araújo, and L. Landesa, “Method-of-moments formulation for the analysis of plasmonic nano-optical antennas,” J. Opt. Soc. Am. A 28, 1341–1348 (2011).
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O. Tsilipakos, A. Pitilakis, A. Tasolamprou, T. Yioultsis, and E. Kriezis, “Computational techniques for the analysis and design of dielectric-loaded plasmonic circuitry,” Opt. Quantum Electron. 42, 541–555 (2011).
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N. C. Lindquist, T. W. Johnson, D. J. Norris, and S.-H. Oh, “Monolithic integration of continuously tunable plasmonic nanostructures,” Nano Lett. 11, 3526–3530 (2011).
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A. Shahmansouri and B. Rashidian, “Comprehensive three-dimensional split-field finite-difference time-domain method for analysis of periodic plasmonic nanostructures: near- and far-field formulation,” J. Opt. Soc. Am. B 28, 2690–2700 (2011).
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A. Trügler, J.-C. Tinguely, J. R. Krenn, A. Hohenau, and U. Hohenester, “Influence of surface roughness on the optical properties of plasmonic nanoparticles,” Phys. Rev. B 83, 081412(R) (2011).
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2010 (5)

H.-Y. Xie, M.-Y. Ng, and Y.-C. Chang, “Analytical solutions to light scattering by plasmonic nanoparticles with nearly spherical shape and nonlocal effect,” J. Opt. Soc. Am. A 27, 2411–2422 (2010).
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2009 (11)

H. Kurkcu and F. Reitich, “Stable and efficient evaluation of periodized Green’s functions for the Helmholtz equation at high frequencies,” J. Comput. Phys. 228, 75–95 (2009).
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J. Hoffmann, C. Hafner, P. Leidenberger, J. Hesselbarth, and S. Burger, “Comparison of electromagnetic field solvers for the 3D analysis of plasmonic nano antennas,” Proc. SPIE  7390, 73900J (2009).

J. Smajic, C. Hafner, L. Raguin, K. Tavzarashvili, and M. Mishrikey, “Comparison of numerical methods for the analysis of plasmonic structures,” J. Comput. Theor. Nanosci. 6, 763–774 (2009).
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J. Bischoff, “Prospects and limits of the Rayleigh Fourier approach for diffraction modelling in scatterometry and lithography,” Proc. SPIE  7390, 73901E (2009).

2008 (4)

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J. Homola, “Surface plasmon resonance sensors for detection of chemical and biological species,” Chem. Rev. 108, 462–493 (2008).
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J. M. Montgomery, T.-W. Lee, and S. K. Gray, “Theory and modeling of light interactions with metallic nanostructures,” J. Phys. Condens. Matter 20, 323201 (2008).
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Y. Otani and N. Nishimura, “A periodic FMM for Maxwell’s equations in 3D and its applications to problems related to photonic crystals,” J. Comput. Phys. 227, 4630–4652 (2008).
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2007 (3)

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D. Barchiesi, B. Guizal, and T. Grosges, “Accuracy of local field enhancement models: toward predictive models?” Appl. Phys. B 84, 55–60 (2006).
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Z. Chen, I. R. Hooper, and J. R. Sambles, “Low dispersion surface plasmon–polaritons on deep silver gratings,” J. Mod. Opt. 53, 1569–1576 (2006).
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2005 (3)

2004 (2)

C.-A. Guérin and A. Sentenac, “Second-order perturbation theory for scattering from heterogeneous rough surfaces,” J. Opt. Soc. Am. A 21, 1251–1260 (2004).
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T. Elfouhaily and C. Guérin, “A critical survey of approximate scattering wave theories from random rough surfaces,” Wave Random Media 14, R1–R40 (2004).
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2003 (1)

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H. Ditlbacher, J. R. Krenn, N. Felidj, B. Lamprecht, G. Schider, M. Salerno, A. Leitner, and F. R. Aussenegg, “Fluorescence imaging of surface plasmon fields,” Appl. Phys. Lett. 80, 404–406 (2002).
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M. Kreiter, S. Mittler, W. Knoll, and J. R. Sambles, “Surface plasmon-related resonances on deep and asymmetric gold gratings,” Phys. Rev. B 65, 125415 (2002).
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I. R. Hooper and J. R. Sambles, “Dispersion of surface plasmon polaritons on short-pitch metal gratings,” Phys. Rev. B 65, 165432 (2002).
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2001 (3)

1999 (3)

1998 (3)

1997 (1)

1996 (4)

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1995 (2)

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1982 (1)

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1981 (1)

M. G. Moharam and T. K. Gaylord, “Rigorous coupled-wave analysis of planar-grating diffraction,” J. Opt. Soc. Am. A 71, 811–818 (1981).
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1980 (1)

J. Chandezon, G. Raoult, and D. Maystre, “A new theoretical method for diffraction gratings and its numerical application,” J. Opt. 11, 235–241 (1980).
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1972 (1)

P. B. Johnson and R. W. Christy, “Optical constants of the noble metals,” Phys. Rev. B 6, 4370–4379 (1972).
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W. L. Barnes, T. W. Preist, S. C. Kitson, and J. R. Sambles, “Physical origin of photonic energy gaps in the propagation of surface plasmons on gratings,” Phys. Rev. B 54, 6227–6244 (1996).
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J. Bischoff and K. Hehl, “Perturbation approach applied to modal diffraction methods,” J. Opt. Soc. Am. A 28, 859–867 (2011).
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C.-C. Chao, S.-H. Tu, C.-M. Wang, H.-I. Huang, C.-C. Chen, and J.-Y. Chang, “Impedance-matching surface plasmon absorber for FDTD simulations,” Plasmonics 5, 51–55 (2010).
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J. Chandezon, M. Dupuis, G. Cornet, and D. Maystre, “Multicoated gratings: a differential formalism applicable in the entire optical region,” J. Opt. Soc. Am. A 72, 839–846 (1982).
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H. Ditlbacher, J. R. Krenn, N. Felidj, B. Lamprecht, G. Schider, M. Salerno, A. Leitner, and F. R. Aussenegg, “Fluorescence imaging of surface plasmon fields,” Appl. Phys. Lett. 80, 404–406 (2002).
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J. Chandezon, M. Dupuis, G. Cornet, and D. Maystre, “Multicoated gratings: a differential formalism applicable in the entire optical region,” J. Opt. Soc. Am. A 72, 839–846 (1982).
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H. Ditlbacher, J. R. Krenn, N. Felidj, B. Lamprecht, G. Schider, M. Salerno, A. Leitner, and F. R. Aussenegg, “Fluorescence imaging of surface plasmon fields,” Appl. Phys. Lett. 80, 404–406 (2002).
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Figures (16)

Fig. 1.
Fig. 1.

Generic grating configuration for SPR coupling.

Fig. 2.
Fig. 2.

(a) (Normalized) reflectivity [“NR” in Eq. (16)] map for a sinusoidal silver profile with period d=650nm at normal incidence in TM polarization; (b) poles (open markers) and zeros (filled markers) of (the [32/32] Padé approximant of) the coefficient B0(h) for complex h and different values of the wavelength of radiation λ (660λ700). Note that no poles are present in a complex neighborhood of the real line (shaded area). The zeros indicated with an arrow correspond to close-to-zero reflectivity, indicating the presence of a plasmon resonance.

Fig. 3.
Fig. 3.

(a) Square of the absolute value |B1+|2 of the 1st order Rayleigh coefficient B1+ in Eq. (10) (“reflectivity at order 1”) for a sinusoidal silver grating of pitch d=400nm and height h=4nm, illuminated with a plane wave of energy 2.26 eV, corresponding to a wavelength λ=548.60nm, as a function of incidence angle; (b) same as in (a) but with height h=6nm; (c) same as in (a) but with height h=10nm.

Fig. 4.
Fig. 4.

Logarithm (base 10) of the maximum relative error in |B1+|2 (circles) and best linear fit (dashed line) corresponding to incidences in the interval [19.2°,19.4°] for the case in Fig. 3(c) (h=10nm). The optical constants used in the simulations are from [87].

Fig. 5.
Fig. 5.

(a) Configuration corresponding to a deep silver grating of height 124 nm and period 258 nm illuminated with a plane wave with an incidence angle of 30°; (b) reflectivity at order 0, |B0|2 [Eq. (10)], as a function of wavelength for a polynomial permittivity [85] as computed from a [32/32] Padé approximant. The dashed vertical line marks the wavelength below which the 1st Rayleigh mode becomes propagating.

Fig. 6.
Fig. 6.

(a) Zeroth order reflectivity as a function of height h for the example in Fig. 5 at a wavelength λ=410nm; the solid lines display the values of the Padé approximations corresponding to each value of N for which the Taylor partial sums are shown. The vertical line at h=hp depicts the absolute value of the closest (noncanceling) pole to the origin as approximated by the corresponding pole of the [32/32] Padé approximant (N=65). (b) Poles and zeros in the complex h plane of the [32/32] Padé approximant for the configuration in (a); the circles and crosses (square and stars) mark the location of the zeros and poles, respectively, that do (do not) cancel out. (c) Same as (a) but for a wavelength λ=495nm. (d) Same as (b) for the configuration in (c).

Fig. 7.
Fig. 7.

(Total) Reflectivity for an elliptical grating configuration, as shown in (a) (period d=630nm, linewidth w=400nm), for different incidence angles. The vertical lines mark the wavelength below which the 1st order becomes propagating.

Fig. 8.
Fig. 8.

Left- (horizontal lines) and right- (curves) hand sides of Eq. (28). Their intersections, marked by vertical lines, define the resonant wavelengths that would be excited for a silver interface of zero height and period d=630, as depicted in (a), under the dispersion derived from (cubic-spline) interpolation of the results in [88].

Fig. 9.
Fig. 9.

Results for a sinusoidal profile of period d=258nm illuminated with an incidence angle θ=30°. (a) Total reflectivity as a function of height and wavelength; the thick dashed vertical line marks the wavelength below which the 1st order becomes propagating. (b) Spatial distribution of the (transverse) reflected magnetic field, |u+,scat|=|Hz+,scat| [Eq. (3)], corresponding to the intersection of the dotted–dashed lines in (a): height h=20nm and wavelength λ=410nm.

Fig. 10.
Fig. 10.

Same as Fig. 9 for the configuration of Fig. 7(a) with incidence angle θ=1° [Fig. 7(d)]. (a) Reflectivity map; (b) spatial (transverse) magnetic field distribution corresponding to the intersection of the dotted–dashed lines in (a): height h=30nm and wavelength λ=640nm.

Fig. 11.
Fig. 11.

(a) A purely sinusoidal profile with amplitude a=5nm; and (b) a modulated profile with a second harmonic component as in Eq. (29) (a=5nm, b=2nm).

Fig. 12.
Fig. 12.

Reflectivities as a function of incidence angle and wavelength for: (a) the purely sinusoidal profile in Fig. 11(a); and (b) for the modulated profile Fig. 11(b). In both cases the simulations correspond to the HOPS approach using a truncated Taylor series of order N=5. The white area in the middle of the latter figure (corresponding to values that exceed the maximum in the plotted scale) indicates the region where the five-term Taylor series is already inaccurate.

Fig. 13.
Fig. 13.

Reflectivities as a function of incidence angle and wavelength for: (a) the purely sinusoidal profile in Fig. 11(a); and (b) a modulated profile with a second harmonic component [Eq. (29)] as depicted in Fig. 11(b). For the latter, the line λ=λ¯ marks the approximate location of the center of the band gap as predicted in Eq. (3b) in [96]. In both cases the simulations correspond to the HOPS approach using [2/2] Padé approximants (N=5).

Fig. 14.
Fig. 14.

(a) Numerical convergence analysis (and best linear fit) for the examples in Figs. 12(a) and 13(a); and (b) same for those in Figs. 12(b) and 13(b). Logarithm of the relative maximum error in the range of wavelengths between 635 and 670 nm, and angles between 1.5° and 1.5°. Note the slow convergence of the Taylor series in (a), and its divergence in (b).

Fig. 15.
Fig. 15.

(a) Reflectivity for a sinusoidal profile [b=0 in Eq. (29)] with period d=634nm and height h=2a=60nm. (b) Reflectivities corresponding to the same profile for varying heights h=1060nm.

Fig. 16.
Fig. 16.

(a) Reflectivity for a Fourier grating [as in Eq. (29)] with period d=634nm, height h=60nm, and a/b=2.5. (b) Reflectivities corresponding to the same profile for varying heights h=1060nm.

Equations (32)

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y=f(x),f(x+d)=f(x),
(E⃗(x,t),H⃗(x,t))=eiωt(E(x),H(x)),x=(x,y,z),
u={Ezfor TE polarization,Hzfor TM polarization,
Δu±,scat+(k±)2u±,scat=0inD±,
{u+,scatu,scat=uinconS,u+,scatnC02u,scatn=uincnonS,
C0={1for   TEk+/kfor   TM.
uinc(x)=u(x,y)=eiαxiβy,
uinc(x+d,y)=eiαdeiαxiβy,
u±,scat(x+d,y)=eiαdu±,scat(x,y),
u±,scat(x,y)=r=Br±eiαrx±iβr±y,valid on±y>maxxR{±f(x)},
αr=α+rK,K=2πd,βr±=(k±)2(αr)2,ris an integer,and Im(βr±)0.
kx,SPP=±k0ε+εε++ε,
|Re(kx,SPP)|>k+,
Re(kx,SPP)αr,
RrUβr+β|Br+|21,
NRrUβr+β|Br+|2|B0+,flat|21,
NR=|B0+|2|B0+,flat|21.
maxxR(f(x))minxR(f(x))=1,
Δxu±,scat(x;h)+(k±)2u±,scat(x;h)=0inDh±,
{u+,scat(·;h)u,scat(·;h)=uinc(·;h),u+,scatn(·;h)C02u,scatn(·;h)=uincn(·;h),
u±,scat(x,y;h)=r=Br±(h)eiαrx±iβr±y,valid on±y>maxxR{±hf(x)}.
Br±(h)=n=0dn,r±hn,
dn,r+dn,r=(iβ)nCn,rk=0n1q=max[kF,r(nk)F]min[kF,r+(nk)F]Cnk,rq[(iβq+)nkdk,q+(iβq)nkdk,q],
iβr+dn,r++C02iβrdn,r=Cn,r(iβ)n1[(iα)(iKr)(iβ)2]+k=0n1q=max[kF,r(nk)F]min[kF,r+(nk)F]Cnk,rq{[iK(rq)](iαq)×[(iβq+)nk1dk,q+C02(iβq)nkdk,q][(iβq+)nk1dk,q+C02(iβq)nkdk,q]},
f(x)=p=FFC1,pei2πpx/d,
f(x)!=p=FFC,pei2πpx/d.
y=h2sin(2πdx),
n=0.060+3.586i.
r=[±ε+εε++εsin(θ)]dλ,
f(x)=asin(Kx)+bsin(2Kx),(K=2πd),
|Kb|1,
(2πλ¯)2(2πλ0)2[12(Kb)2],

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