\documentclass[aps,prl,reprint,longbibliography,floatfix]{revtex4-1} \usepackage[utf8]{inputenc} \usepackage{amsmath,graphicx,upgreek,amssymb,xcolor} \usepackage[colorlinks=true,urlcolor=purple,citecolor=purple,filecolor=purple,linkcolor=purple]{hyperref} \usepackage[english]{babel} \definecolor{mathc1}{html}{5e81b5} \definecolor{mathc2}{html}{e19c24} \definecolor{mathc3}{html}{8fb032} \definecolor{mathc4}{html}{eb6235} \makeatletter % A change to a babel macro -- Don't ask! \def\bbl@set@language#1{% \edef\languagename{% \ifnum\escapechar=\expandafter`\string#1\@empty \else\string#1\@empty\fi}% %%%% ADDITION \@ifundefined{babel@language@alias@\languagename}{}{% \edef\languagename{\@nameuse{babel@language@alias@\languagename}}% }% %%%% END ADDITION \select@language{\languagename}% \expandafter\ifx\csname date\languagename\endcsname\relax\else \if@filesw \protected@write\@auxout{}{\string\select@language{\languagename}}% \bbl@for\bbl@tempa\BabelContentsFiles{% \addtocontents{\bbl@tempa}{\xstring\select@language{\languagename}}}% \bbl@usehooks{write}{}% \fi \fi} % The user interface \newcommand{\DeclareLanguageAlias}[2]{% \global\@namedef{babel@language@alias@#1}{#2}% } \makeatother \DeclareLanguageAlias{en}{english} \newcommand{\brad}[1]{{\color{red} #1}} % Our mysterious boy \def\urusi{URu$_{\text2}$Si$_{\text2}$} \def\e{{\text{\textsc{elastic}}}} % "elastic" \def\i{{\text{\textsc{int}}}} % "interaction" \def\Dfh{D$_{\text{4h}}$} % Irreducible representations (use in math mode) \def\Aog{{\text A_{\text{1g}}}} \def\Atg{{\text A_{\text{2g}}}} \def\Bog{{\text B_{\text{1g}}}} \def\Btg{{\text B_{\text{2g}}}} \def\Eg {{\text E_{\text g}}} \def\Aou{{\text A_{\text{1u}}}} \def\Atu{{\text A_{\text{2u}}}} \def\Bou{{\text B_{\text{1u}}}} \def\Btu{{\text B_{\text{2u}}}} \def\Eu {{\text E_{\text u}}} % Variables to represent some representation \def\X{\text X} \def\Y{\text Y} % Units \def\J{\text J} \def\m{\text m} \def\K{\text K} \def\GPa{\text{GPa}} \def\A{\text{\r A}} % Other \def\op{\textsc{op}} % order parameter \def\ho{\textsc{ho}} % hidden order \def\rus{\textsc{rus}} % resonant ultrasound spectroscopy \def\Rus{\textsc{Rus}} % Resonant ultrasound spectroscopy \def\afm{\textsc{afm}} % antiferromagnetism \def\recip{{\{-1\}}} % functional reciprocal \begin{document} \title{Elastic properties of hidden order in \urusi\ are reproduced by staggered nematic order} \author{Jaron Kent-Dobias} \author{Michael Matty} \author{Brad Ramshaw} \affiliation{ Laboratory of Atomic \& Solid State Physics, Cornell University, Ithaca, NY, USA } \date\today \begin{abstract} We develop a phenomenological mean field theory of the hidden order phase in \urusi\ as a ``staggered nematic" order. Several experimental features are reproduced when the order parameter is a nematic of the $\Bog$ representation, staggered along the c-axis: the topology of the temperature--pressure phase diagram, the response of the elastic modulus $(C_{11}-C_{12})/2$ above the hidden-order transition at zero pressure, and orthorhombic symmetry breaking in the high-pressure antiferromagnetic phase. In this scenario, hidden order is characterized by broken rotational symmetry that is modulated along the $c$-axis, the primary order of the high-pressure phase is an unmodulated nematic state, and the triple point joining those two phases with the high-temperature paramagnetic phase is a Lifshitz point. \end{abstract} \maketitle \emph{Introduction.} \urusi\ is a paradigmatic example of a material with an ordered state whose broken symmetry remains unknown. This state, known as \emph{hidden order} (\ho), sets the stage for unconventional superconductivity that emerges at even lower temperatures. At sufficiently large hydrostatic pressures, both superconductivity and \ho\ give way to local moment antiferromagnetism (\afm) \cite{hassinger_temperature-pressure_2008}. Despite over thirty years of effort, the symmetry of the \ho\ state remains unknown, and modern theories \cite{kambe_odd-parity_2018, haule_arrested_2009, kusunose_hidden_2011, kung_chirality_2015, cricchio_itinerant_2009, ohkawa_quadrupole_1999, santini_crystal_1994, kiss_group_2005, harima_why_2010, thalmeier_signatures_2011, tonegawa_cyclotron_2012, rau_hidden_2012, riggs_evidence_2015, hoshino_resolution_2013, ikeda_theory_1998, chandra_hastatic_2013, harrison_hidden_nodate, ikeda_emergent_2012} propose a variety of possibilities. Our work here seeks to unify two experimental observations: one, the $\Bog$ ``nematic" elastic susceptibility $(C_{11}-C_{12})/2$ softens anomalously from room temperature down to T$_{\mathrm{HO}}=17.5~$ K \brad{find old citations for this data}; and two, a $\Bog$ nematic distortion is observed by x-ray scattering under sufficient pressure to destroy the \ho\ state \cite{choi_pressure-induced_2018}. Recent \emph{resonant ultrasound spectroscopy} (\rus) measurements examined the thermodynamic discontinuities in the elastic moduli at T$_{\mathrm{HO}}$ \cite{ghosh_single-component_nodate}. The observation of discontinues only in compressional, or $\Aog$, elastic moduli requires that the point-group representation of \ho\ is one-dimensional. This rules out a large number of order parameter candidates \brad{cite those ruled out} in a model-free way, but still leaves the microscopic nature of \ho~ undecided. Recent X-ray experiments discovered rotational symmetry breaking in \urusi\ under pressure \cite{choi_pressure-induced_2018}. Above 0.13--0.5 $\GPa$ (depending on temperature), \urusi\ undergoes a $\Bog$ nematic distortion. While it is still unclear as to whether this is a true thermodynamic phase transition, it may be related to the anomalous softening of the $\Bog$ elastic modulus---$(c_{11}-c_{12})/2$ in Voigt notation---that occurs over a broad temperature range at zero-pressure \brad{cite old ultrasound}. Motivated by these results, hinting at a $\Bog$ strain susceptibility associated with the \ho\ state, we construct a phenomenological mean field theory for an arbitrary \op\ coupled to strain, and the determine the effect of its phase transitions on the elastic response in different symmetry channels. We find that only one \op\ symmetry reproduces the anomalous $\Bog$ elastic modulus, which softens in a Curie-Weiss like manner from room temperature, but which cusps at T$_{\mathrm{HO}}$. That theory associates \ho\ with a $\Bog$ \op\ \emph{modulated along the $c$- axis}, the \afm\ state with uniform $\Bog$ order, and the triple point between them with a Lifshitz point. Besides the agreement with ultrasound data across a broad temperature range, the theory predicts uniform $\Bog$ strain at high pressure---the same distortion which was recently seen in x-ray scattering experiments \cite{choi_pressure-induced_2018}. This theory strongly motivates future ultrasound experiments under pressure approaching the Lifshitz point, which should find that the $(c_{11}-c_{12})/2$ diverges once the uniform $\Bog$ strain sets in. \emph{Model.} The point group of \urusi\ is \Dfh, and any coarse-grained theory must locally respect this symmetry in the high-temperature phase. Our phenomenological free energy density contains three parts: the elastic free energy, the \op, and the interaction between strain and \op. The most general quadratic free energy of the strain $\epsilon$ is $f_\e=C^0_{ijkl}\epsilon_{ij}\epsilon_{kl}$. Linear combinations of the six independent components of strain form five irreducible components of strain as \begin{equation} \begin{aligned} & \epsilon_{\Aog,1}=\epsilon_{11}+\epsilon_{22} \hspace{0.15\columnwidth} && \epsilon_\Bog=\epsilon_{11}-\epsilon_{22} \\ & \epsilon_{\Aog,2}=\epsilon_{33} && \epsilon_\Btg=2\epsilon_{12} \\ & \epsilon_\Eg=2\{\epsilon_{11},\epsilon_{22}\}. \end{aligned} \label{eq:strain-components} \end{equation} All quadratic combinations of these irreducible strains that transform like $\Aog$ are included in the free energy, \begin{equation} f_\e=\frac12\sum_\X C^0_{\X,ij}\epsilon_{\X,i}\epsilon_{\X,j}, \end{equation} where the sum is over irreducible representations of the point group and the bare elastic moduli $C^0_\X$ are \begin{equation} \begin{aligned} & C^0_{\Aog,11}=\tfrac12(C^0_{1111}+C^0_{1122}) && C^0_{\Bog}=\tfrac12(C^0_{1111}-C^0_{1122}) \\ & C^0_{\Aog,22}=C^0_{3333} && C^0_{\Btg}=C^0_{1212} \\ & C^0_{\Aog,12}=C^0_{1133} && C^0_{\Eg}=C^0_{1313}. \end{aligned} \end{equation} The interaction between strain and an \op\ $\eta$ depends on the point group representation of $\eta$. If this representation is $\X$, the most general coupling to linear order is \begin{equation} f_\i=-b^{(i)}\epsilon_\X^{(i)}\eta. \end{equation} If there doesn't exist a component of strain that transforms like the representation $\X$ there can be no linear coupling, and the effect of the \op\ condensing at a continuous phase transition is to produce a jump in the $\Aog$ elastic modului if $\eta$ is single-component \cite{luthi_sound_1970, ramshaw_avoided_2015, shekhter_bounding_2013}, and jumps in other elastic moduli if multicompenent \cite{ghosh_single-component_nodate}. Because we are interested in physics that anticipates the phase transition, we will focus our attention on \op s that can produce linear couplings to strain. Looking at the components present in \eqref{eq:strain-components}, this rules out all of the \emph{u}-reps (which are odd under inversion) and the $\Atg$ irrep. If the \op\ transforms like $\Aog$ (e.g. a fluctuation in valence number), odd terms are allowed in its free energy and any transition will be first order and not continuous without fine-tuning. Since the \ho\ phase transition is second-order \brad{cite something}, we will henceforth rule out $\Aog$ \op s as well. For the \op\ representation $\X$ as any of $\Bog$, $\Btg$, or $\Eg$, the most general quadratic free energy density is \begin{equation} \begin{aligned} f_\op=\frac12\big[&r\eta^2+c_\parallel(\nabla_\parallel\eta)^2 +c_\perp(\nabla_\perp\eta)^2 \\ &\qquad\qquad\qquad\quad+D_\perp(\nabla_\perp^2\eta)^2\big]+u\eta^4, \end{aligned} \label{eq:fo} \end{equation} where $\nabla_\parallel=\{\partial_1,\partial_2\}$ transforms like $\Eu$, and $\nabla_\perp=\partial_3$ transforms like $\Atu$. Other quartic terms are allowed---especially many for an $\Eg$ \op---but we have included only those terms necessary for stability when either $r$ or $c_\perp$ become negative. The full free energy functional of $\eta$ and $\epsilon$ is \begin{equation} \begin{aligned} F[\eta,\epsilon] &=F_\op[\eta]+F_\e[\epsilon]+F_\i[\eta,\epsilon] \\ &=\int dx\,(f_\op+f_\e+f_\i). \end{aligned} \end{equation} The only strain relevant to the \op\ at linear coupling is $\epsilon_\X$, which can be traced out of the problem exactly in mean field theory. Extremizing with respect to $\epsilon_\X$, \begin{equation} 0=\frac{\delta F[\eta,\epsilon]}{\delta\epsilon_\X(x)}\bigg|_{\epsilon=\epsilon_\star}=C^0_\X\epsilon^\star_\X(x) -b\eta(x) \end{equation} \textbf{talk more about the functional-ness of these parameters!, also, why are we tracinig out strain?} gives the optimized strain conditional on the \op\ as $\epsilon_\X^\star[\eta](x)=(b/C^0_\X)\eta(x)$ and $\epsilon_\Y^\star[\eta]=0$ for all other $\Y$. Upon substitution into the free energy, the resulting effective free energy $F[\eta,\epsilon_\star[\eta]]$ has a density identical to $f_\op$ with $r\to\tilde r=r-b^2/2C^0_\X$. \begin{figure}[htpb] \includegraphics[width=\columnwidth]{phase_diagram_experiments} \vspace{1em} \includegraphics[width=0.51\columnwidth]{phases_scalar}\hspace{-1.5em} \includegraphics[width=0.51\columnwidth]{phases_vector} \caption{ Phase diagrams for (a) \urusi\ from experiments (neglecting the superconducting phase) \cite{hassinger_temperature-pressure_2008} (b) mean field theory of a one-component ($\Bog$ or $\Btg$) Lifshitz point (c) mean field theory of a two-component ($\Eg$) Lifshitz point. Solid lines denote continuous transitions, while dashed lines denote first order transitions. Later, when we fit the elastic moduli predictions for a $\Bog$ \op\ to data along the zero (atmospheric) pressure line, we will take $\Delta\tilde r=\tilde r-\tilde r_c=a(T-T_c)$. } \label{fig:phases} \end{figure} With the strain traced out, \eqref{eq:fo} describes the theory of a Lifshitz point at $\tilde r=c_\perp=0$ \cite{lifshitz_theory_1942, lifshitz_theory_1942-1}. The properties discused below can all be found in a standard text, e.g.,~\cite{chaikin_principles_2000}. For a one-component \op\ ($\Bog$ or $\Btg$) and positive $c_\parallel$, it is traditional to make the field ansatz $\langle\eta(x)\rangle=\eta_*\cos(q_*x_3)$. For $\tilde r>0$ and $c_\perp>0$, or $\tilde r>c_\perp^2/4D_\perp$ and $c_\perp<0$, the only stable solution is $\eta_*=q_*=0$ and the system is unordered. For $\tilde r<0$ there are free energy minima for $q_*=0$ and $\eta_*^2=-\tilde r/4u$ and this system has uniform order with the \op\ symmetry, e.g., $\Bog$ or $\Btg$. For $c_\perp<0$ and $\tilde r0$, and the modulated phase is now characterized by helical order with $\langle\eta(x)\rangle=\eta_*\{\cos(q_*x_3),\sin(q_*x_3)\}$. The uniform to modulated transition is now continuous. This does not reproduce the physics of \ho, which has a first order transition between \ho\ and \afm, and so we will henceforth neglect the possibility of a multicomponent order parameter. The schematic phase diagrams for this model are shown in Figure~\ref{fig:phases}. \emph{Results.} We will now derive the \emph{effective elastic tensor} $C$ that results from the coupling of strain to the \op. The ultimate result, found in \eqref{eq:elastic.susceptibility}, is that $C_\X$ differs from its bare value $C^0_\X$ only for the symmetry $\X$ of the \op. Moreover, the effective elastic moduli does not vanish at the unordered to modulated transition---as it would if the transition were a $q=0$ structural phase transition---but instead exhibits a \emph{cusp}. To show this, we will first compute the susceptibility of the \op, which will both be demonstrative of how the modulus is calculated and prove useful in expressing the functional form of the modulus. Then we will compute the elastic modulus using techniques from functional calculus. The generalized susceptibility of a single component ($\Bog$ or $\Btg$) \op\ is \begin{equation} \begin{aligned} &\chi^\recip(x,x') =\frac{\delta^2F[\eta,\epsilon_\star[\eta]]}{\delta\eta(x)\delta\eta(x')}\bigg|_{\eta=\langle\eta\rangle} =\big[\tilde r-c_\parallel\nabla_\parallel^2 \\ &\qquad\qquad-c_\perp\nabla_\perp^2+D_\perp\nabla_\perp^4+12u\langle\eta(x)\rangle^2\big] \delta(x-x'), \end{aligned} \label{eq:sus_def} \end{equation} where $\recip$ indicates a \emph{functional reciprocal} in the sense that \begin{equation} \int dx''\,\chi^\recip(x,x'')\chi(x'',x')=\delta(x-x'). \end{equation} Taking the Fourier transform and integrating over $q'$ we have \begin{equation} \chi(q) =\big(\tilde r+c_\parallel q_\parallel^2+c_\perp q_\perp^2+D_\perp q_\perp^4 +12u\sum_{q'}\langle\tilde\eta_{q'}\rangle\langle\tilde\eta_{-q'}\rangle\big)^{-1}. \end{equation} Near the unordered to modulated transition this yields \begin{equation} \begin{aligned} \chi(q) &=\frac1{c_\parallel q_\parallel^2+D_\perp(q_*^2-q_\perp^2)^2 +|\Delta\tilde r|} \\ &=\frac1{D_\perp}\frac{\xi_\perp^4} {1+\xi_\parallel^2q_\parallel^2+\xi_\perp^4(q_*^2-q_\perp^2)^2}, \end{aligned} \label{eq:susceptibility} \end{equation} with $\xi_\perp=(|\Delta\tilde r|/D_\perp)^{-1/4}=\xi_{\perp0}|t|^{-1/4}$ and $\xi_\parallel=(|\Delta\tilde r|/c_\parallel)^{-1/2}=\xi_{\parallel0}|t|^{-1/2}$, where $t=(T-T_c)/T_c$ is the reduced temperature and $\xi_{\perp0}=(D_\perp/aT_c)^{1/4}$ and $\xi_{\parallel0}=(c_\parallel/aT_c)^{1/2}$ are the bare correlation lengths \brad{needs a descriptor like "in and perpendicular to the x-y plane" or something like that}. Notice that the static susceptibility $\chi(0)=(D_\perp q_*^4+|\Delta\tilde r|)^{-1}$ does not diverge at the unordered to modulated transition. Though it anticipates a transition with Curie--Weiss-like divergence at $\Delta\tilde r=-D_\perp q_*^4$, this is cut off with a cusp at $\Delta\tilde r=0$ \brad{this will all be clearer if you remind the reader that this is Tc, or the new renormalized Tc, or whatever it is}. The elastic susceptibility, which corresponds with the reciprocal of the elastic modulus, is given in a similar way to the \op\ susceptibility: we must trace over $\eta$ and take the second variation of the resulting effective free energy functional of $\epsilon$. Extremizing over $\eta$ yields \begin{equation} 0=\frac{\delta F[\eta,\epsilon]}{\delta\eta(x)}\bigg|_{\eta=\eta_\star}= \frac{\delta F_\op[\eta]}{\delta\eta(x)}\bigg|_{\eta=\eta_\star}-b\epsilon_\X(x), \label{eq:implicit.eta} \end{equation} which implicitly gives $\eta_\star[\epsilon]$, the optimized \op\ conditioned on the strain. Since $\eta_\star$ is a functional of $\epsilon_\X$ alone, only the modulus $C_\X$ can be modified from its bare value $C^0_\X$. Though this differential equation for $\eta_*$ cannot be solved explicitly, we can make use of the inverse function theorem. First, denote by $\eta_\star^{-1}[\eta]$ the inverse functional of $\eta_\star$ implied by \eqref{eq:implicit.eta}, which gives the function $\epsilon_\X$ corresponding to each solution of \eqref{eq:implicit.eta} it receives. This we can immediately identify from \eqref{eq:implicit.eta} as $\eta^{-1}_\star[\eta](x)=b^{-1}(\delta F_\op[\eta]/\delta\eta(x))$. Now, we use the inverse function theorem to relate the functional reciprocal of the derivative of $\eta_\star[\epsilon]$ with respect to $\epsilon_\X$ to the derivative of $\eta^{-1}_\star[\eta]$ with respect to $\eta$, yielding \begin{equation} \begin{aligned} \bigg(\frac{\delta\eta_\star[\epsilon](x)}{\delta\epsilon_\X(x')}\bigg)^\recip &=\frac{\delta\eta_\star^{-1}[\eta](x)}{\delta\eta(x')}\bigg|_{\eta=\eta^*[\epsilon]} =b^{-1}\frac{\delta^2F_\op[\eta]}{\delta\eta(x)\delta\eta(x')}\bigg|_{\eta=\eta^*[\epsilon]}. \end{aligned} \label{eq:inv.func} \end{equation} Next, \eqref{eq:implicit.eta} and \eqref{eq:inv.func} can be used in concert with the ordinary rules of functional calculus to yield the second variation \begin{widetext} \begin{equation} \begin{aligned} \frac{\delta^2F[\eta_\star[\epsilon],\epsilon]}{\delta\epsilon_\X(x)\delta\epsilon_\X(x')} &=C^0_\X\delta(x-x')- 2b\frac{\delta\eta_\star[\epsilon](x)}{\delta\epsilon_\X(x')} -b\int dx''\,\frac{\delta^2\eta_\star[\epsilon](x)}{\delta\epsilon_\X(x')\delta\epsilon_\X(x'')}\epsilon_\X(x'') +\int dx''\,\frac{\delta^2\eta_\star[\epsilon](x'')}{\delta\epsilon_\X(x)\delta\epsilon_\X(x')}\frac{\delta F_\op[\eta]}{\delta\eta(x'')}\bigg|_{\eta=\eta_\star[\epsilon]}\\ &\qquad\qquad+\int dx''\,dx'''\,\frac{\delta\eta_\star[\epsilon](x'')}{\delta\epsilon_\X(x)}\frac{\delta\eta_\star[\epsilon](x''')}{\delta\epsilon_\X(x')}\frac{\delta^2F_\op[\eta]}{\delta\eta(x'')\delta\eta(x''')}\bigg|_{\eta=\eta_\star[\epsilon]} \\ &=C^0_\X\delta(x-x')- 2b\frac{\delta\eta_\star[\epsilon](x)}{\delta\epsilon_\X(x')} -b\int dx''\,\frac{\delta^2\eta_\star[\epsilon](x)}{\delta\epsilon_\X(x')\delta\epsilon_\X(x'')}\epsilon_\X(x'') +\int dx''\,\frac{\delta^2\eta_\star[\epsilon](x'')}{\delta\epsilon_\X(x)\delta\epsilon_\X(x')}(b\epsilon_\X(x''))\\ &\qquad\qquad+b\int dx''\,dx'''\,\frac{\delta\eta_\star[\epsilon](x'')}{\delta\epsilon_\X(x)}\frac{\delta\eta_\star[\epsilon](x''')}{\delta\epsilon_\X(x')} \bigg(\frac{\partial\eta_\star[\epsilon](x'')}{\partial\epsilon_\X(x''')}\bigg)^\recip\\ &=C^0_\X\delta(x-x')- 2b\frac{\delta\eta_\star[\epsilon](x)}{\delta\epsilon_\X(x')} +b\int dx''\,\delta(x-x'')\frac{\delta\eta_\star[\epsilon](x'')}{\delta\epsilon_\X(x')} =C^0_\X\delta(x-x')-b\frac{\delta\eta_\star[\epsilon](x)}{\delta\epsilon_\X(x')}. \end{aligned} \label{eq:big.boy} \end{equation} \end{widetext} The elastic modulus is given by the second variation evaluated at the extremized strain $\langle\epsilon\rangle$. To calculate it, note that evaluating the second variation of $F_\op$ in \eqref{eq:inv.func} at $\langle\epsilon\rangle$ (or $\eta_\star(\langle\epsilon\rangle)=\langle\eta\rangle$) yields \begin{equation} \bigg(\frac{\delta\eta_\star[\epsilon](x)}{\delta\epsilon_\X(x')}\bigg)^\recip\bigg|_{\epsilon=\langle\epsilon\rangle} =b^{-1}\chi^\recip(x,x')+\frac{b}{C^0_\X}\delta(x-x'), \label{eq:recip.deriv.op} \end{equation} where $\chi^\recip$ is the \op\ susceptibility given by \eqref{eq:sus_def}. Upon substitution into \eqref{eq:big.boy} and taking the Fourier transform of the result, we finally arrive at \begin{equation} C_\X(q) =C^0_\X-b\bigg(\frac1{b\chi(q)}+\frac b{C^0_\X}\bigg)^{-1} =C^0_\X\bigg(1+\frac{b^2}{C^0_\X}\chi(q)\bigg)^{-1}. \label{eq:elastic.susceptibility} \end{equation} Though not relevant here, this result generalizes to multicomponent \op s. At $q=0$, which is where the modulus measurements used here were taken, this predicts a cusp in the static elastic modulus $C_\X(0)$ of the form $|\Delta\tilde r|^\gamma$ for $\gamma=1$. \brad{I think this last sentence, which is the point of the whole paper, needs to be expanded upon and emphasized. It needs to be clear that what we have done is consider a general OP of B1g or B2g type modulated along the c-axis. For a general Landau free energy, it will develop order at some finite q, but if you measure at q=0, which is what ultraound typically does, you still see "remnant" behaviour that cusps at the transition} \begin{figure}[htpb] \centering \includegraphics[width=\columnwidth]{fig-stiffnesses} \caption{ \Rus\ measurements of the elastic moduli of \urusi\ as a function of temperature (green, solid) alongside fits to theory. The vertical yellow lines show the location of the \ho\ transition. (a) $\Btg$ modulus data and fit to standard form \cite{varshni_temperature_1970}. (b) $\Bog$ modulus data and fit to \eqref{eq:elastic.susceptibility}. The fit gives $C^0_\Bog\simeq\big[71-(0.010\,\K^{-1})T\big]\,\GPa$, $b^2/D_\perp q_*^4\simeq6.2\,\GPa$, and $a/D_\perp q_*^4\simeq0.0038\,\K^{-1}$. Addition of an additional parameter to fit the standard bare modulus \cite{varshni_temperature_1970} led to sloppy fits. (c) $\Bog$ modulus data and fit of \emph{bare} $\Bog$ modulus. (d) $\Bog$ modulus data and fit transformed using $[C^0_\Bog(C^0_\Bog/C_\Bog-1)]]^{-1}$, which is prediced from \eqref{eq:susceptibility} and \eqref{eq:elastic.susceptibility} to be linear above $T_c$. The failure of the Ginzburg--Landau prediction below the transition is expected on the grounds that the \op\ is too large for the free energy expansion to be valid by the time the Ginzburg temperature is reached. } \label{fig:data} \end{figure} \emph{Comparison to experiment.} \Rus\ experiments \cite{ghosh_single-component_nodate} yield the full elasticity tensor; the moduli broken into the irrep components defined in \eqref{eq:strain-components} is shown in Figure \ref{fig:data}. The $\Btg$ modulus doesn't appear to have any response to the presence of the transition, exhibiting the expected linear stiffening upon cooling from room temperature, with a low-temperature cutoff at some fraction of the Debye temperature\cite{varshni_temperature_1970}. The $\Bog$ modulus, on the other hand, has a dramatic response, softening over the course of roughly $100\,\K$, and then cusping at the \ho\ transition. While the low-temperature response is not as dramatic as the theory predicts, mean field theory---which is based on a small-$\eta$ expansion---will not work quantitatively far below the transition where $\eta$ has a large nonzero value and higher powers in the free energy become important. The data in the high-temperature phase can be fit to the theory \eqref{eq:elastic.susceptibility}, with a linear background modulus $C^0_\Bog$ and $\tilde r-\tilde r_c=a(T-T_c)$, and the result is shown in Figure \ref{fig:fit}. The data and theory appear quantitatively consistent in the high temperature phase, suggesting that \ho\ can be described as a $\Bog$-nematic phase that is modulated at finite $q$ along the $c-$axis. We have seen that the mean-field theory of a $\Bog$ \op\ recreates the topology of the \ho\ phase diagram and the temperature dependence of the $\Bog$ elastic modulus at zero pressure. This theory has several other physical implications. First, the association of a modulated $\Bog$ order with the \ho\ phase implies a \emph{uniform} $\Bog$ order associated with the \afm\ phase, and moreover a uniform $\Bog$ strain of magnitude $\langle\epsilon_\Bog\rangle^2=b^2\tilde r/4u(C^0_\Bog)^2$, which corresponds to an orthorhombic structural phase. Orthorhombic symmetry breaking was recently detected in the \afm\ phase of \urusi\ using x-ray diffraction, a further consistency of this theory with the phenomenology of \urusi\ \cite{choi_pressure-induced_2018}. Second, as the Lifshitz point is approached from low pressure, this theory predicts that the modulation wavevector $q_*$ should vanish continuously. Far from the Lifshitz point we expect the wavevector to lock into values commensurate with the space group of the lattice, and moreover that at zero pressure, where the \rus\ data here was collected, the half-wavelength of the modulation should be commensurate with the lattice spacing $a_3\simeq9.68\,\A$, or $q_*=\pi/a_3\simeq0.328\,\A^{-1}$ \cite{meng_imaging_2013, broholm_magnetic_1991, wiebe_gapped_2007, bourdarot_precise_2010}. In between these two regimes, the ordering wavevector should shrink by jumping between ever-closer commensurate values in the style of the devil's staircase \cite{bak_commensurate_1982}. This motivates future \rus\ experiments done at pressure, where the depth of the cusp in the $\Bog$ modulus should deepen (perhaps with these commensurability jumps) at low pressure and approach zero like $q_*^4\sim(c_\perp/2D_\perp)^2$ near the Lifshitz point. \brad{Should also motivate x-ray and neutron-diffraction experiments to look for new q's - mentioning this is important if we want to get others interested, no one else does RUS...} Alternatively, \rus\ done at ambient pressure might examine the heavy fermi liquid to \afm\ transition by doping. \brad{We have to be careful, someone did do some doping studies and it's not clear exctly what's going on}. The presence of spatial commensurability known to be irrelevant to the critical behavior at a one-component disordered to modulated transition, and therefore is not expected to modify the critical behavior otherwise \cite{garel_commensurability_1976}. There are two apparent discrepancies between the orthorhombic strain in the phase diagram presented by \cite{choi_pressure-induced_2018} and that predicted by our mean field theory. The first is the apparent onset of the orthorhombic phase in the \ho\ state prior to the onset of \afm. As \cite{choi_pressure-induced_2018} notes, this could be due to the lack of an ambient pressure calibration for the lattice constant. The second discrepancy is the onset of orthorhombicity at higher temperatures than the onset of \afm. Susceptibility data sees no trace of another phase transition at these higher temperatures \cite{inoue_high-field_2001}, and therefore we don't expect there to be one. We do expect that this could be due to the high energy nature of x-rays as an experimental probe: orthorhombic fluctuations could appear at higher temperatures than the true onset of an orthorhombic phase. \brad{I think this paragraph should probably be tigtened up a bit, we need to be more specific about "don't expect there to be one" and "fluctuations"}. Three dimensions is below the upper critical dimension $4\frac12$ of a one-component disordered to modulated transition, and so mean field theory should break down sufficiently close to the critical point due to fluctuations, at the Ginzburg temperature \cite{hornreich_lifshitz_1980, ginzburg_remarks_1961}. Magnetic phase transitions tend to have Ginzburg temperature of order one. Our fit above gives $\xi_{\perp0}q_*=(D_\perp q_*^4/aT_c)^{1/4}\simeq2$, which combined with the speculation of $q_*\simeq\pi/a_3$ puts the bare correlation length $\xi_{\perp0}$ at about what one would expect for a generic magnetic transition. The agreement of this data in the $t\sim0.1$--10 range with the mean field exponent suggests that this region is outside the Ginzburg region, but an experiment may begin to see deviations from mean field behavior within around several degrees Kelvin of the critical point. A \rus\ experiment with more precise temperature resolution near the critical point may be able to resolve a modified cusp exponent $\gamma\simeq1.31$ \cite{guida_critical_1998}, since the universality class of a uniaxial modulated one-component \op\ is $\mathrm O(2)$ \cite{garel_commensurability_1976}. We should not expect any quantitative agreement between mean field theory and experiment in the low temperature phase since, by the point the Ginzburg criterion is satisfied, $\eta$ is order one and the Landau--Ginzburg free energy expansion is no longer valid. \emph{Conclusion and Outlook.} We have developed a general phenomenological treatment of \ho\ \op s with the potential for linear coupling to strain. The two representations with mean field phase diagrams that are consistent with the phase diagram of \urusi\ are $\Bog$ and $\Btg$. Of these, only a staggered $\Bog$ \op is consistent with zero-pressure \rus\ data, with a cusp appearing in the associated elastic modulus. In this picture, the \ho\ phase is characterized by uniaxial modulated $\Bog$ order, while the \afm\ phase is characterized by uniform $\Bog$ order. \brad{We need to be a bit more explicit about what we think is going on with \afm - is it just a parasitic phase? Is our modulated phase somehow "moduluated \afm" (can you modualte AFM in such as way as to make it disappear? Some combination of orbitals?)} The corresponding prediction of uniform $\Bog$ symmetry breaking in the \afm\ phase is consistent with recent diffraction experiments \cite{choi_pressure-induced_2018} \brad{needs a caveat about temperature, so that we're being transparent}. This work motivates both further theoretical work regarding a microscopic theory with modulated $\Bog$ order, and preforming \rus\ experiments at pressure that could further support or falsify this idea. \begin{acknowledgements} This research was supported by NSF DMR-1719490 and DMR-1719875. \end{acknowledgements} \bibliographystyle{apsrev4-1} \bibliography{hidden_order} \end{document}