1. Introduction
Flow in porous media is frequently idealized by defining a domain with homogeneous porosity and permeability, a strong simplification of reality which frequently involves heterogeneities and anisotropies with variations of both properties. In many environments stratification is quite common, often with impermeable layers separating more or less porous lenses with higher permeability representing a preferential path for flows. Such lenses are included in the model by means of spatial averaging, which can be quite effective if a significant anisotropy is present, but is less suitable to interpret the processes if, for example, layering is regular. It is also common that long lenses of high permeability are separated by thin sealing layers of much lower permeability, hence a pressure gradient may develop between adjacent lenses, which are characterised by a flow dynamics largely independent of the neighbouring layers. This layering can be observed as a consequence of sedimentary processes in coastal regions and lakes, where flooding transported muds covering sands and pebbles, or as a consequence of volcanic eruptions, spreading ash and fine particles on previous sand and coarse grained layers. Layers are often accompanied by scattered fracture bands or gaps, which connect them allowing exchange of fluid in the presence of a pressure gradient. The two extreme configurations, with layers separated by impermeable boundaries or with gaps, show a significantly different equivalent permeability in the horizontal and in the vertical direction, as a consequence of the complex pattern of the flow arising in the medium (Phillips Reference Phillips2009). Point sources of buoyant fluid show a dynamics quite influenced by the presence of a strong layering (Hesse & Woods Reference Hesse and Woods2010; Rayward-Smith & Woods Reference Rayward-Smith and Woods2011); hence, detailed analyses are required to infer the effect of this idealized model on buoyancy-driven flow and transport.
A model capturing the essential behaviour of gravity currents advancing in a regular array of horizontal, independent layers having uniform properties has been developed for Newtonian fluids by Farcas & Woods (Reference Farcas and Woods2015). A constant overpressure at the upstream boundary generates in each layer a gravity current propagating with a speed different between layers and increasing with depth. In each layer the body of the current progresses under confined conditions, while the head of the current slumps in free-surface conditions. Upon comparing the total flow rate pertaining to a single layer and to an array of $N$ layers arranged in a package having the same total thickness, it is seen that the layered configuration is associated with a relatively modest increase in flow rate (in the order of a few percent at most if
$N>10$) with respect to the single layer. Of significant interest for field applications is the increased buoyancy-driven dispersion produced by the layering: the mechanical dispersion associated with the speed difference among layers is shown to be dominant, at the field scale, upon in-layer local dispersion linked to pore scale mechanisms. In a layered system with flow parallel to the bedding, this dominance has long been recognized for pressurized domains (Salandin, Rinaldo & Dagan Reference Salandin, Rinaldo and Dagan1991), but not in the context of a buoyancy-driven flow. Significantly, irrespective of the flow details, in both cases the apparent dispersion coefficient increases without bounds with the square root of time (Matheron & De Marsily Reference Matheron and De Marsily1980).
The perfectly layered domain, an archetype in the upscaling literature (Dagan, Fiori & Jankovic Reference Dagan, Fiori and Jankovic2013), seems useful to investigate additional effects of great relevance in subsurface fluid mechanics, such as the effects of fluid rheology on effective flow parameters (Di Federico, Pinelli & Ugarelli Reference Di Federico, Pinelli and Ugarelli2010). The implications of complex, nonlinear rheological fluid models on a flow in a porous and fractured media have been extensively analysed theoretically and experimentally in recent years, since there exist numerous practical applications requiring artificial fluids with a non-Newtonian rheology or foam-like nature, like aquifer remediation (Kananizadeh et al. Reference Kananizadeh, Chokejaroenrat, Li and Comfort2015), fracking technology (Kreipl & Kreipl Reference Kreipl and Kreipl2017), well drilling (Pelipenko & Frigaard Reference Pelipenko and Frigaard2004) and soil reinforcing (Gullu Reference Gullu2015). Also, carbon dioxide storage in aquifers may include some phases when the brine behaves like a non-Newtonian fluid (Wang & Clarens Reference Wang and Clarens2012), and permeability control of porous media can be obtained with polymers activated by temperature (Tran-Viet, Routh & Woods Reference Tran-Viet, Routh and Woods2014). Polyacrylamide solutions are quite frequent in oil field displacement: these solutions can be controlled to develop a cross-linked gel structure, which is prone to shear-thinning behaviour due to alignment of the polymer chains in the presence of a shear (Lee et al. Reference Lee, Morad, Teng and Poh2012). For an increasing concentration of polyacrylamide, yield stress also appears (Yang Reference Yang2001).
A key modelling choice is then the adoption of a specific constitutive equation, as rheological uncertainty is a crucial factor affecting predictions of environmental phenomena (for an example in a related field at a different scale see Bodur & Rey Reference Bodur and Rey2019). A detailed description of the possible constitutive equations of relevance in subsurface applications is outside the scope of the present paper, for a review see Bird, Dai & Yarusso (Reference Bird, Dai and Yarusso1983) and Sochi (Reference Sochi2010). Limiting to time-independent models, the relationship between shear stress and shear rate is described by non-Newtonian models having a variable number of parameters, starting from the Ostwald–De Waele (power-law) model (two parameters, Morrell & De Waele Reference Morrell and De Waele1920; Ostwald Reference Ostwald1929), Herschel–Bulkley (three parameters, Herschel & Bulkley Reference Herschel and Bulkley1926), Cross (four parameters, Cross Reference Cross1965) and Carreau–Yasuda (four or five parameters, Carreau Reference Carreau1972; Yasuda, Armstrong & Cohen Reference Yasuda, Armstrong and Cohen1981). The selected constitutive equation is ready to use for discrete modelling of fractured media, while it is often upscaled to a Darcy-type relationship between flow rate and velocity for porous media; both cases are relevant to the present work. The paper by Savins (Reference Savins1969) provides an early, comprehensive review on non-Newtonian flow in porous media, including scale-up via the capillary bundle model. Pioneering and state-of-the-art work, linked to oil and gas applications, on the extension of Darcy's law to power-law and yield-stress fluids is reported in Barenblatt, Entov & Ryzhik (Reference Barenblatt, Entov and Ryzhik1990). Pearson & Tardy (Reference Pearson and Tardy2002) describe continuum models and the role of length scales in upscaling non-Newtonian flows in porous media. Other than the capillary bundle model, approaches to upscaling include homogeneization (Hornung Reference Hornung1997; Orgeas et al. Reference Orgeas, Geindreau, Auriault and Bloch2007) and network modelling, used for shear-thinning fluids (Balhoff & Thompson Reference Balhoff and Thompson2006) and yield-stress fluids (Sochi & Blunt Reference Sochi and Blunt2008). Direct verification of upscaled laws is also important: a Darcy-type law including a threshold was derived experimentally for Herschel–Bulkley fluids by Chevalier et al. (Reference Chevalier, Chevalier, Clain, Dupla, Canou, Rodts and Coussot2013), obtaining a relationship between pressure gradient and flow rate with a form similar to the rheological model adopted.
Among a multiplicity of industrial applications, significant research efforts have been devoted to non-Newtonian buoyancy-driven flows in pipes, considering exchange yield-stress flow in inclined channels (Fenie & Frigaard Reference Fenie and Frigaard1999), displacement viscoplastic flows in quasi-horizontal (Taghavi et al. Reference Taghavi, Seon, Martinez and Frigaard2009) and horizontal ducts (Eslami, Frigaard & Taghavi Reference Eslami, Frigaard and Taghavi2017) and eccentric annuli (Carrasco-Teja & Frigaard Reference Carrasco-Teja and Frigaard2010). As to porous media flow, single-phase, unconfined gravity currents of power-law nature have been recently thoroughly analysed with a combination of analytical and experimental techniques by various authors (e.g. Pascal & Pascal Reference Pascal and Pascal1993; Bataller Reference Bataller2008; Di Federico, Archetti & Longo Reference Di Federico, Archetti and Longo2012a,Reference Di Federico, Archetti and Longob; Longo et al. Reference Longo, Di Federico, Chiapponi and Archetti2013; Di Federico et al. Reference Di Federico, Longo, Chiapponi, Archetti and Ciriello2014; Longo et al. Reference Longo, Ciriello, Chiapponi and Di Federico2015a; Ciriello et al. Reference Ciriello, Longo, Chiapponi and Di Federico2016; Lauriola et al. Reference Lauriola, Felisa, Petrolo, Di Federico and Longo2018). Typically, closed-form solutions of self-similar form were derived and supported by laboratory experiments conducted both in tanks filled with glass ballotini and Hele-Shaw cells. The solutions incorporate constant and variable-volume, channelized flow and longitudinal or vertical variations of permeability and porosity, and allow modelling, by virtue of the Hele-Shaw analogy, also flow in fractures with a continuous, power-law aperture variation. The approach was extended to, and experimentally validated for, gravity currents of a Herschel–Bulkley fluid by Di Federico et al. (Reference Di Federico, Longo, King, Chiapponi, Petrolo and Ciriello2017), employing an expansion method applicable to other configurations and set-ups. In summary, non-Newtonian scenarios of unconfined gravity currents in single fractures and porous domains with simple geometry seem to be adequately described for Herschel–Bulkley fluids and their sub-case, power-law fluids.
The next step is investigating the impact of fluid rheology on buoyancy-driven dispersion at the field scale for an understanding of contamination mechanisms and remediation issues. A logical starting point is the aforementioned perfectly layered porous medium or array of parallel fractures. Building on the conceptualization of Farcas & Woods (Reference Farcas and Woods2015), we present a theoretical model and its experimental validation for flow of a Herschel–Bulkley fluid in a layered vertical fracture or porous medium. Our approach yields a closed-form self-similar solution describing flow of power-law fluids, while an expansion method is applied for the Herschel–Bulkley fluid which, in general, does not allow such a solution except for special combinations of parameters, (see Balmforth et al. Reference Balmforth, Burbidge, Craster, Salzig and Shen2000; Di Federico et al. Reference Di Federico, Longo, King, Chiapponi, Petrolo and Ciriello2017), or for dominant yield stress (Blake Reference Blake and Fink1990). The joint influence of fluid rheology and layering on longitudinal dispersion around the centre of mass of a solute pulse injected into the fractured or porous system is then theoretically examined with a simplified model. A series of flow tests are conducted to validate the theory in a Hele-Shaw cell for diverse fluids ranging from Newtonian (N) to power-law (PL) and Herschel–Bulkley (HB) fluids; a subset of experiments is devoted to the verification of transport results for N and PL fluids. The main results are consistent with the hypotheses and support the relevance of the models in describing the flow within the experimental uncertainty. As to HB fluids, their flow properties in a porous medium have already been experimentally tested (see Di Federico et al. Reference Di Federico, Longo, King, Chiapponi, Petrolo and Ciriello2017), showing the correctness of the theoretical extension from Hele-Shaw cell to a porous medium, but transport characteristics have not been tested. We expect significant differences in transport in a layered porous medium with respect to a fractured layered domain. This analysis has not been performed and is left for future activity, including its experimental validation.
The paper is structured as follows. Section 2 presents the model for power-law flow in a layered Hele-Shaw cell or narrow fracture, and the corresponding self-similar solution; developments for a layered porous medium are illustrated in appendix A and lead to a dimensionless formulation which is identical to the former, albeit with different scales. The solution to the flow problem is expanded in § 3 to include a yield stress; the corresponding developments for porous media are illustrated in appendix B. Section 4 describes the approach adopted and the results obtained for macro-dispersion, valid for both fractured and porous cases. Section 5 illustrates the flow and transport experiments in the Hele-Shaw cell and their results. The final § 6 reports on the conclusions and perspectives for future work.
2. Flow of a power-law fluid in a layered Hele-Shaw cell
We consider a non-Newtonian fluid described by the Ostwald–De Waele (power-law) model, which reads for simple shear flow as

where $\tau$ is the shear stress,
$\dot {\gamma }$ is the shear rate,
$n$ and
$\mu _0$ are the fluid behaviour index and the flow consistency index (dimensions
$[ML^{-1}T^{n-2}]$), respectively; for
$n=1$,
$\mu _0$ reduces to Newtonian viscosity
$\mu$. Typically the power-law relationship is a simplification of the rheology of non-Newtonian fluids, which can exhibit a yield stress (Herschel–Bulkley model), and/or a time dependency; in many situations, however, the power-law model adequately describes the flow of real fluids. Such a fluid, of density
$\rho$, is supplied by a vertical fracture intersecting
$N$ horizontal layers, each having height
$H_l$ and width
$b$ separated from the top and lower one by an impervious seal of negligible thickness. This array of layers is subject to a constant overpressure (head) equal to
$PH_l$ above the top of the uppermost layer, producing
$N$ different currents, each advancing independently as there is no connection between the layers, see figure 1. If a layered porous medium is considered in lieu of a Hele-Shaw cell, each layer has permeability
$k$ and porosity
$\phi$; for the description of the porous flow, see appendix A.

Figure 1. Schematic of a layered formation with the intruding gravity current arising from a vertical fracture that opens into each of the horizontal layers. Here $x_s$ and
$x_f$ represent the contact point of the current with the top and the base of the layer. The inset shows an enlargement of a single layer with the front separating the saturated and unsaturated domains.
We assume the validity of the shallow water approximation, with the horizontal length scale much larger than the vertical one except at early times. Thus, the pressure gradient along the flow direction is controlled by the gradient of the thickness of the current, $h$, and the difference of density
$\Delta \rho$ between the current fluid and the ambient fluid, according to
$\partial p/\partial x=\Delta \rho \,g\,(\partial h/\partial x)$, where
$g$ is gravity. We further neglect losses in the vertical fracture feeding the layers. Based on the Hele-Shaw/porous medium analogy, the gap-averaged velocity in each layer is hence equal to

where $\varOmega$ is a velocity scale.
The mass conservation equation reads for each layer as

Within each layer, we distinguish a domain $0<x<x_s(t)$ where the flow is confined by the impervious upper and lower boundaries, and a domain
$x_s(t)<x<x_f(t)$ where the current spreads on the lower boundary maintaining a nose shape. In the former domain, the gap-averaged flow velocity is

where the subscript $r$ refers the variables to the
$r$-th layer. As the overpressure
$(P+r-1)H_l$ is constant, the velocity is expected to decay
$\propto x_s^{-1/n}$, as the resistance to flow increases with the length of the current
$x_s$; as a consequence, we expect
$x_s\propto t^{n/(1+n)}$. In the nose region
$x_s(t)<x<x_f(t)$, the mass conservation equation becomes

with the following boundary conditions, valid $\forall t$,

where $Q_r$ is the flow rate in the
$r$-th layer. Defining the velocity, length and time scales as




where we have assumed that $\partial h/\partial x<0$ or
$\partial H/\partial X<0$ .
For the flow in the nose region, we seek a self-similar solution of the form $H_r=f(\eta )T^{\delta }$, where
$X=\alpha \eta T^{\beta }$. Upon substitution of
$H_r$ and
$X$ into (2.9) and in the boundary conditions (2.10a–c), the following identities emerge:

Hence,

Furthermore, (2.9) becomes the nonlinear ordinary differential equation

with the boundary conditions

Defining a normalized variable

(2.13) becomes

with the boundary conditions

This is a boundary value problem parametric in $\eta _f$ and
$\eta _s$, which can be solved by converting it into an initial value problem by imposing a further condition on the first derivative at the front for
$\xi =1$, and then finding the values of the parameters which satisfy the two boundary conditions in
$\xi =0$. The solution becomes singular in
$\xi =1$ and we seek for a power series by expressing
$f_r(\xi )$ near
$\xi =1$ as
$f(\xi )=a_0(1-\xi )^b+a_1(1-\xi )^{b+1}+\cdots$.
This function already satisfies the first boundary condition in (2.17a–c) and upon substitution into (2.16), we calculate

Although the expansion is approximately $\xi =1$, it constitutes a very good approximation of the solution in the whole domain and can be adopted for computing
$\eta _f$ and
$\eta _s$ without solving the differential problem. Substituting (2.18) into the last two boundary conditions in (2.17a–c) yields, for
$\xi =0$,

to be solved numerically. For $n=1$, the results are identical to those reported in Farcas & Woods (Reference Farcas and Woods2015) except for the minor last term in (2.18).
Figure 2 shows the curves representing $\eta _f$ and
$\eta _s$ for
$n=0.5,0.7,1$ obtained upon solving the differential problem, and symbols representing the approximate solution given by (2.19) for
$n=0.5$. The approximate solution matches the numerical results very well in the entire head range; the difference between the two is negligible also for other values of the fluid behaviour index (not shown).

Figure 2. Values of $\eta _f$ and
$\eta _s$ for varying fluid behaviour index
$n$ and for increasing
$P+r-1$. The symbols are the numerical solution of (2.19) for
$n=0.5$.
Figure 3 shows the theoretical profile of the current for three different values of the fluid behaviour index in a configuration with $N=10$ layers. In this format, profiles for shear-thinning fluids are more extended than those for a Newtonian fluid; however, any comparison is best drawn in dimensional form as scales depend on consistency and fluid behaviour indices.

Figure 3. Profile of the currents with $P=0.1$ for fluid behaviour index
$n=0.5, 0.7, 1$.
Similar developments can be carried out for a stratified porous medium having independent layers, leading to an identical dimensionless formulation with different scales; results are reported in appendix A.
3. Extension to Herschel–Bulkley fluids
The Herschel–Bulkley model exhibits a three-parameter relation between shear rate and shear-stress given by

where $\tau _p$ is the yield stress, and
$\mu _0$ and
$n$ are the consistency and fluid behaviour index, respectively. Flow of such a fluid brings a new expression of the gap-averaged velocity based on the analogy established by Di Federico et al. (Reference Di Federico, Longo, King, Chiapponi, Petrolo and Ciriello2017):

Here $\varOmega$ is defined in (2.2) and
$\kappa =2\tau _p/(\Delta \rho \,gb)$ is the dimensionless ratio between yield stress and gravity related stress, and is similar to the Bingham number (Balmforth et al. Reference Balmforth, Burbidge, Craster, Salzig and Shen2000). For the gravity current to advance, it must be
$\kappa <|\partial h/\partial x|$. This condition is generally met at the beginning of the inflow process, but as the fluid invades more of the formation, the flow gradually slows down with a progressively smaller
$|\partial h/\partial x|$ which could lead to a halt and with the shape of the current reflecting the yield-stress value. In theory the distance at which the current stops could be evaluated, but in practice in this condition the present model fails for several reasons: we bear in mind that the shallow water hypotheses are violated near the tip of the current, where, for example, the bottom stress is of the same order of the vertical walls stress; also, higher-order terms in the governing equation are requested for a proper description of the flow field (Lipscomb & Denn Reference Lipscomb and Denn1984). The complexity of the analysis increases if we consider that many real fluids described by a Herschel–Bulkley model show thixotropy, with two different critical yield-stress values (Hewitt & Balmforth Reference Hewitt and Balmforth2013).
In the fully saturated domain $0<x<x_s$ the gap-averaged velocity in the
$r$-th layer transforms into

while within the nose ($x_s(t)<x<x_f(t)$), the continuity equation becomes

with the boundary conditions (2.6a–c).
By adopting the same velocity, length and time scales of the power-law case (2.7a–c), (3.3)–(3.4) become


with the boundary conditions


Again, we have assumed that $\partial h_r/\partial x<0$ or
$\partial H_r/\partial X<0$.
The yield stress introduces a new length scale in the process, and a self-similar solution is not possible in general. We do not pursue this approach and instead apply an asymptotic analysis similar to that developed for a gravity current of Herschel–Bulkley fluid slumping in a vertical fracture (see Di Federico et al. Reference Di Federico, Longo, King, Chiapponi, Petrolo and Ciriello2017).
For $\kappa \to 0$, (3.5)–(3.7a,b) collapse to (2.8)–(2.10a–c) and admit a self-similar solution where
$H_r=f_r(\eta )$ and
$X=((n+1)/n)^{n/(n+1)}\eta T^{n/(n+1)}$ as shown above. For
$\kappa > 0$, we define

and propose the expansion


where $f_{0}(\eta )\equiv f_{r}(\eta )$ is the self-similar solution for power-law fluids, and
$\chi _1,\chi _2,\ldots$ are constant coefficients to be evaluated. For ease of notation, we have omitted the subscript
$r$ relative to the
$r$-th layer. The variable
$\psi$ introduces a new time scale, hence a new velocity scale, and satisfies the balance of the terms in (3.6). In fact, according to (3.9), the expansion is valid for

with $T_c$ a critical time value inversely depending on a power of
$\kappa$, as the yield stress delays the spreading. Figure 4 depicts
$T_c(n,\kappa )$; it is seen that the temporal domain of validity of the expansion is more extended for smaller
$\kappa$ and
$n$, i.e. very shear-thinning fluids with relatively small yield stress with respect to the effect of the horizontal pressure gradient. For a Bingham fluid (
$n=1$),
$T_c=1/\kappa ^2$ and
$X \propto T^{1/2}$.

Figure 4. Dimensionless critical time $T_c$ as a function of the fluid behaviour index
$n$ and dimensionless yield stress
$\kappa$. The hatched area indicates the domain where the condition
$\psi <1$ is satisfied for a Bingham fluid with
$n=1$.
By inserting (3.10)–(3.11) in (3.6) and balancing the terms with the same power of time, at order $O(\psi ^0)$ we recover the fundamental solution for power-law fluids. At
$O(\psi )$ the continuity equation (3.6) becomes

where the unknown function $f_{1}$ is forced by the fundamental solution
$f_{0}$ and
$\eta _s,\eta _f$ refer to the fundamental solution. The boundary conditions are computed by expanding the terms in (3.7a,b) near
$\eta _s$,
$\eta _f$. At
$O(\psi )$ the boundary conditions are

Introducing the normalized variable $\xi =(\eta -\eta _s)/(\eta _f-\eta _s)$, the governing differential equation becomes

with the boundary conditions

The solution $f_1$ of (3.15) with the last two boundary conditions at
$\xi =0$ in (3.16a–c) is parametric in
$\chi _1$, which is unknown. Following the same approach adopted for a power-law fluid,
$f_1$ and
$\chi _1$ are computed by imposing that the first boundary condition at
$\xi =1$ is satisfied.
Figure 5 shows the values of $\chi _1$ for different values of the upstream overpressure
$P+r-1$ and for different values of the fluid behaviour index. The coefficient
$\chi _1$ is always negative, increases for increasing overpressure and increasing
$n$, is generally of
$O(1)$ and exhibits more marked variation for smaller values of overpressure and
$n$. Figure 6 shows the correction to the position of the front due to the presence of the yield stress for different values of the fluid behaviour index
$n$ and different values of
$\kappa$. The reduction of the speed of the front is more evident for increasing
$\kappa$, while the more shear-thinning the fluid, the more noticeable the slowdown of the front.

Figure 5. Values of $\chi _1$ (a) for different dimensionless overpressure
$P+r-1$ versus
$n$, and (b) for varying fluid behaviour index
$n$ versus overpressure.

Figure 6. The effects of the first-order correction on the position of the front for $P+r-1=3$, (a)
$\eta _f$ and (b)
$\eta _s$. The thick, mid and thin curves refer to
$n=1$,
$n=0.7$ and
$n=0.5$, respectively. The continuous, dashed and dot–dashed curves refer to
$\kappa =0$ (power-law fluid),
$\kappa =0.01$ and
$\kappa =0.05$, respectively. The grey curves for
$n=1$ and
$\kappa =0.05$ are unphysical since they indicate an overturning of the front with
$\eta _f<\eta _s$.
Figure 7 shows the functions $f_0$ and
$f_1$ for different values of
$n$, with the three curves representing
$f_0$ practically overlapping. The function
$f_1$ is of the same order of
$f_0$ and its contribution to
$H_r$ is small since
$f_1$ is multiplied by
$\psi \ll 1$. The derivative of
$f_1$ is always negative, hence the approximation
$\partial H_r/\partial x<0\equiv f_0'+\psi f_1'<0$ is satisfied.

Figure 7. The computed solutions $f_0$ and
$f_1$ for
$P+r-1=3$ and for different values of fluid behaviour index
$n$. The three curves representing
$f_0$ are almost coincident.
Figure 8 shows the profiles of the current at $T=10$ for
$P+r-1=3$. The case
$\kappa =0$ corresponds to the power-law solution, for
$\kappa >0$ the presence of yield stress reduces the advancement of the front position without other significant variations. For larger values of time and overpressure, the profiles are more shifted, showing an increasing delay with respect to the base power-law case (not shown).

Figure 8. The effects of the first-order correction on the current profiles at time $T=10$ for
$P+r-1=3$. The thick, mid and thin curves refer to
$n=1,0.7,0.5$, respectively. The continuous, dashed and dotted curves refer to
$\kappa =0$ (power-law fluid) and to
$\kappa =0.01,0.05$, respectively.
Similar developments can be carried out for a perfectly stratified porous medium having independent layers; results are reported in appendix B.
4. Effects of layering on total flow rate and migration of a solute cloud for a power-law fluid
Layering affects both the flow rate and transport properties of buoyancy-driven gravity currents. In the following, our results are illustrated for the fractured case with the length, velocity and time scales defined by (2.7a–c) in § 2, but are equally valid for the porous case described in appendix A by substituting the former scales with those defined in (B 6). Upon comparing a single layer of total height $NH_l$ and
$N$ independent layers each of uniform height
$H_l$, it is first seen that the flow rate increases, albeit not too significantly. Second, the simultaneous release of a solute pulse in each layer produces intra-layer (local) dispersion, and macro-dispersion at the system scale. We concentrate on the latter phenomenon and derive an expression for the spatial moment of a solute cloud around its centre of mass. The order of magnitude of the dispersive phenomena is then discussed and compared.
4.1. Increase in flow rate
For a given overpressure, the flux in the $r$-th layer is

where $\eta _s(n,P,r)\equiv \eta _s(n,P+r-1)$, and the sum of the fluxes for all layers is

The flux in a single layer of height $NH_l$ is

where the overpressure $P/N$ appears in order to obtain the same dimensional value
$PH_l$ of overpressure at the roof of the first layer in both configurations. The relative variation of the flux is

In a similar manner we can calculate the variation of the foremost front, obtaining

where $x_{f1}$ is the front position of a single equivalent layer of height
$NH_l$ and
$x_{f\kern0.5pt N}$ is the most advanced front among the
$N$ layers (the deepest one).
Figure 9(a) shows the relative variation of the flow rate versus number of layers, with $0<\Delta Q<\approx 6\,\%$. For a given overpressure, a stratified aquifer allows a higher flow rate than a single layer with the same total height, with a maximum variation that increases with the number of layers and decreases with the fluid behaviour index
$n$. On the contrary, figure 9(b) shows that the front most advanced position within a layered medium (the front in the deepest layer) is less advanced than the front position of a single equivalent layer (
$-12\,\% <\Delta X_f<0$), with greater differences for an increasing number of layers and larger
$n$ values. An increase of
$P$ reduces the effects of layering for both variables, as the overpressure is relatively more uniform throughout the layers. In summary, a shear-thinning fluid flow reduces the effects of layering on both flow rate and maximum current spread.

Figure 9. Power-law fluids. The effects of layering and of flow behaviour index $n$ on (a) the flow rate and (b) the foremost front position for
$P=1$ (filled circles) and
$P=5$ (empty circles).
4.2. Effects on transport
To quantify dispersion at the system scale, we now consider a pulse of dye injected at a given section of the $r$-th layer of a fracture or porous formation when a current of power-law fluid is already present.
The dye trajectory satisfies the differential equation $\text {d}X/\text {d}T= U$, where the flow speed for a power-law fluid is given by (2.4)

where we have inserted the expression for $X_s$. In the nose of the
$r$-th layer, we obtain

The approximation (2.18) gives

If the parcel is left at $X_0$ at time
$T_0$, it advances in the saturated portion of the lens and enters the nose region at
$X=X_1\equiv X_s(T_1)$ at the time
$T_1$, where

or, in explicit form,

Then, in the nose region ($X>X_1$ and
$T>T_1$) the path is given by the differential equation

to be integrated by imposing that $X(T_1)=X_1$. For
$n=1$, the result is equal to the result given in Farcas & Woods (Reference Farcas and Woods2015) except for a typo. Equation (4.11) admits an analytical solution also for
$n=1/2$,

while for different values of $n$ a numerical solution is needed. The differences between the fully numerical solution and both the analytical solutions for
$n=1$ (Farcas & Woods Reference Farcas and Woods2015) and
$n=1/2$ in (4.12) are negligible except for very large values of
$T$ (not shown).
If we consider a slug of unitary width released in a single layer at time $T_0$ in the interval
$0<X<1$, it first advances in the pressurized domain (
$X<X_s$) with constant speed of the two fronts, without changing its width. In an intermediate position, with part of the slug in the nose and part in the pressurized portion of the layer, the width of the slug decreases or increases according to the value of
$n$. Then the slug enters fully into the nose and increases its width due to the greater speed of the free-surface front. These different positions occupied by the slug with respect to the two fronts
$X_s$ and
$X_f$ are illustrated in figure 10, which shows its time evolution for fluids with a different flow behaviour index; differences in dimensionless coordinates are relatively modest among the fluids.

Figure 10. Evolution of slugs of unitary width released at $T=5$ in the interval
$X\in [0-1]$ in a 10-layer aquifer, with
$P+r-1=1$ for (a) a power-law fluid with
$n=0.5$, (b) a power-law fluid with
$n=0.7$ and (c) a Newtonian fluid. The grey lines are the trajectories of the front and of the back of the slug.
Figure 11 illustrates the slug evolution over time and space for fluids with a different behaviour index. The centre of mass of the slug advances and the dispersion generally increases in time. The direct comparison between fluids with a different $n$ is impractical in dimensionless coordinates since the velocity and time scales are different.

Figure 11. The advancement of a finite slug released at $T=5$ and of unitary width
$X\in [0,1]$. (a) Snapshots at
$T=5,100,500,1000$ for a power-law fluid with
$n=0.5$; (b) and (c) as in (a) for a power-law fluid with
$n=0.7$ and for a Newtonian fluid, respectively.
In order to evaluate longitudinal macro-dispersion, we consider the ratio between the standard deviation and the centre of mass of the slugs in different layers, assuming that the mass of each slug is concentrated in its own centre of mass, hence neglecting local dispersion occurring for the single slug (a comparison between the two phenomena is performed later). The centre of mass of the slugs is

where $X_r$ and
$A_r$ are the centre of mass and the mass of the slug in the
$r$-th layer. The variance (second spatial moment of the solute cloud) is

and the ratio STD/mean is equal to $\sigma /\bar {X}$.
An approximate expression for the centre of mass and variance can be derived as follows. We assume that the slugs are released by injecting fluid in the same section for a finite-time interval. The mass of the slug in the $r$-th layer varies as the volume flux in that layer, which is proportional to
$(P+r-1)^{1/n}/\eta _s^{1/n}$; see (2.8). Equation (2.19) can be approximated as

hence, the mass of the slug scales with $\eta _f$. The position of the slug is also proportional to
$\eta _f$, see (4.12) for the special case
$n=1/2$; therefore, the centre of mass of the
$N$ slugs varies as

and the variance with respect to the centre of mass varies as

Hence, the approximate relations based on summations are

and

The approximate time scalings from (4.18) to (4.19), $\bar {X} \sim T^{n/(n+1)}$ and
$\sigma ^2 \sim T^{2n/(n+1)}$, reduce to
$\bar {X} \sim \mathrm {const}$ and
$\sigma ^2 \sim \mathrm {const}$ for
$n \rightarrow 0$, and to
$\bar {X} \sim T$,
$\sigma ^2 \sim T^2$ for
$n \rightarrow \infty$. For
$n=1$, they collapse to the expressions given by Farcas & Woods (Reference Farcas and Woods2015) with
$\bar {X} \sim T^{1/2}$,
$\sigma ^2 \sim T$. The time scaling of the variance is sublinear for shear-thinning fluids with
$n<1$, and supralinear for shear-thickening fluids with
$n>1$. Correspondingly, one could define a longitudinal macro-dispersion coefficient
$D_{LM}$ according to the classical definition
$1/2 \times \text {d}(\sigma ^2)/\text {d}t$, and
$D_{LM}$ would scale with time as
$T^{(n-1)/(n+1)}$.
Figure 12(a) shows the exact values (derived via integration of the differential problem) of the ratio $\sigma /\bar {X}$, a measure of dispersion, for an aquifer with
$N=10$ layers. For large overpressure
$P$, the ratio is almost the same for all fluids, with slightly higher values for shear-thinning fluids, and asymptotically reaches a common value
${\approx }5\,\%$. Variation over time in the examined range is negligible, with results at
$T=100$ almost coincident with the results at
$T=1000$. In terms of dispersion, these results indicate that the length scale of the tracer cloud scales with time identically to the position of the centre of mass, being asymptotically a constant fraction (here approximately
$5\,\%$) of the distance travelled by the plume. However, since the differences addressed to the
$n$ fluid behaviour index are modest, the data field obtained with different fluids can be hardly useful to detect the aquifer structure. Figure 12(b) shows also the approximated values of
$\sigma /\bar {X}$ (bold curves) according to (4.18) and (4.19). The approximation is fairly good providing that the overpressure
$P$ is sufficiently high.

Figure 12. Comparison of the ratio of the standard deviation to the mean as a function of the overpressure $P$ for fluids with a different behaviour index
$n$ at (a)
$T=100$ and (b) at
$T=1000$. The aquifer has
$N=10$ layers and the slug of unitary width is released at
$T=5$. Dashed curves refer to numerical computation, bold curves refer to the approximation in (4.18) and (4.19).
Figure 13(a) plots again the ratio $\sigma /\bar {X}$ for fluids with different fluid behaviour index
$n$ as a function of overpressure
$P$ and number of layers
$N$, the total height being the same. Smaller values refer to a Newtonian fluid, whilst shear-thinning fluids favour larger dispersion values. Minimum dispersion is obtained with large
$P$ and a limited number of layers (
$\sigma /\bar {X}=0$ for a single layer), maximum dispersion corresponds to reduced overpressure
$P$ values and many layers. Figure 13(b) compares
$\sigma /\bar {X}$ for varying
$N$ and
$n$, with
$P=N$. Shear-thinning fluids slightly increase dispersion with respect to a Newtonian fluid.

Figure 13. (a) Comparison of the asymptotic ratio $\sigma /\bar {X}$ for varying overpressure
$P$ and number of layers
$N$, and for fluids with
$n=0.5,0.7,1$. Contours refer to
$\sigma /\bar {X}=5\,\%$ (bold curves),
$10\,\%$ (dashed curves),
$15\,\%$ (dash–dotted curves),
$20\,\%$ (dash–dot–dot curves) and
$25\,\%$ (dotted curves). (b) The variation of the asymptotic ratio
$\sigma /\bar {X}$ for varying
$n$ and varying overpressure
$P$, with a number of layers
$N=P$.
These comparisons do not involve scales, and depend only on $n$, although the dimensional time required to reach the asymptotes depends on the time scale given by (2.2) or (A 10) that, in turn, are related to the nature of the medium (fractured or porous), its geometry and the rheological and physical properties of the fluid.
To compare large-scale (inter-layer) dispersion thus examined with local (intra-layer) transport phenomena, we recall that for power-law flow in porous media of grain size $\delta$ and flow speed
$u$ the Peclet number is (Delgado Reference Delgado2006)

where $D_L$ is the (local) longitudinal dispersion coefficient. With
$\delta = 10^{-4} \ \mathrm {m}$,
$u \approx 10^{-7} \ \mathrm {m}\,\text {s}^{-1}$,
$\rho \approx 1000 \ \mathrm {kg}\,\text {m}^{-3}$,
$\mu _0 \approx 1 \ \mathrm {Pa} \, \text {s}^{{n}}$ and
$n \approx 0.5$, the second term in (4.20) is negligible and hence
$D_L \sim \delta u$. According to (4.6), for a power-law fluid in a porous layer, the velocity scales with time as

with $H$ and
$V$ being length and velocity scales. Coupling these two results leads to the relationship

for the typical length scale of a localised slug of tracer. On the other hand, (4.19) yields

for the length scale of the cloud of tracer slugs. Comparing the two scales, macro-dispersion prevails when

In a field setting, $H \approx 1 \ \mathrm {m}$,
$V \approx 10^{-7} \ \mathrm {m}\,\text {s}^{-1}$, and with
$\delta = 10^{-4} \ \mathrm {m}$ this yields
$t_0 = 10^{-5} \ \mathrm {s}$ for
$n = 0.5$,
$t_0 = 0.1 \ \mathrm {s}$ for
$n = 1$ and
$t_0 = 2.15 \ \mathrm {s}$ for
$n = 1.5$, showing the high sensitivity of threshold time
$t_0$ to the value of
$n$ in the shear-thinning range.
In a layered HS cell or fracture of gap $b$, the scalings are partially different as
$D_L \sim (\bar {u}^2b^2)/D_0$ (Dejam Reference Dejam2019), where
$D_0$ is the molecular diffusion coefficient; this leads to

for the length scale of a localised slug. The typical length scale of the solute cloud is unchanged, hence the condition for the macro-dispersion to prevail is independent from scales $H$,
$V$ and flow behaviour index
$n$ and reads
$t \gg t_0 \equiv b^2/D_0$. With
$b = 10^{-3} \ \mathrm {m}$ and
$D_0 \approx 10^{-9}\ \mathrm {m}^2\,\text {s}^{-1}$, one has a threshold time of the order of
$10^3 \ \mathrm {s}$.
5. The experiments
In order to validate the theoretical model, a set of experiments were conducted in a Hele-Shaw cell with a small gap, simulating a layered fracture. Figure 14 shows the experimental device, a $75\ \mathrm {cm}$ long,
$18\ \mathrm {cm}$ high cell made of two parallel plates of transparent plastic, separated by a gap
$b=0.25\ \text {cm}$. Six layers, each with height
$1.9\ \mathrm {cm}$, were separated with sealing baffles. On one side, a large well, fed either by a vane pump or a small tank (for higher viscosity fluids) and having an overflow, guarantees a constant overpressure during fluid propagation. A lock gate (a plastic tube acting as an O-ring) was initially put in place and then fast removed at the start of the experiment. The profiles of the advancing noses of the current in the layers were detected with a high-resolution video camera (Canon Legria
$1920 \times 1080$ pixels) operating at
$25\ \mathrm {frames}\,\mathrm {s}^{-1}$ and with a single shot camera (Canon EOS 2D,
$3456\times 2304$ pixels). The images, extracted with a varying time step according to the speed of the current, were then post-processed in order to be referenced to a lab coordinate system. With this set-up, the overall accuracy in detecting the profile of the current was approximately
$0.1\ \text {cm}$, whilst the uncertainty in measuring timings was negligible.

Figure 14. A sketch of the Hele-Shaw cell. Front (a) and side (b) view, and (c) a snapshot.
A first set of experiments was conducted with Newtonian fluids, a second set with power-law shear-thinning fluids and a third set with Herschel–Bulkley fluids. Newtonian fluids were glycerol and a mixture of water and glycerol, power-law fluids were obtained by adding Xanthan Gum with different concentration, HB fluids were obtained by mixing deionized water, Carbopol 980 ($0.05\text {--}0.14\,\%$), ink, with subsequent neutralization by adding NaOH. The mass density was measured with a hydrometer (STV3500/23 Salmoiraghi) or with a pycnometer, or with a DMA 5000 by Anton Paar, with an accuracy
$\Delta \rho /\rho \le 0.1\,\%$. The rheologic behaviour of the fluid was analysed with a Ubbelohde viscometer and with a parallel plate rheometer (Anton Paar Physica MCR 101) conducting several different tests to evaluate the fluid behaviour index (for power-law and HB fluids), the consistency index and the yield stress (for HB fluids). Figure 15(a,b) shows the shear-stress/shear-rate experimental data and the apparent viscosity, respectively, for five measurements taken for the same power-law fluid, three at
$\varTheta =22\ ^{\circ }\text {C}$ and two at
$\varTheta =24\ ^{\circ }\text {C}$. There is an adequate repeatability of the data and a minor dependence on temperature, with apparent viscosity decreasing with increasing temperature. It is also evident that the range of shear rate considered for interpolation affects the values of the consistency index
$\mu _0$ and of the fluid behaviour index
$n$, which is smaller in the low shear-rate domain. The estimated accuracy is
$\Delta n/n\le 4.5\,\%$,
$\Delta \mu _0/\mu _0\le 7\,\%$ and
$\Delta \tau _p/\tau _p\le 10\,\%$. The parameters of the experiments are listed in table 1.

Figure 15. Rheometry measurements for the power-law fluid used in experiment 8. (a) Shear-stress/shear-rate experimental values and (b) apparent viscosity. Open circles refer to measurements at $\varTheta =24\ ^{\circ }\text {C}$, stars refer to measurements at
$\varTheta =22\ ^{\circ }\text {C}$. The bold line is the interpolation for high shear rates, with
$n=0.39$ and
$\mu _0=0.87\ \text {Pa}\,\text {s}^n$, the dashed line is the interpolation for low shear rates, with
$n=0.36$ and
$\mu _0=0.91\ \text {Pa}\,\text {s}^n$.
Table 1. Parameters for the experiments in a layered cell/fracture. The last column indicates if tracer was injected at $x=7.2\ \text {cm}$.

Figure 16 shows three snapshots of experiment 7 with power-law fluids and $N=6$ layers. It is evident the different speed of the current in the layers and the different shape of the nose, more elongated for the upper than for the lower layers.

Figure 16. Three snapshots of experiment 7. Photographs were taken 1, 5, 10 s after the removal of the gate lock.
Figure 17(a,b) shows the nose front position $X_f$ and the limit of the pressurized current
$X_s$ within the six layers for a Newtonian and a power-law fluid, respectively. The agreement between theory and experiments improves with time, as a consequence of the asymptotic intermediate self-similarity of the theoretical model. A similar behaviour is shown in figure 18(a,b), depicting the front position of the nose and the limits of the pressurized current in the six layers for a HB fluid. Again, experimental values tend asymptotically to theoretical ones, faster for the lower layers and with a slightly better match for the nose position than for the upper layers.

Figure 17. Dimensionless positions of the nose front and of the limit of pressurized current (a) for experiment 4 with a Newtonian fluid and (b) for experiment 7, a power-law fluid. The green lines refer to the theoretical nose front $X_f$, the red lines refer to the theoretical limit of the pressurized current
$X_s$, the symbols are the experiments. Data for
$X_f$ are translated in the vertical of 10 units for an easy visualization.

Figure 18. Dimensionless front positions for experiment 11 with an HB fluid. (a) Front position at contact with the bottom and (b) limit of the pressurized current. The lines are the theoretical values, the symbols are the experiments.
There is initially a certain discrepancy between theory and experiments; table 2 summarizes for three selected experiments with N, PL and HB fluids, the minimum (for upper layers) and maximum (for lower layers) relative error for the position of the pressurized body $X_s$ and of the nose
$X_f$ at
$t = 180\ \text {s}$ for Newtonian and power-law fluids, and at
$t = 18\ \text {s}$ for the Herschel–Bulkley fluid (the HB experiments are shorter); corresponding root-mean square relative errors averaged over the remaining duration of each experiment are also listed. The error at
$t = 180\ \text {s}$ is slightly larger for power-law fluids than for Newtonian, and lowest (at
$t = 18\ \text {s}$) for HB fluids for both body and nose. The error in the late portion of the experiments follows the same trend but somewhat decreases.
Table 2. Minimum and maximum relative error in the six layers for the position of the pressurized body $X_s$ and of the nose
$X_f$, computed as
$\Delta _{X_s}= \vert X_{s-exp}-X_{s-th} \vert / X_{s-th}$ and
$\Delta _{ X_f} = \vert X_{f-exp}-X_{f-th} \vert / X_{f-th}$ at
$t=180\ \text {s}$ for N and PL, at
$t=18\ \text {s}$ for HB, respectively. The third and fourth columns
$\overline {\Delta _{X_s}}$ and
$\overline {\Delta _{X_f}}$ list the root mean square values for
$t>180\ \text {s}$ (till the end of each experiment) for N and PL fluid experiments, and for
$t>18\ \text {s}$ for the HB fluid experiment. Here N, PL and HB stands for Newtonian, power-law and Herschel–Bulkley fluid, respectively.

By observing figures 17(a,b) and 18(a,b) one would conclude that experiments with Newtonian and PL fluids show a better agreement with theory than those with HB fluids, with a faster approach to intermediate asymptotics. Data listed in table 2 indicate that the opposite is true, with the HB fluid experiment showing less deviation from the theory than the Newtonian and PL fluid experiments. However, this last result is odd and cannot be generalized since HB fluid experiments are expected to deviate from theory more than the Newtonian and PL fluid experiments, as a consequence of (i) model approximations, as the solution for HB fluids is based on an expansion of the self-similar solution for PL fluids; (ii) the accuracy of rheometric measurements, which are known to be less accurate in the presence of yield stress; (iii) possible slip at the wall. Further sources of discrepancy common to experiments with PL and HB fluids are effects missing in the model, such as surface tension phenomena in the nose and inaccuracies in the channel geometry. Finally, there is an intrinsic approximation in the adoption of relatively simple constitutive equations (2/3 parameters for PL/HB fluids) to interpret non-Newtonian behaviour over a wide range of shear rates. More complex rheological models could give a better fit to the experiments. We also recall the difficulty in comparing experiments where the rheology and duration is different.
Figure 19 shows the flow front for experiment 9, PL fluid, and for experiment 12, HB fluid, rescaled to the value $\eta =X/[T(n+1)/n]^{n/(n+1)}$ and
$\eta =X/[T(n+1)/n]^{n/(n+1)}(1+\psi \chi _1)$, respectively, and the vertical position from the base of the cell has been rescaled to the interval
$[0,1]$. The collapse to a single curve, predicted by the theory, is better for PL than for HB fluids, improves in time and with the overpressure (the depth) of the layer.

Figure 19. Comparison of the experimental data for the height of the current as a function of the distance from the source with the theoretical prediction. (a) Experiment 9, power-law fluid and (b) experiment 12, Herschel–Bulkley fluid. The horizontal axis shows the dimensionless $\eta$ from (2.12a,b) and from (3.11), respectively. The experimental data refer to different times as shown in the legend, the curves refer to the theoretical model.
In a subset of experiments, conducted with Newtonian and power-law fluids, pulses of tracer were simultaneously injected in the six layers at mid-height and mid-width and a distance $x=7.2\ \text {cm}$ from the origin, in order to validate the model for the transport properties of the domain.
Figure 20 shows the migration of four series of pulses injected in the flow during an experiment with a power-law fluid. The pulses travel noticeably faster than the mean velocity of the current, elongate during travel, spread little in the other two directions and tend to accumulate near the front when they reach the nose. There is specific evidence that the injected tracer concentrates at the mid gap position; see figure 21 where the red shadow of the slugs is visible on the back wall of the Hele-Shaw cell. The slug advances like a comet, with the tail due to the transversely diffused fluid that moves towards areas near the walls characterized by a lower speed. In correspondence of the axis of the cell the shear rate tends to zero reducing the dispersion of the slug nose, which maintains its structure and advances with a reduced loss of mass. In the nose the tracer is convected down to the bottom of the layer, see the coloured nose of the upper layer, and the flow field is three-dimensional, hence cannot be properly described within the present model.

Figure 20. A snapshot of experiment 9, power-law fluid, at $t=35\ \text {s}$. The circles are the injection points, the dashed curves are the envelopes of the slugs generated by four injections at
$t=7, 15, 25, 33\ \text {s}$, synoptic for all layers.

Figure 21. A snapshot of experiment 10, power-law fluid. The shadow projected by the slugs on the back wall indicates that the slugs do not fill the entire gap of the cell. Photographs was taken 15 s after the removal of the gate lock.
Faster plumes in lower layers spread more in the longitudinal direction due to increased shear dispersion, proportional to mean velocity squared for pressurized flow of Newtonian (Berkowitz & Zhou Reference Berkowitz and Zhou1996) and power-law fluids (Dejam Reference Dejam2019) in a single fracture. An order of magnitude analysis based on the latter formulation confirms shear dispersion prevails over molecular diffusion; the Peclet number $Pe = \bar {u}(b/2)/D_0$ in each fracture is of order
$5 \times 10^3 \ \text {s}$ (with mean velocity
$\bar {u}=0.005 \ \text {m}\,\text {s}^{-1}$, half-gap
$b/2 = 1.25 \times 10^{-3} \ \text {m}$ and diffusion coefficient in water
$D_0 \approx 10^{-9} \ \text {m}^2\,\text {s}^{-1}$); the diffusion time across the gap is
$t_D=(b/2)^2/D_0$, decidedly larger than the duration of the experiments. As all evidence confirms that each pulse remains localized at the centre of the current, the comparison with the experiments is drawn considering the actual path of particles by introducing the maximum speed of the current in (4.6), with
$U_{max}/U=(2n+1)/(n+1)$.
Figure 22 shows a comparison of $X_s$,
$X_f$ and of the path of the slugs for
$r=1,3,6$. The experimental position of the slugs is still ahead of the theoretical one, as a consequence of secondary effects such as the lack of symmetry of the velocity field with respect to the fully confined scheme, leading to enhanced dispersion, but the agreement is quite satisfactory for slugs advancing in the pressurised portion of the current (all series of injections in layers
$r=3$ and
$r=6$ and for
$i2$,
$i3$ and
$i4$ for layer
$r=1$) and in the nose (injection
$i1$ for layer
$r=1$).

Figure 22. Advancement of the slugs for experiment 5, Newtonian fluid. (a) Experimental $X_s$ (orange x) and
$X_f$ (green +), and slug position at four different time of injection (
$i1,i2,i3,i4$) for
$r=1$, (b) for
$r=3$, and (c) for
$r=6$. Symbols are experimental data, curves are theoretical values.
6. Conclusions
Buoyancy-induced dispersion is often encountered in the subsurface environment, which is plagued by pollution problems and remediation needs. Both environmental contaminants and remediation fluids have often a rheologically complex nature, prompting the need to investigate the influence of the fluid nature on flow and transport in geologic media. We have developed novel analytical models to describe the motion of Ostwald–De Waele and Herschel–Bulkley fluids slumping under gravity in a narrow stratified fracture or porous medium constituted of independent layers, and analysed the long-term dispersion associated with the simultaneous release of a solute pulse in all layers. Experiments were conducted in an analogue Hele-Shaw cell with $N=6$ layers to validate the theoretical findings using fluids of different rheology. Our results lead to the following major conclusions.
(a) Gravity currents, regardless of the fluid nature, are constituted by a body occupying the full layer height and a nose slumping ahead of the body. Each portion is associated with a different velocity of propagation, which depends upon rheology and time
$t$. For a power-law fluid of rheological index
$n$, the nose of the current advances as
$x_s \propto {t^n/(n+1)}$. Gravity currents advancing under a larger pressure gradient propagate faster and show a different overall shape with a proportionally shorter nose section.
(b) The analytical solution for power-law fluids is extended to Herschel–Bulkley rheology via an expansion valid below a dimensionless critical time
$T_c = \kappa ^{-(n+1)/n}$, a function of rheological index
$n$ and dimensionless yield stress
$\kappa$.
(c) The presence of independent layers affects both flow and transport, the second much more than the first. The relative increase in flow rate associated with the presence of
$N$ layers with respect to a single one, all other conditions being equal, tends to an asymptotic value of a few percent at most as
$N \rightarrow \infty$, depending on flow behaviour index and overpressure.
(d) For a large number of layers
$N$ and large overpressure
$P$, the centre of mass of an instantaneously released solute cloud scales with time as
$\bar {X} \sim T^{n/(n+1)}$, the variance as
$\sigma ^2 \sim T^{2n/(n+1)}$. Correspondingly, the ratio
$\sigma /\bar {X}$ (a measure of dispersion of the solute cloud around its centre of mass) is remarkably independent of time, while it decreases with increasing overpressure and modestly depends on flow behaviour index. The time scaling of the variance is sublinear for shear-thinning fluids with
$n<1$, and supralinear for shear-thickening fluids with
$n>1$.
(e) The macro-dispersion induced by the layering prevails upon local dispersion beyond a threshold time that is very sensitive to rheology for fractured media.
(f) Experimental results match the theoretical predictions of the profile of the advancing currents for large time, a consequence of the asymptotic intermediate self-similarity of the theoretical model. The match is better for a large number of layers and for the nose position.
(g) The advancement of the solute slugs in transport experiments is typically under-predicted by the theory as a consequence of molecular diffusion and other effects, such as the lack of symmetry of the velocity field with respect to the fully confined scheme, leading to enhanced dispersion.
(h) The set-up with multiple independent layers, admittedly a limit case, allows better exploring of the parameter space for flow experiments, reducing uncertainty associated with fluid rheology.
Our study is amenable to several possible extensions of interest for flow/transport and Newtonian/non-Newtonian fluids. In terms of geometry, these include realistic geological features such as (i) fractures/layers of uniform but different gap/permeability; (ii) localized changes in the gap/permeability; (iii) localized or diffused inter-layer mass exchange; (iv) domains of finite length, limited downstream by an edge. From a conceptual viewpoint, the study of macro-dispersion can be extended to HB fluids. Future experimental work is also of interest to shed light on HB dispersion in porous media. We are working on these topics and hope to report on them in the near future.
Acknowledgements
S.L. gratefully acknowledges the financial support from Anton Paar for co-funding Anton Paar DMA 5000 density metre. The Department of Engineering and Architecture, with its Rheometrica lab, is in the list of the Anton Paar reference research centres for the testing and development of experimental apparatus in the field of rheology. The cost of the equipment used for this experimental investigation was partly supported by the University of Parma through the Scientific Instrumentation Upgrade Programme 2018. V.D.F. gratefully acknowledges financial support from Università di Bologna Almaidea 2017 Linea Senior grant.
Declaration of interests
The authors report no conflict of interest.
Appendix A. Model description for flow in a two-dimensional (2D) porous medium
The flow in a porous medium of a power-law fluid is described by a modified Darcy's law based on the capillary bundle approach, proposed by Bird, Lightfoot & Stewart (Reference Bird, Lightfoot and Stewart1960), verified experimentally by Cristopher & Middleman (Reference Cristopher and Middleman1965) and Yilmaz, Bakhtiyarov & Ibragimov (Reference Yilmaz, Bakhtiyarov and Ibragimov2009), applied by Pascal (Reference Pascal1983); Wu & Pruess (Reference Wu and Pruess1996); Velten, Lutz & Friedrich (Reference Velten, Lutz and Friedrich1999); Kim & Hyun (Reference Kim and Hyun2004); Vajravelu, Sreenadh & Reddy (Reference Vajravelu, Sreenadh and Reddy2006); Degan, Akowanou & Awanou (Reference Degan, Akowanou and Awanou2007) and, in the specific context of gravity currents, by Longo et al. (Reference Longo, Di Federico, Chiapponi and Archetti2013, Reference Longo, Ciriello, Chiapponi and Di Federico2015a); Longo, Di Federico & Chiapponi (Reference Longo, Di Federico and Chiapponi2015b) among others. A power-law relationship at the representative elementary volume scale between flux and pressure drop for a fluid with power-law rheology was also obtained based on a network model by Balhoff & Thompson (Reference Balhoff and Thompson2006). The flow law reads as

where $p$ is the pressure,
$z$ is the vertical coordinate positive upwards,
$g$ is the specific gravity,
$\boldsymbol {u}$ is the Darcy velocity,
$\mu _{eff}$ is the effective viscosity and
$k$ is the intrinsic permeability. The ratio
$k/\mu _{eff}$ is the mobility ratio expressed as

where $\phi$ is the porosity. It is convenient to express (A 1) as

In the limits of the present approximations, we are dealing with a horizontal current advancing partly in a fully saturated domain, partly in a gravity-driven nose ahead of it. The velocity has only a horizontal component

and, in the shallow water approximation, the pressure is hydrostatic, with $p(x,y,t)=p_1+\Delta \rho \,g h(x,t)-\rho \,gz$, where
$p_1=p_0+(\rho -\Delta \rho )gh_0$, with
$p_0$, the pressure at
$y=h_0$, being constant. Hence, (A 4) can be written as

In writing (A 4) and (A 5) we have assumed that $\partial p/\partial x<0$ or
$\partial h/\partial x<0$.
In the fully saturated domain ($0<x<x_s$) the darcian uniform velocity is

where the subscript $r$ refers the variable to the
$r$-th layer. In the nose (
$x_s<x<x_f$) the mass conservation equation reads as

where $\bar {u}$ is the darcian velocity. Introducing the velocity results in

with the boundary conditions

and $h_r(x_f,t)=0\,\forall t$,
$h_r(x_s,t)=H_l\ \forall t$. Here
$Q_r$ is the flow rate in the
$r$-th layer.
Introducing the velocity, length and time scales



with the boundary conditions

The search for a self-similar solution brings to a differential problem identical to that already solved for a narrow gap. Dimensional results are obtained by considering the new scales.
Appendix B. Extension to Herschel–Bulkley fluids in a 2D porous medium
The extension of the model to a HB fluid requires the formulation of the Darcy's law, which may be written as (Chevalier et al. Reference Chevalier, Chevalier, Clain, Dupla, Canou, Rodts and Coussot2013)

where $d$ is the diameter of grains,
$\boldsymbol {\nabla } p$ the pressure gradient,
$\tau _p$ the yield stress,
$\mu _0$ the consistency index,
$n$ the flow behaviour index,
$\bar {u}$ the darcian velocity, and
$\chi$ and
$\beta$ are coefficients. The two coefficients
$\chi$ and
$\beta$ depend on the maximum width of the widest path of the flowing current and on pore size distribution and structure, respectively. Their values change for different porous media and should be measured by experiments. As a pragmatic approach, here we will use
$\chi =5.5$ and
$\beta =85$, the values reported in the experimental work of Chevalier et al. (Reference Chevalier, Chevalier, Clain, Dupla, Canou, Rodts and Coussot2013) and referred to glass beads with diameters from 0.26 to 2 mm, the same diameter range adopted in the present experiments. The flow described by (B 1) is possible only if
$|\partial p/\partial x |> \chi \tau _p/d$, otherwise a plug is formed. However, it has been demonstrated (Chevalier et al. Reference Chevalier, Rodts, Chateau, Chevalier and Coussot2014) that at very low darcian velocity only a residual, small region of the fluid domain is at rest, whereas the prevailing part of the fluid domain is in motion with a velocity density distribution similar to that obtained for a Newtonian fluid, producing percolation. This physical behaviour represents a strong advantage from an analytical point of view, since if we neglect the separation between fluid at rest and in motion, the analysis is much simpler and the results are expected to be coherent with the experiments (see Di Federico et al. Reference Di Federico, Longo, King, Chiapponi, Petrolo and Ciriello2017).
In the shallow water approximation, the pressure gradient becomes $\partial p/\partial x=\Delta \rho \,g\,(\partial h/\partial x)$ and inverting equation (B 1) the darcian velocity is equal to

under the constraint $|\partial h/\partial x|>\kappa _p$ and where
$\kappa _p=5.5\tau _p/(d\Delta \rho \,g)$. In the fully saturated domain (
$0<x<x_s$) the darcian uniform velocity is

In the nose ($x_s<x<x_f$), inserting the darcian velocity in the local mass conservation for
$\partial h/\partial x <0$ yields

where $\phi$ is the porosity. The boundary conditions are

and $h_r(x_f,t)=0\,\forall t$,
$h_r(x_s,t)=H_l\ \forall t$. Here
$Q_r$ is the flow rate in the
$r$-th layer.
Introducing velocity, length and time scales

(B 3) in dimensionless form becomes

while (B 4) becomes

with the boundary conditions

and $H_r(X_f,T)=0\ \forall T$,
$H_r(X_s,T)=1\ \forall T$.
The approach for the solution is similar to that adopted for a narrow gap, with minor differences for the coefficients of the $O(\psi )$ differential problem and with different scales for the dimensional results.